This content was uploaded by our users and we assume good faith they have the permission to share this book. If you own the copyright to this book and it is wrongfully on our website, we offer a simple DMCA procedure to remove your content from our site. Start by pressing the button below!
in the ground state for each mode as a function of the GVD. As shown, a dominant high frequency mode dictates an OC for all others, while a low frequency one allows for different OCs for each. Dephasing reduces the amplitudes of vibronic coherences in both modes (particularly in second case) and g j
g j
g j
e j
g m
e m
m≠ j
331
also shifts OC toward the transform-limit (TL). This must be due to rapid erasure of transition dipole coherence, which is the vehicle of all chirp control schemes, leading to a narrower "window" of coherent dynamics. These results are trivially obtained from the analytical expression and are in accordance with our intuitive understanding of the wave packet dynamics. Case 2: Low-Frequency Mode is Highly Displaced
Case 1: High-Frequency Mode is Highly Displaced -3
-60
-40
-20
0
20
2
Amp.
- 1,500 cm
-1
3
6
2
4 1 2 -1
υ1=1500 cm-1, ∆1=0.1 υ2= 600 cm , ∆2=4.0
0 -80
40
-60
-40
-20
0
20
0
-1
-1
υ1=1500 cm , ∆1=4.0 -1 υ2= 600 cm , ∆2=0.1
-1
Amp.
- 1,500 cm
5
X10
8
Amp.
- 600 cm
3
Amp.
- 600 cm
4
10
-1
5
-2
X10
X10
15
0 -80
-3
-4
X10
40
2
GVD [fs /RAD]
2
GVD [fs /RAD]
Fig. 1. Simulations results for two extreme cases of displacements (displacements are shown inside panels) - with (dashed lines) and without (solid lines) electronic dephasing.
Experiment A single stage NOPA centered at ~550 nm and pumped by a 30 fsec amplified Ti:Sapphire system was used to derive pump, probe and reference pulses. Precompression in BK7 prisms was followed by a deformable mirror 4-f shaper, allowing nearly full compression to the TL of ~6.5 fs. Chirp was varied by insertion or removal of differing amounts of fused silica in the pump beam path, covering a GVD range of (-140)-(+140) fs2/rad, and characterized using X-PG-FROG. Transmitted probe and reference pulses were co-dispersed in an imaging spectrograph to obtain delay dependent ∆OD spectra from 500-700 nm. Analysis of the results involves subtraction of a slowly varying background and Fourier transformation of the isolated residuals. -1
Fourier Power (a.u.) -120
-80
-40
0
40
Oxazine-1
-1
Betaine-30
1,640 cm -1 650 cm -1 230 cm
>1,500 cm -1 560 cm -1 270 cm
-120
-80
-40
0
40
2
GVD[fs /rad] Fig. 2. Modulation depth vs. GVD of pulse for a few representative modes in two systems.
332
The results, depicted in figure 2, demonstrate a substantial enhancement of the amplitude of ground state vibronic coherences using NC, particularly for high frequency modes of both molecules. Both systems demonstrate a nearly uniform OC, of only ~20 fs2/rad, regardless of the vibrational frequencies (250-1,600 cm-1). This OC broadens the pulse by a factor of only ~1.5. While in Betaine this might be indicative of rapid dephasing, [5,6], this is not the case for Oxazine-1 which exhibits a narrow absorption band with pronounced vibronic structure.
Conclusions To account for similarity of OC in both dyes, the instantaneous shift in the vertical difference potential for Betaine was calculated based on RR data [5] (Fig 3). In addition, we present a graph of the chirp rate (CR) of Gaussian pulses as a function of the GVD. The first suggest that a ~5-7 fs pulse isn't broad enough to compensate for the changes in the difference potential as required from the OC. It also demonstrates that for short times, the OC for following ∆P should be quadratic and not linear. The latter shows that for a given TL pulse width, CR reaches extrema at distinct values of GVD which broaden the pulse by 2 in agreement with the data of both dyes. Accordingly we conclude that the measured OC reflects the limited bandwidth of our source, not the polyatomic molecular systems under study. Nonetheless NC clearly enhances coherences of ground state wave packets in all active modes, and can still serve to separate excited and ground state modes in a polyatomic to some extent.
-6
2
Typical bandwidth (FWHM)
-3
Chirp Rate - [1/time , a.u.]
1.0
3
-1
∆Peg(10 cm )
0
0.5
0
5
10
Time (fsec)
15
τ max = 2τ 0
0.0 Γ0=1
-0.5
Γ0=2 Γ0=4
-1.0
-9
2
Γ0=τ0/2ln(2)
Γ0=6
-2 0 2 2 GVD - φ''(ω) - [time , a.u.]
Fig. 3. Demonstration of the limitations of excitation pulse to dynamically follow the instantaneous potential gap in Betaine (see text for details).
References: 1 2 3 4 5 6
Ruhman S. and Kosloff R., J. Opt. Soc. Am. B 7, 1748 (1990). Bardeen C.J., Wang Q. and Shank C.V., J. Phys. Chem. A 102, 2759 (1998). Malkmus S., Durr R., Sobotta C. and Pulvermacher H., Zinth W. and Braun M., J. Phys. Chem. A 109, 10488 (2005). Kovalenko S.A., Eilers-Konig N., Senyushkina T.A. and Ernsting N.P., J. Phys. Chem. A 105, 4834 (2001). Zhao X., Burt J.A. and McHale J.L., J. Chem. Phys. 121(22), 11195 (2004). Hwang H. and Rossky, J. Phys. Chem. B 108, 6723 (2004).
333
Determining Vibrational Huang-Rhys Factors by Photon Echo Spectroscopy N. Christensson, A. Yartsev and T.Pullerits Department of Chemical Physics, Lund University, P.O. Box 124, SE-22100, Lund, Sweden E-mail: [email protected] Abstract. Electronic and vibrational dephasing dynamics of Rhodamine 800 has been studied with 3PEPS. With careful analysis, the S-factors of the vibrational modes can be accurately determined. The vibrational dephasing rate displays abnormal frequency dependence.
Introduction The interaction of electronic states with the surrounding bath of nuclear motions gives rise to dephasing and population relaxation in the condensed phase. The time dependence of the interaction of the nuclear bath with the chromophore gives rise to time dependent fluctuations of the transition frequency. The auto correlation of these fluctuations, C(t), contains the information on the timescale and amplitude of the nuclear motions affecting the electronic states. The normalised correlation function, M(t), can be determined in the time domain by the three-pulse photon echo peak shift (3PEPS) experiment. In the impulsive limit it has been shown that this measurement can directly follow the slow decay of correlation function on timescales longer than the bath correlation time [1, 2]. However, intermolecular modes generally have a frequency of a few hundred wave numbers and are thus comparable to the width of the pulses used in experiments. To what extent it is possible to correctly determine the absolute coupling strength of these modes from time domain experiments with finite pulses is the main topic of this work. To get and independent measure of the spectra of the vibrational modes of the molecule in this study we employ Fluorescence Line Narrowing (FLN) [3].
Fig. 1. Temperature dependence of the FLN signal of Rhodamine 800 in ethanol matrix. Excitation laser line is at 713.8 nm. Blue dotted lines are experiment points and the dashed lines are simulation based on the theory in ref [3]. The simulations have been displaced vertically by 0.15.
334
We extract the frequencies and Huang-Rhys factors of all modes that couple to the electronic transition in Rhodamine 800. Figure 1 shows the temperature dependence from 10 to 80 K of the FLN signal of Rhodamine 800 excited at 713.8 nm using a narrow band CW Ti:Sapphire laser. The low temperature trace reveals a large number of vibrational (under-damped) modes above 100 cm-1 that couple to the electronic transition. We resort to simulations to accurately determine the correct shape of the spectral density[3]. Figure 2 shows the 3PEPS signal of Rhodamine 800 in ethanol at room temperature recorded with ~20 fs pulses centred at 700 nm (red edge of the absorption spectrum) with a FWHM of 34 nm. The experimental set-up has been describe previously [4]. The peak shift shows a multi-exponential decay originating from solvation and a complex beating pattern of multiple vibrations. A least–square curve fitting routine was employed to extract the frequencies and the dephasing times of all vibrations in the signal. We directly identify the 5 vibrational modes below 500 cm-1 seen in the FLN experiment. The apparent dephasing rate of each of the modes is shown as a function of the vibrational frequency in figure 3. We start with a direct simulation of the 3PEPS signal based on the non-linear response function formalism in the impulsive limit [1, 5]. To phenomenologically account for the finite duration of the pulses we introduce an effective spectral density via a multiplication by the power spectra of the pulses used in the experiments. The results are shown together with the experimental trace in figure 2 and we find a nice agreement between the experiment and the simulations.
Fig. 2. 3PEPS signal of Rhodamine 800 in ethanol at 290 K. Dotted line show the experimental data points and the solid line shows the simulated trace. The simulations have been displaced vertically by 1 fs for clarity. However, it’s desirable to approach the inverse problem, i.e to obtain the S-factors directly from the time domain data. To do so we assume that the peak shift signal is equal to the transition frequency correlation function. Any mode that enters the correlation function does so with relative amplitude that corresponds to its reorganisation energy. To obtain the reorganisation energy of the modes probed by the experiment we determine the total reorganisation energy by simulations of the peak shift in the impulsive limit. Since the finite bandwidth of the pulses discriminate against the high frequency vibrational modes we divide the amplitude of each mode by the amplitude of the power spectra of the pulses at the vibrational frequency. We thus arrive at the following equation:
335
Sj =
A vib j
∑A
λtot
k
1 v j E (Ω − v j )
(1)
where S is the Huang-Rhys factor, νj is the mode frequency, Aj is the amplitude obtained from the fit of the peak shift,, λtot is the total reorganisation energy, E is the power spectra of the pulses and Ω is the position of the pure electronic transition. The simulations allow for a determination of the reorganisation energy of the overdamped modes of the peak shift. When the reorganisation energy of these modes are known, we can directly calculate the relative value of the reorganisation energy of the vibrational modes as the ratio of the amplitudes obtained from the curve fitting of the 3PEPS trace. Dividing the reorganisation energy of each mode by the laser pulse envelope we obtain the S-factors of the modes impulsively excited by the laser pulses. Table 1 show the S-factors obtained, in an independent fashion, from the 3PEPS and the FLN measurements. If we use the method for obtaining the coupling strength without simulations proposed by Christensson [6], we can directly obtain the S-factors from curve fitting of the peak shift and from the knowledge of the duration of the pulses. This would amount to a considerable simplification compared to other methods like FLN or resonance Raman spectra where simulations of the experimental signals are needed to obtain the S-factors. Table 1. Comparison of the Huang-Rhys factor obtained from FLN and Photon echo experiments for the modes that can be impulsively excited with the fs pulses
ν /cm-1 S(Echo) S(FLN) ”diff” %
91 0.155 0.143 +8
219 0.120 0.137 -12
345 0.093 0.091 +2
373 0.051 0.059 -13
455 0.104 0.124 -16
Conclusions We have shown by a direct comparison of Fluorescence Line Narrowing spectra and three pulse photon echo peak shift that the amplitude of the high frequency vibrational modes are scaled by the power spectra of the laser pulses. Having obtained the total reorganisation energy of the overdamped solvent mode via impulsive limit simulations we use a simple expression to obtain the S-factors of the impulsively excited vibrational modes. 1 2 3
4
5 6 7
Joo, T.H., Y.W. Jia, J.Y. Yu, M.J. Lang, and G.R. Fleming, "Third-order nonlinear time domain probes of solvation dynamics". J. Chem. Phys., 104(16): p. 6089 (1996). deBoeij, W.P., M.S. Pshenichnikov, and D.A. Wiersma, "On the relation between the echo-peak shift and Brownian-oscillator correlation function". Chem. Phys. Lett., 253(1-2): p. 53 (1996). Personov, R.I., ed. Site selection spectroscopy of complex molecules in solutions and its applications. Spectroscopy and excitation dynamics of condensed molecular systems, ed. V.M. Agranovich and R.M. Hochstrasser. (North-Holland: Amsterdam, 1983). Dietzek, B., N. Christensson, P. Kjellberg, T. Pascher, T. Pullerits, and A. Yartsev, "Appearance of intramolecular high-frequency vibrations in two-dimensional, time-integrated three-pulse photon echo data". Phys. Chem. Chem. Phys., 9(6): p. 701 (2007). Mukamel, S., Principles of Non-linear Optical Spectroscopy (New York: Oxford University Press, 1995). Christensson N., Dietzek B., Yartsev A. and Pullerits T. submitted (2008) May, V. and O.Kuhn, Charge and Energy Transfer Dynamics in Molecular Systems (Berlin: Wiley/VHC, 1999).
336
Observation of High-Frequency Coherent Vibrational Motion with Strongly Chirped Probe Pulses D. Polli1, D. Brida1, G. Lanzani1, and G. Cerullo1 1
Dipartimento di Fisica, Politecnico di Milano, Piazza L. da Vinci 32, 20133 Milano, Italy E-mail: [email protected]
Abstract. We observe time-domain coherent vibrational wavepackets at 1585-cm-1 frequency (21-fs period) using broadband probe pulses strongly chirped up to 150-fs duration. The results are explained using the chronocyclic (Wigner) representation of the chirped pulse.
During the last decade, the availability of broadly tunable few-optical-cycle light pulses has dramatically improved the resolution of time-domain vibrational spectroscopy, allowing the direct detection of vibrational motions at frequencies as high as 2100 cm-1 (16 fs period) and providing new insights into chemical and structural rearrangements in all phases of matter [1]. The pump pulse excites a molecule on a timescale shorter than that of nuclear vibrational motion, generating coherent vibrational wavepackets in both the excited and the ground potential energy surfaces. The wavepacket motion is tracked by following the time-dependent transmission modulations of a delayed probe pulse. Traditionally, both pump and probe pulses with nearly transform-limited (TL) duration are used and the time resolution of the experiment is taken as the instrumental response function, which is the cross-correlation function between pump and probe pulses. While the role of pump pulse chirp in changing the relative weight of ground and excited state coherences has been studied [2], the effects of probe pulse chirp have not been investigated. In this work, we combine sub-10-fs visible pulses with a high-sensitivity broadband detection setup using a fast spectrometer to study coherent vibrational dynamics in organic molecular systems. We demonstrate that it is possible to observe high-frequency coherent wavepacket dynamics employing strongly chirped broadband probe pulses with duration almost 10 times longer than the period of the detected vibrational mode. This surprising result is explained by analysing the chirped probe pulse in terms of its chronocyclic (Wigner) representation. The experimental setup starts with a commercial regeneratively-amplified Ti:Sapphire laser delivering 150-fs pulses at 790 nm with 500-μJ energy and 1-kHz repetition rate. A home-built non-collinear optical parametric amplifier then generates ultrabroadband pulses with μJ-level energy and spectrum spanning the 500-700 nm wavelength range. The pulses are compressed down to nearly TL ≈6-fs duration, measured by Frequency Resolved Optical Gating (FROG), by chirped mirrors and sent to a degenerate pump-probe setup. After the sample the probe beam is sent to an optical multichannel analyser with fast electronics, allowing single-shot recording of the probe spectrum at the full 1-kHz repetition rate [3]. Differential transmission (ΔT/T) maps as function of probe wavelength and delay are acquired. The probe pulse is chirped by inserting different blocks of dispersive material along its path before the sample.
337
We studied thin films with blends of poly-phenylene vinylene (PPV) and C60, with 1:3 molecular weight ratio. The 6-fs pump pulse resonantly excites PPV, creating a singlet exciton; rapid electron transfer (ET) to the fullerene then occurs with ≈50 fs time constant, leaving the polymer chain in a metastable polaron excited state [4]. The vibrational coherence created by the pump pulse in the excited state is rapidly quenched by the ET process, while the ground state coherence, induced via impulsive stimulated Raman scattering, is preserved. The periodic modulations in the ΔT/T signal observed at time delays longer than ≈200 fs are thus solely attributed to ground state vibrational coherence.
Fig. 1. (a, b, c) ΔT/T plots as a function of probe wavelength and delay in PPV-C60 for different values of probe pulse chirp; (d, e) pump-probe dynamics at selected probe wavelengths (as indicated) for transform-limited (d) and strongly-chirped (e) probe pulses; (f) solid lines: spectral phase of the Fourier transform at 1585 cm-1 frequency (C=C stretching) and corresponding group delay for the transform-limited probe pulse (TL) and the chirped ones; dashed lines: simulated group delays introduced by the sapphire plates.
Figure 1(a) shows the experimental results for the oscillatory component of the ΔT/T signal in the 510-580 nm wavelength region and in the 200-400 fs temporal window using TL probe pulses. We observe a complex vibrational pattern due to the beating of two modes at ν1≈1320 cm-1 (T1≈25 fs) and ν2≈1585 cm-1 (T2≈21 fs), corresponding to the single and double carbon bond stretching respectively. The phase of the oscillations is almost flat across the whole probe spectrum (see also Fig. 1(d)), in accordance with expectations for a ground-state mode to the red of the steady-state absorption peak. Figures 1(b-c) show the results for strongly-chirped probe pulses obtained by dispersing them in sapphire plates with thickness 1 mm (Fig. 1(b)) and 2.5 mm (Fig. 1(c)), corresponding to group delay dispersions of 100 fs2 and 250 fs2, respectively; the corresponding probe pulsewidths, measured by FROG, are 55 fs and 150 fs. Astonishingly, even with probe pulse that are ≈7 times longer than the period of oscillation, the oscillatory pattern is still visible with high contrast, although it is now strongly bent (note the 2π phase drift occurring in the 15-nm spectral region plotted in Fig. 1(e)). This observation is possible only thanks to the combination of high temporal and spectral resolution of our instrument, because the fringe pattern would be cancelled in open-band or pass-band (≈10 nm bandwidth) detection schemes. Fig. 1(f) shows the probe wavelength dependence of the spectral phase φ for the 1585 cm-1 mode (solid line) and the corresponding group delay (GD) calculated as
338
GD=φ*T2/2π. The results are in very good agreement with the calculated GD introduced by the sapphire windows (dashed lines in Fig. 1(f)). To understand our results, we analysed the broadband probe pulse using the chronocyclic representation [5], also known as Wigner Distribution (WD), which provides information on the temporal distribution of the different spectral components of the pulse. Fig. 2(b) shows the WDs for the TL pulse and the strongly chirped one. While for the TL pulse all the spectral components occur in a short time window (see dashed line in Fig. 1(a)), the WD for the chirped pulse is elongated but maintains a narrow shape. Therefore, while the wavelength integrated pulse temporal profile is long (≈150 fs, see solid line in Fig. 2(a)), cuts at selected wavelengths are still very short, almost comparable to the TL pulse duration (see Fig. 2(c)). It is thus possible to preserve a very high temporal resolution in the pump-probe experiments even with a strongly chirped probe pulse.
Fig. 2. Chronocyclic representation of the TL and chirped probe pulses (panel (b)) and corresponding wavelength integrated pulse profiles (panel (a)) and cuts at selected wavelengths of the chirped pulse (panel (c)).
In conclusion, we have shown that it is possible to measure very fast dynamics and/or high frequency coherent vibrational oscillations using strongly chirped broadband pulses, provided that a suitable detection setup with high spectral resolution is used. This result becomes important in those spectral ranges where it is easy to generate broadband spectra (e.g. by self-phase modulation) but temporal compression is not easy to implement. References 1 2 3 4 5
S. De Silvestri, G. Cerullo, G. Lanzani, eds., Coherent Vibrational Dynamics, CRC Press, Boca Raton, 2008. C. J. Bardeen, Q. Wang, and C. V. Shank, Phys. Rev. Lett. 75, 3410 (1995). D. Polli, L. Lüer and G. Cerullo, Rev. Sci. Instrum. 78, 103108 (2007). Ch. Brabec, G. Zerza, G. Cerullo, S. De Silvestri, S. Luzzati, J.C. Hummelen, S. Sariciftci, Chem. Phys. Lett. 340, 232 (2001). J. Paye, IEEE J. Quantum Electron. 28, 2262 (1992).
339
Coherent Transfer of Molecular Vibrations in the Electronic Excited States Chul Hoon Kim, Sohyun Park, Intae Eom, and Taiha Joo Department of Chemistry, Pohang University of Science and Technology (POSTECH), Pohang 790-784, South Korea E-mail: [email protected] Abstract. Coherent wave packet motions in the electronic excited states prepared by impulsive nuclear rearrangements such as electronic transition, internal conversion, and chemical reaction are observed exclusively by ultrafast 35 fs time-resolved spontaneous fluorescence. Direct information on the excited state dynamics, reaction coordinates, and coupling between excited states can be obtained.
Introduction When a molecule is excited impulsively by a short pulse of light, coherent vibrational wave packets are created, which are manifested in the oscillations of the time trace in various time-resolved spectroscopies, pump/probe transient absorption (TA) being a typical example. The nature of the vibrational modes excited and their decay provide a wealth of information on the dynamics and molecular structures of the states involved. In general, the coherent nuclear wave packet can be launched by any nuclear rearrangements such as chemical reaction and internal conversion that occur faster than roughly half of the vibrational period. Excited state intramolecular proton transfer (ESIPT) is one such example. Lochbrunner et al. have observed characteristic coherent oscillation components in the excited state absorption and stimulated emission from the ESIPT reaction product keto form of the 2-(2’hydroxyphenyl)benzothiazole by TA experiments [1,2]. They showed that these oscillations are nearly the same as the low frequency skeletal motions which modulate the distance between the proton donor and acceptor groups. In this work, we report the observation of the coherent wave packet motions of the molecules in the excited states created by various processes such as chemical reaction and internal conversion from Sn (n≥2) to S1. These observations give direct information on the structures of the excited states and the reaction coordinates by examining the modes excited. We used femtosecond time-resolved spontaneous fluorescence (TRF), because interpretation of a TA signal is not always straightforward due to the fact that a TA signal consists of several contributions originating from the ground state, excited state, and all product states. On the other hand, TRF provides information on the dynamics of the excited state exclusively, although typical time resolution of the TRF measurement is limited to around 200 fs, which is not high enough to observe coherent vibrational motions.
Experimental Methods A home-built Kerr lens mode-locked cavity-dumped Ti:sapphire laser and a homebuilt cavity-dumped optical parametric oscillator (OPO) employing a periodicallypoled lithium niobate were used as light sources. Details of the OPO have been described elsewhere [3].
340
TRF was measured by fluorescence up-conversion technique. Detailed description of the up-conversion apparatus employing noncollinear sum frequency generation has been described elsewhere [4]. To achieve ultrafast time resolution of ~35 fs comparable to that of TA experiment, we carefully minimized the group velocity dispersion, group velocity mismatch, and phase front mismatch in a nonlinear crystal. Instrument response of the apparatus was estimated to be as short as 35 fs (full width at half maximum) from the cross-correlation of the scattered pump and gate pulses.
Results and Discussion First, we prepared rhodamine B in S1 state by the internal conversion from Sn (n≥2) state. It is well known as the Kasha’s rule [5] that the internal conversion from the higher excited state Sn (n≥2) to S1 is ultrafast to give fluorescence from the S1 state exclusively. Figure 1. shows the TRF of rhodamine B in methanol detected at different wavelengths followed by the excitation to S2 state. At early times, TRF signals show a detection wavelength independent slow rise of around 130 fs due to the vibronic relaxation in S1. Stokes shift of rhodamine B is rather small to give minor detection wavelength dependence due to the intermolecular solvation process. At longer times, the TRF signals show the typical behavior in picosecond time scales due to the dielectric relaxation of the solvent; rise at longer detection wavelengths and a decay at shorter wavelengths. More interestingly, oscillations of the TRF signals at 155, 203, and 277 cm-1 are observed at all detection wavelengths, which are due to the coherent wave packet motions. Thus, the internal conversion from S2 to S1 is much faster than the 100 fs period of the highest frequency vibrational modes observed. Once the structure of the S1 state is calculated quantum mechanically, assignments of the observed vibrational modes lead to the qualitative description of the structure of the S2 state, since the vibrational excitation in the product state (S1) is directly proportional to the projection of the displacement between the reactant (S2) and the product onto the vibrational normal modes of the product state [6], in analogy to the Franck-Condon principle in the electronic transition of the molecules. (b)
80
(c)
6
40 4
0
2
-40 650 nm
-80
0 6
80
1600
Intensity (cps)
Residual
(a)
1400 1200
40
1000
-80
800
80
600
40
400
0
200
-40
0 -200 -100
4
0
2
-40 600 nm
0 2
1
560 nm
-80
0
100
200
300
Time Delay (fs)
400
500
0
0
200 400 600 800 1000 1200 1400
Time Delay (fs)
0
100
200
300
400
500
600
700
Wavenumber (cm-1)
Fig. 1. (a) TRF of rhodamine B in methanol. The exponential fit and the instrument response are also indicated. (b) Residuals from the exponential fits of the TRF signals at different wavelengths. (c) Frequency spectra of the oscillation components obtained by the linear prediction singular value decomposition method.
341
We also prepared a molecule in the excited state impulsively by an ultrafast chemical reaction. Figure 2 shows the intramolecular charge transfer (ICT) of laurdan. The ICT of laurdan is essentially the same as the ICT of much studied dimethylaminobenzonitrile (DMABN). The ICT state of DMABN is usually called a twisted intramolecular charge transfer (TICT) state. The TICT hypothesis, however, is highly controversial, and recent work found strong evidence that the perpendicular twist of the amino group is not necessary to reach the ICT state [7]. To study the molecular dynamics of the ICT reaction, we have obtained the TRF spectra over the whole emission range of the reactant and the product directly at 50 fs resolution without the spectral reconstruction method, which are required to separate the ICT reaction and the solvation dynamics. The ICT reaction in methanol occurs by 50 fs, 300 fs, and 10 ps, where the 10 ps time constant correlates well with the dielectric relaxation time of the solvent molecules. Moreover, when the fluorescence is detected at the product ICT state with 35 fs resolution, we have observed an oscillation at 510 cm-1 due to the wave packet motion of the ICT state. Ab initio calculations identified that the vibration corresponds to the twisting motion of the amino group, which unambiguously identifies the twisting motion as the ICT reaction coordinate.
Fig. 2. Frequency spectra of the oscillation components in the TRF of laurdan detected at the emission of the intramolecular charge transfer state obtained by the sliding window (500 fs width) Fourier transformation of the residual from the exponential fits
Conclusions Coherent nuclear motions are observed for molecules in the excited state in a variety of systems prepared by direct electronic excitation, non-adiabatic electronic transition, and chemical reactions. This was possible by the ultimate time-resolution in the TRF measurement. The wave packet dynamics provides a wealth of information on the dynamics of the system and the structures of the states involved. 1 2 3 4 5 6 7
342
S. Lochbrunner, A. J. Wurzer, and E. Riedle, J. Chem. Phys. 112, 10699, 2000. S. Lochbrunner, K. Stock, and E. Riedle, J. Mol. Struct. 700, 13, 2004. C. K. Min and T. Joo, Opt. Lett. 30, 1855, 2005. H. Rhee and T. Joo, Opt. Lett. 30, 96, 2005. M. Kasha, Discuss. Faraday. Soc. 9, 14, 1950. J. R. Reimers, J. Chem. Phys. 115, 9103, 2001. K. A. Zachariasse, S. I. Druzhinin, W. Bosch, and R. Machinek, J. Am. Chem. Soc. 126, 1705, 2004.
Ultrafast Isomerization Dynamics of Biomimetic Photoswitches J. Briand1, D. Sharma1, J. Léonard1, J. Helbing2, A. Cannizzo3, M. Chergui3, V. Zanirato4, S. Haacke1 and M. Olivucci5 1 Institut de Physique et Chimie des Matériaux de Strasbourg, UMR 7504 ULP – CNRS, F-67034 Strasbourg, France 2 Physikalisch-Chemisches Institut, Universität Zürich Witerthurerstr. 190, CH-8057 Zürich, Switzerland 3 Laboratoire de Spectroscopie Ultrarapide, ISIC – EPFL, CH-1015 Lausanne, Switzerland 4 Dipartimento di Scienze Farmaceutiche, Università di Ferrara, 44100 Ferrara, Italy 5 Dipartimento di Chimica, Università degli Studi di Siena, 53100 Siena, Italy & Chemistry Dept., Bowling Green State University, Bowling Green, OH 43403, USA Email: [email protected]
Abstract. Femtosecond UV-VIS and mid-IR experiments show that a new class of biomimetic photoswitches photo-isomerizes in less than 300 fs. In close analogy to rhodopsin, the isomerization is driven by ultrafast motion along the stretch and the torsional coordinates.
Introduction The ultrafast reaction dynamics of the methylated methoxy-IP Schiff base (MeO-IP) compound (Inset fig. 1) is presented. This molecule is part of a family of indanylidene-pyrroline (IP) compounds recently synthesized with the aim of mimicking the photochemistry of retinal Schiff bases in rhodopsins [1]. Combined ab initio quantum mechanics/molecular mechanics (CASPT2/CASSCF/AMBER) calculations in methanol indicate that the relaxation in the excited state is barrierless for torsion around the central double bond, leading to a conical intersection (CI) at almost 90° torsion. This motivated the expectation for the photo-initiated twist to occur on a sub-picosecond time scale. Absorption spectra of the Z and E forms are dominated by intense π−π* transitions at 392 and 385 nm, respectively. The isomerization quantum yield is 0.20-0.22 [1]. The present work shows that the Z→E photo-isomerization occurs indeed within less than 300 fs; similar to the record values held by retinal proteins [2]. As for the latter, the excited state dynamics leading to the CI seems to be well described by a two-mode scenario involving motion along stretch and torsional coordinates.
Results and discussion We used polychromatic fluorescence up-conversion [3] to obtain the time-resolved spectra and kinetic traces of Z-MeO-IP after 400 nm excitation (fig. 1A). The fluorescence, centred at 530 nm, covers most of the visible spectral region, rises instantaneously and shows a bi-phasic decay. The decay constants, obtained from bi-exponential fits convoluted with the 120 fs instrument response function, are τ1 < 40 fs (limited by IRF) and τ2 = 300 ± 30 fs. At wavelengths > 600 nm, a mono-exponential fit with the longer time suffices to reproduce the data.
343
We attribute the observed fluorescence to the directly excited singlet state S1, as it lies to the red of the absorption band and appears promptly. The fast fluorescence decay at short wavelengths could reflect an initial ultrafast relaxation along the stretching coordinate, or may be due to molecules with high torsional kinetic energy affording fast isomerisation (see below). On the other hand, the slower component represents an excited state population decaying non-radiatively within 0.3 ps. Transient absorption data show prompt excited state absorption (ESA, 1 ps), the anisotropy becomes the same at all frequencies, due to the slow components (~1 ps) of the spectral diffusion of liquid water [6]. If the absorption band is pumped in the center or in the red wing, the initial anisotropy will be close to 0.4 and has the same dynamics at all probe frequencies (Fig. 2a). This frequency independence follows from the fact that the excited molecules do not directly reorient, but first have to diffuse spectrally to the blue wing before they can undergo the reorientation and frequency jumping process. After the jump the O-D oscillator can return at all frequencies in the absorption band, leading to the same decay of the anisotropy at all frequencies, with an effective time constant of 2.5 ps. This finding agrees with previous femtosecond studies of the O-D stretch vibration of HDO:H2O in which the anisotropy dynamics were also observed to be the same at all probe frequencies following excitation in the center [4,5].
503
Fig. 2. Anisotropy as a function of delay for pump frequencies of 2500 cm-1 (a) and 2650 cm-1 (b), and 2650 cm-1 and probe frequencies of and probe frequencies of 2500 cm-1 (circles), 2550 cm-1 (squares), and 2600 cm-1 (triangles). The solid curves in the figures are calculated with the same model that was used to calculate Fig. 1b.
In Fig. 2b the anisotropy dynamics are shown at different probe frequencies following excitation in the blue wing. When probing in the red wing, the anisotropy acquires a low initial value, because the signal is dominated by molecules that have jumped from the excited blue wing of the spectrum to lower frequencies. Then the anisotropy increases because of the spectral equilibration with the molecules in the blue wing that have not reoriented and that still have a high value of the anisotropy. After reaching a maximum at ~500 fs, the anisotropy shows the same decay of 2.5 ps that is observed at all other pump and probe frequencies for delays >1 ps. This time constant is determined by the fraction of the molecules that show the reorientation and frequency jump, and the rate at which molecules diffuse to this state. The experimental results are all well reproduced with a model that includes the spectral diffusion and a frequency-dependent probability for frequency jumping and reorientation (Fig. 1b and solid curves of Fig. 2).
Discussion and Conclusion Our results are consistent with the recent observation that the blue wing of the absorption band shows a very rapid spectral diffusion effect that is absent in the red wing [6]. Our findings also agree with the recently proposed molecular jump model for reorientation [1,2]. In this model the reorientation of a water molecule proceeds via a bifurcated hydrogen bond with two other water molecules. The bifurcated transition state decays by breaking one of the hydrogen bonds while strengthening the other bond, which leads to a large change in the frequency of the O-D vibration. According to the theoretical work, the probability to evolve to the bifurcated transition state does not depend on frequency, which disagrees with the present findings. We hope that the present results will stimulate further theoretical investigations of the relation between the hydroxyl frequencies of the water molecule and the probability for reorientation. 1 D. Laage and J. T. Hynes, Science 311, 832, 2006. 2 D. Laage and J. T. Hynes, Chem. Phys. Lett. 433, 80, 2006. 3 H.-K. Nienhuys, R.A. van Santen, and H.J. Bakker, J. Chem. Phys. 112, 8487, 2000. 4 T. Steinel, J. B. Asbury, J. Zheng, and M. D. Fayer, J. Phys. Chem. A 108, 10957, 2004. 5 Y. L. A. Rezus and H. J. Bakker, J. Chem. Phys. 123, 114502, 2006. 6 J. J. Loparo, S. T. Roberts, and A.Tokmakoff, J. Chem. Phys. 125, 194522, 2006.
504
Ultrafast Temperature Jumps in Liquid Water Studied by Infrared-Pump and X-ray AbsorptionProbe Spectroscopy G. Gavrila1, Ph. Wernet1, K. Godehusen1, C. Weniger1, E. T. J. Nibbering2, Th. Elsaesser2, and W. Eberhardt1 1
BESSY, Albert-Einstein-Str. 15, D-12489 Berlin, Germany Max-Born-Institut für Nichtlineare Optik und Kurzzeitspektroskopie, Max-Born-Str. 2 A, D12489 Berlin, Germany
2
Abstract. We report the first time-resolved x-ray absorption study of liquid water. Structural changes in the hydrogen-bond network as induced by resonant femtosecond-infrared excitation are monitored via transient x-ray absorption at the oxygen K-edge.
Liquid water consists of an extended molecular network with highly polar water molecules that are coupled via intermolecular hydrogen bonds. Structural dynamics of this network including the breaking and making of hydrogen bonds occur on femto- to picosecond time scales set by the strength of intermolecular interactions and the different vibrational and translational motions. Ultrafast vibrational spectroscopy has provided detailed information on such couplings and the resulting ultrafast processes [1]. Transient vibrational spectra give, however, only indirect spectroscopic insight into time-dependent molecular arrangements. In contrast, x-ray diffraction and x-ray absorption have recently been applied to water [2,3], providing direct access to the time-averaged equilibrium structure through radial distribution functions of the oxygen atoms [2] and local configurations of hydrogen bonds [3]. In view of the extremely fast structural fluctuations, an extension of x-ray methods into the ultrafast time domain holds great promise for unraveling structural dynamics in liquid water on their intrinsic time scales. Here we investigate the structural changes in the hydrogen bond network of liquid water upon an ultrafast temperature jump of a few degrees Kelvin. We apply femtosecond infrared (IR) pulses to excite the intramolecular O-H stretching vibration, and monitor the transient response in the oxygen K-edge x-ray absorption spectrum (Fig. 1 (a)) with picosecond x-ray pulses. Picosecond x-ray absorption spectroscopy has successfully been used before in studies of transient structure of electronically excited molecules [4]. Here, we used it for the first time to unravel ultrafast structural fluctuations in the electronic ground state of a molecular liquid with an unprecedented combination of ultrafast IR and x-ray spectroscopies. The experimental set-up consists of an amplified femtosecond Ti:sapphire laser system driving an optical parametric amplifier to generate intense mid-infrared pulses. For our experiment the infrared pulse is tuned to the O-H stretching band (3400 cm-1) of liquid water. The infrared pulse energies are about 2.2 μJ at the sample and the spot size is in the order of 100 μm (diameter for the full width at half maximum intensity). The x-ray probe pulses are generated in the soft x-ray undulator beamline UE56-1 PGM-B at the synchrotron radiation source BESSY. We used the single bunch mode of the electron storage ring, with one electron bunch in the ring
505
and with a corresponding repetition rate of 1.25 MHz (800 ns between two x-ray pulses). The x-ray pulse width was 70 ps. X-ray spot sizes larger than the infrared spot and single-bunch mode were used to avoid x-ray irradiation induced changes of the samples. Samples consisted of thin liquid water films with typical thicknesses of 100-500 nm held in the vacuum chamber between two lithographically made x-ray transparent silicon nitride (Si3N4) membranes. The oxygen K-edge absorption spectrum (Fig. 1 (a)) arises from transitions of oxygen 1s electrons into empty molecular orbitals of the probed water molecules. Transitions to these empty states are particularly sensitive to hydrogen bonding with the nearest neighbours and the spectrum reflects the ensemble average of spectral contributions from water molecules in various configurations with frozen-in geometries. The three prominent features, labelled as the pre-, main- and post-edge and, are related to different hydrogen bonding configurations [3]. In particular, the pre-edge peak at 535 eV (see Fig. 1 (b)) can be assigned to locally asymmetric configurations with one strong and one weak/broken hydrogen bond on the hydrogen side (donor hydrogen bond).
Fig. 1. (a) Oxygen K-edge x-ray absorption spectrum of liquid water [7]. (b) Steadystate transmission close to the pre edge (535 eV). (c) Change of x-ray transmission ΔT/T0 at the pre-edge as a function of pump-probe delay. Measured data are shown as markers and the fit using a Gaussian (FWHM=70 ps) broadened step function is shown as solid line.
Fig. 2. (a) X-ray transmission of liquid water at the oxygen K edge. (b) Transient changes of the x-ray transmission at a pump-probe delay of 280 ps. (Transmission change ΔT/T0, ΔT =(T-T0) with T0 and T transmission before and after excitation). Measured raw data are shown as markers (squares) and smoothed data are shown as solid line [7].
The change of x-ray transmission ΔT/T0 with ΔT=(T-T0) at the pre-edge photon energy of 535 eV (see Fig 1 (b)) is plotted as a function of pump-probe delay (symbols) in Fig. 1 (c). T0 and T are the transmission before and after excitation. The sample transmission decreases in a step-like fashion by approximately 0.2 % and stays constant up to the longest measured delay time of 280 ps. The time evolution of the transmission decrease follows the time-integrated cross-correlation function of the
506
femtosecond pump and the 70 ps x-ray probe pulse (solid line). The width of the measured transmission step is hence determined by the duration of the probe pulses. Upon femtosecond excitation, the O-H stretching vibration displays a population relaxation with a 200 fs time constant by vibrational redistribution through the intramolecular bending and intermolecular librational and hydrogen bond vibrational modes of the hydrogen bond network. Full equilibration into a heated water sample is reached within a few picoseconds [5]. The related ultrafast temperature jump is monitored here for the first time with the changes in x-ray transmission at the oxygen K edge. The associated structural changes are characterized in more detail with the transient x-ray absorption spectrum measured at a fixed pump-probe delay of 280 ps (Fig. 2 (b)). The pre-edge transmission decrease and the concurrent post-edge increase directly indicate that the number of molecules in locally asymmetric configurations with one weak/broken donor hydrogen bond increases [3], fully consistent with a temperature jump in the probed water volume. With the heat capacity of water and with our experimental parameters we estimate a temperature jump of 1–2 K, averaged over the entire probed volume. A comparison with steadystate x-ray absorption spectra measured at different temperatures [6] corroborates this estimate [7]. This demonstrates that our method serves as a very sensitive probe of transient structural changes in liquid water. In conclusion, we report on the first infrared pump – x-ray probe study of structural dynamics in liquid water on a picosecond time scale. We demonstrate the feasibility of such combined laser-synchrotron experiments with very high structural sensitivity. Our results pave the way for future experiments with substantially shorter x-ray pulses such as generated with a femtosecond slicing scheme. In this way, the ultrafast structural changes of the hydrogen bond network during and immediately after the disposal of excess energy will become accessible. This establishes a direct link between femtosecond vibrational spectroscopy and structural information. Acknowledgements. We thank Nils Huse for valuable support during the early stages of this experiment and Karsten Holldack, Christian Stamm, Ulrich Schade and Torsten Quast for their assistance. We also gratefully acknowledge financial support from the Deutsche Forschungsgemeinschaft (No. SPP1134). 1 2 3 4 5 6 7
E. T. J. Nibbering, T. Elsaesser, Chem. Rev. 104, 1887, 2004. T. Head-Gordon, G. Hura, Chem. Rev. 102, 2651, 2002. Ph. Wernet, D. Nordlund, U. Bergmann, M. Cavalleri, M. Odelius, H. Ogasawara, L. Å. Näslund, T. K. Hirsch, L. Ojamäe, P. Glatzel, L. G. M. Pettersson, A. Nilsson, Science 304, 995, 2004. M. Khalil, M.A. Matcus, A. L. Smeigh, J. K. McCusker, H.H. W. Chong, R. W. Schoenlein, J. Phys. Chem. A 110, 38, 2006. S. Ashihara, N. Huse, A. Espagne, E. T. J. Nibbering, T. Elsaesser, J. Phys. Chem. A 111, 743, 2007. U. Bergmann, D. Nordlund, Ph. Wernet, M. Odelius, L. G. M. Pettersson, A. Nilsson, Phys. Rev. B 76, 024202, 2007. Ph. Wernet, G. Gavrila, K. Godehusen, C. Weniger, E.T.J. Nibbering, T. Elsaesser, W. Eberhardt, Appl. Phys. A, DOI 10.1007/s00339-008-4726-5
507
Influence of the Environment on Reaction Dynamics: Excited State Intramolecular Proton Transfer in the Gas Phase and in Solution C. Schriever, S. Lochbrunner1, and E. Riedle Lehrstuhl für BioMolekulare Optik, Ludwig-Maximilians-Universität München, Oettingenstr. 67, D-80538 Munich, Germany 1 present address: Institut für Physik, Universität Rostock, Universitätsplatz 3, D-18055 Rostock, Germany e-mail: [email protected] Abstract. Femtosecond transient absorption reveals very similar excited state intramolecular proton transfer and associated wavepacket dynamics in the gas phase and in solution. There are striking differences for the kinetics associated with the subsequent internal conversion.
Unified probe process in the gas phase and in solution Ultrafast molecular processes are governed by intramolecular motions at the speed of skeletal vibrations as well as the interaction with the surrounding medium. In particular only little is known about the influence of the environment on the coherent wavepacket motion and how environment induced variations of the wavepacket motion change the outcome of a process. To understand this interplay we investigate the ultrafast intramolecular excited state proton transfer (ESIPT) of 2-(2'-hydroxyphenyl)benzothiazole (HBT; cf. Fig. 1a) in the gas phase and compare the dynamics to the one found in solution [1]. ESIPT is a prototypical process for very fast chemical reactions since it leads to the breaking of the bond between the reactive hydrogen atom and the donating oxygen atom (enol tautomer, UV absorption) and the simultaneous formation of a new bond between the reactive hydrogen atom and the accepting nitrogen atom (keto tautomer, fluorescence in the visible).
Fig. 1. (a) Transient absorption probed at 510 nm after exciting HBT at 350 nm in the gas phase and in a cyclohexane solution. The depicted sliding window Fourier transforms [(b) gas phase, (c) solution] are obtained after subtracting the exponential contributions from the transients.
508
It is crucial to use the same probe process for both experiments; Otherwise the probe process projects the wavefunction onto different manifolds of final states, resulting in different signatures even if there are no changes in the molecular dynamics. We choose transient absorption as probe signal since it provides a rich amount of spectroscopic information revealing detailed insight into the dynamics and it allows for a time resolution superior to most other techniques. A 20 fs time resolution and a sensitivity of 10-6 ΔOD allow us for the first time to compare directly the coherent wavepacket motion of low vapor pressure, medium sized molecules in the gas phase to the one in solution [2].
Mechanism of the excited state intramolecular proton transfer: Transfer time and coherent wavepacket motion We measure the evolution of the transient HBT absorption at various probe wavelengths after photoexcitation at 325 and 350 nm (first absorption band) in the gas phase as well as in a cyclohexane solution. Figure 1 shows the transient transmission change at 510 nm for 350 nm excitation and the sliding window Fourier transforms of the oscillatory components. The transients show a transmission decrease at time zero which is caused by the excited state absorption (ESA) and occurs immediately after the pump pulse has promoted the molecule to the S1 state. The transmission increase follows with a delay of 35 fs and originates from the emission that occurs when the electronically excited keto form is populated by the ESIPT. The delay is therefore identified with the transfer time [1]. A comparison with a precision of better than 5 fs shows that it occurs in both environments with the identical delay of 35 ± 5 fs [2]. From the similarity of the proton transfer time and the subsequent signal signatures which result from the coherent wavepacket motion during and after the proton transfer we conclude that the ESIPT proceeds in the same way in the gas phase and in solution and the environment has only a negligible influence on the ultrafast reaction. For the transients recorded in cyclohexane solution we find dominant modes at 113 cm-1 and 255 cm-1. A comparison to ab-initio calculations reveals that the modes are in-plane deformations of the molecular skeleton which reduce the donor acceptor distance. This contraction allows for a barrierless reaction path and an efficient mixing between the electronic configurations of the enol and the keto form. In the gas phase we observe the 255 cm-1 oscillation with the same frequency and relative phase as in solution (Fig. 1a). A significant contribution of the 113 cm-1 mode cannot be established. In the gas phase a very strong modulation with an even lower frequency of 41 cm-1 is observed which is absent in solution (see Fig.1). However, this mode is not related to the ESIPT but to the IC as it is discussed below. It most likely obscures contributions from the 113 cm-1 mode in the gas phase. The dephasing of the oscillatory contributions occurs with a time constant of ~1 ps for the 255 cm-1 mode both in the gas phase and in solution. The dephasing times in both environments slightly decrease with the excess energy. These findings clearly show that the vibrational dephasing is in this case intrinsic to the molecule and purely of intramolecular origin. High vibrational levels are populated which experience strong coupling to other modes and anharmonicities at the S1 keto minimum of the potential energy surface (PES). In line with this minor influence of the solvent, the vibrational frequencies also do not change. This is in striking contrast to the electronic dephasing which is accelerated by many orders of magnitude via the solvent interaction.
509
Internal conversion through a conical intersection In cyclohexane solution the decay of the electronically excited state, as measured by the stimulated emission decrease, occurs with a time constant of ~100 ps. It reflects the internal conversion (IC) back to the electronic ground state. For the isolated molecule, we observe that the emission decays much faster with a time constant of 2.6 ps for excitation at 325 nm and somewhat slower for 350 nm excitation. Both high level calculations of the PES and classical mechanics trajectories support that the 41 cm-1 mode found in the transients is associated with a torsion of the molecule around the phenyl-thiazole bond leading to a conical intersection at a torsional angle of 90° [3]. Upon IC the molecule can return to the S0 enol configuration or form the metastable S0 trans-keto tautomer. Clear spectral signatures for the latter process are found in our data. The time constant of the trans-keto formation corresponds to the decay of the stimulated emission, supporting this model.
Fig. 2. Model for the potential energy surface along the torsion around the phenyl-thiazole bond leading to a conical intersection of the S1 and the S0 state.
The 41 cm-1 contribution to the transients shows that at least the first step in the IC of the isolated molecule is associated with wavepacket dynamics. Contrary to the ESIPT, the reaction path seems to involve a significant energy barrier which leads to the excess energy dependence (see Fig. 2). The pronounced initial decay of the gas phase transients indicates that a compact wavepacket is partially leaving the detection window of the probe process before it spreads and dephases. For later times, the emission decay can be described as a rate like process. The 41 cm-1 motion is associated with a large amplitude twisting of the entire molecular skeleton and should be subject to frictional forces in solution which cause a significant dissipation of kinetic energy. The wavepacket motion along this coordinate is overdamped in solution and no oscillatory behavior is observed. The excess energy of the torsional motion is much higher in the gas phase and the IC much faster than in solution. Friction changes the character of the process from a more or less ballistic wavepacket motion to a rate governed process in solution. 1 2 3
510
S. Lochbrunner, A. J. Wurzer, and E. Riedle, J. Phys. Chem. A 107, 10580 (2003). C. Schriever, S. Lochbrunner, E. Riedle, D. J. Nesbitt, Rev. Sci. Instrum. 79, 013107 (2008). M. Barbatti, A. J. A. Aquino, H. Lischka, C. Schriever, S. Lochbrunner, and E. Riedle, in preparation.
Ultrafast 2D-IR spectroscopy of a molecular monolayer Jens Bredenbeck1,2, Avishek Ghosh1, Marc Smits1, Mischa Bonn1. 1
FOM Institute for Atomic and Molecular Physics, Kruislaan 407, 1098 SJ, Amsterdam, the Netherlands 2 Institut für Biophysik, Universität Frankfurt, Max von Laue-Str. 1, 60438 Frankfurt, Germany E-mail: [email protected] Abstract. We report on ultrafast 2-dimensional vibrational surface spectroscopy, providing information on coupling and energy transfer between vibrations of surface molecules. As a 4th order technique, it is bulk-forbidden in centrosymmetric materials and hence surface specific.
Introduction Coupling and energy flow through vibrational modes at surfaces and interfaces are important in areas as diverse as heterogeneous catalysis, electrochemistry, and membrane biophysics and –chemistry [1]. Furthermore, vibrational coupling patterns contain information on molecular structure, a feature already explored in bulk experiments to measure structure parameters with femtosecond time resolution [2]. However, measuring vibrational mode coupling at surfaces is challenging, because it requires both distinguishing the signal of a small number of surface molecules from a much larger bulk response and recording this signal within typical vibrational lifetimes (i.e. on sub-picosecond timescales). For bulk studies, femtosecond two-dimensional infrared (2D-IR) spectroscopy is ideally suited to reveal vibrational mode coupling. In 2D-IR, a vibrational mode A is excited, and the effect of this excitation on a different mode B is probed. If the modes are uncoupled, mode B remains unaffected by excitation of mode A, and a spectral response is only observed for mode A at identical pump and probe frequencies, i.e. on the diagonal of the 2D-IR spectrum. Inversely, the off-diagonal peaks between modes A and B are determined by the strength of their coupling and depend on their relative orientation and distance. As such, 2D-IR spectroscopy is increasingly useful in determining (sub-)molecular structures and dynamics; 2D-IR analogues of NMR experiments like NOESY, COSY and EXSY have been demonstrated [2-6]. Here we introduce femtosecond sumfrequency generation 2D-IR spectroscopy (SFG-2D-IR) with submonolayer sensitivity and surface specificity. Closely related surface 2D vibrational techniques have been recently proposed theoretically [7,8].
Experimental Method In bulk 2D-IR spectroscopy, a sequence of coherent interactions between the sample and the IR laser fields is designed such that an odd (typically third) order coherence is detected. To apply 2D-IR spectroscopy to surfaces, we gain monolayer sensitivity and interface specificity through an additional interaction with a nonresonant near-IR laser pulse. This additional interaction upconverts the third-order coherence to a fourth order coherence, which radiates a field in the visible, at the sum frequency of near-IR and IR. This upconversion process is beneficial in two ways. Firstly, it ensures surface specificity for materials whose optical response is dominated by dipole contributions.
511
Most materials are centrosymmetric, and the upconverted, even-order response can only originate from the surface molecules [9]. Secondly, the upconverted signal is background free and lies in the visible spectral range, where CCD cameras with high quantum efficiencies are available, readily providing sub-monolayer sensitivity. Specifically, we used a commercial Ti:Sapphire femtosecond amplified laser system to generate mid-IR pump and probe pulses. The pump pulse was shaped using a Fabry-Perot filter, resulting in a tunable narrowband ~15 µJ, 20 cm-1 pulse to excite specific vibrations. After a variable delay, the surface was probed by simultaneous mid-IR and near-IR pulses. The ~10 µJ IR probe pulse sustained 200 cm-1 bandwidth. To maintain spectral resolution in the upconversion step, a narrow-band (~12 cm-1) 800 nm pulse was used. The resulting sum frequency generation (SFG) spectrum reveals the effect of the pump pulse on all resonances within the probe bandwidth.
Results and Discussion We investigated a dodecanol monolayer on water – a model system for biological membranes – in the region of the C–H stretching modes [10]. The self-assembled monolayer was prepared by putting a small crystal of 1-dodecanol in contact with water. Fig. 1a and b show the IR spectrum of bulk dodecanol and the static 1D SFG spectrum of the monolayer. The C–H stretching region (Fig. 1a) features vibrational modes assigned to symmetric and antisymmetric CH2 and CH3 stretching (ss and as) perturbed by Fermi resonances (fr) with bending modes. For symmetry reasons, only some of these modes are SFG active and appear in the static SFG spectrum in Fig. 1b. Figs. 1c and d show SFG-2D-IR spectra of the self-assembled monolayer. In addition to the diagonal peaks, several off-diagonal peaks appear. As SFG selection rules apply for the probe process, off-diagonal peaks reporting on vibrational coupling appear at the two frequencies corresponding to the two peaks in the static SFG spectrum. The collective molecular alignment present at the interface (as opposed to the bulk) allows us to enhance the sensitivity to specific modes, by controlling the polarization of the pump laser pulse, as evident from a comparison of Figs. 1c and d. CH3(fr) CH3(as)
CH3(ss)
CH3(ss)
CH3(fr) CH3(as)
b c
-1
pump frequency [cm ]
a 2960 2920 2880 2840
d
CH3(as) CH3(fr) CH2(as) CH2(fr) CH3(ss) CH2(ss) p-pol
s-pol
2800 2840
2880
2920
2960 2840 2880 2920 -1 SFG probe frequency [cm ]
2960
Fig. 1. Fig. 1. SFG-2D-IR spectra of a dodecanol monolayer on D2O. Plain (shaded): pumpinduced decrease (increase) in SFG intensity. (a) IR spectrum of crystalline dodecanol at 150 K. (b) SFG spectrum of the monolayer, polarizations: SFG/VIS/IR: s/s/p. (c) SFG-2D-IR spectrum, p-pol. pump, t = 0.7 ps. (d) s-pol. pump, in-plane CH2 modes are efficiently excited, leading to the dominance in the 2D spectrum of the cross-peak between the CH2(as) and CH3(ss) modes. Lines: diagonal; arrows: off-diagonal peaks indicative of vibrational coupling.
512
When comparing the new surface 2D-IR technique to bulk 2D-IR spectroscopy, a few striking characteristics are apparent from the data: Firstly, the different selection rules for pump interactions (IR activity) and probe interactions (SFG activity requiring IR and Raman activity as well as broken centrosymmetry at the surface) make off-diagonal peaks appear that report on vibrational coupling between SFG inactive modes that are visible only in the IR absorption spectrum (shown along the pump axis) and SFG active modes (shown along the probe axis). Secondly, there is remarkably little increase in SFG-2D-IR intensity (positive red features in Figs. 1c and d). In third order bulk 2D-IR spectroscopy, an individual 2D-IR spectral response consists of a positive and a negative feature due to bleach, stimulated emission and excited state absorption. These paired features are largely suppressed in SFG-2D-IR spectra owing to the relative insensitivity of SFG to excited state transitions (the homodyne SFG signal is proportional to the square of the population difference between vibrational levels, heterodyne detection will allow to detect excited state dynamics as well). Thirdly, coherent interferences between the different modes play an important role in the SFG-2D-IR spectrum. These coherent interferences result from collective molecular alignment at the interface in combination with homodyne detection. Interestingly, the sign of off-diagonal peaks resulting from coherent interference depends on relative orientation of interfering oscillators, directly revealing structural information. Also interference terms could be suppressed by heterodyne detection.
Conclusions In summary, we have demonstrated the implementation of ultrafast surface 2D-IR spectroscopy. We expect this technique to be useful for a variety of applications, including the study of: the structure and reactivity of (mixed) molecular adsorbate layers in catalytic systems, the structures and interactions of membranes and membrane proteins as well as the structure and dynamics of interfacial water in various systems. T. Komeda, Y. Kim, M. Kawai, B. N. J. Persson, and H. Ueba, Science 295, 2055, 2002. R. M. Hochstrasser, Proc. Nat. Acad. Sci. USA 104, 14190, 2007. M. T. Zanni and R. M. Hochstrasser, Curr. Opin. Struct. Biol. 11, 516, 2001. M. L. Cowan, B. D. Bruner, N. Huse, J. R. Dwyer, B. Chugh, E. T. J. Nibbering, T. Elsaesser, and R. J. D. Miller, Nature 434, 199, 2005. 5 J. D. Eaves, J. J. Loparo, C. J. Fecko, S. T. Roberts, A. Tokmakoff, and P. L. Geissler, Proc. Nat. Acad. Sci. USA 102, 13019, 2005. 6 C. Kolano, J. Helbing, M. Kozinski, W. Sander, and P. Hamm, Nature 444, 469, 2006. 7 Y. Nagata, Y. Tanimura, and S. Mukamel, J. Chem. Phys. 126, 204703, 2007. 8 M. Cho, J. Chem. Phys. 112, 9978, 2000. 9 Y. R. Shen, Nature 337, 519, 1989. 10 J. Bredenbeck, A. Ghosh, M. Smits, and M. Bonn, J. Am. Chem. Soc. 130, 2152, 2008.
1 2 3 4
513
Frozen Dynamics and Insulation of Water at the Lipid Interface Artem A. Bakulin, Dan Cringus, Maxim S. Pshenichnikov, and Douwe A. Wiersma Zernike Institute for Advanced Materials, University of Groningen, Nijenborgh 4, 9747 AG Groningen, The Netherlands E-mail: [email protected] Abstract. 2D IR correlation spectroscopy reveals extremely slow dynamics and splitting of the OH-stretching mode of water in anionic micelles. Water at the lipid interface behaves as if the molecules were isolated in a “frozen” environment.
Introduction The properties of the dense and flexible 3D hydrogen bond network among water molecules are substantially altered in proximity of chemical or biological interfaces. At the same time, these properties play an important role in a wide range of processes like chemical reactions, protein secondary structure stabilization, proton and energy transfer, etc. The recent sum-frequency generation studies on water-silica [1] and water-air [2] interfaces have shown that despite noticeable differences in spectral responses of bulk and surface water, the OH stretch vibrational dynamics of the surface water molecules are hardly distinguishable from those in the bulk. Yet, similar experiments on water-lipid interfaces exposed a noticeable decrease of population relaxation rates at high frequencies [3]. This was attributed to the decelerated spectral diffusion that otherwise equalizes frequency-dependent relaxation rates. However, conclusions about dephasing (that is, the T2 time) were based on the OH-stretch vibration lifetimes (i.e., T1) while these two values are not necessarily interconnected. In this contribution, we present a 2D IR correlation spectroscopy study of the effect of the lipid interface on dynamics of the neighbouring water hydrogen-bond network. 2D correlation spectroscopy is capable of disclosing features otherwise hidden either beneath the broad absorption or in the broadband nonlinear response. As a model system for a membrane surface we used reverse micelles (fig. 1a) - the nanosize water droplets covered with a monolayer of lipid-like surfactant and floating in a nonpolar solvent. By varying the relative concentrations of H2O and surfactant (AOT), the amount of interface water can be accurately controlled. To probe to the dynamics of hydrogen bond network, the environmentally sensitive OH vibration of an H2O molecule was excited by a pair of 70-fs IR pulses, allowed to evolve, and finally probed by another pulse pair.
Results and Discussion The Figure 1b presents FTIR spectra of the OH stretch mode for 1 nm and 10 nm diameter H2O micelles, and for bulk H2O. The absorption spectrum for large micelles (d=10 nm diameters) follows the absorption spectrum of neat H2O because 90% of water molecules are displaced from the micelle interface and create a “bulk-like” core. In contrast, in the small micelles (d=1 nm) 90% of water molecules border the lipid membrane which leads to a clear blue shift of the OH stretch frequency indicating weaker hydrogen bonding.
514
Fig. 1. a) Schematics of the reverse micelle. b) Absorption spectra of bulk (grey shape), 10 nm and 1 nm reversed micelles in the OH stretch region. c) Normalized absorptive components of 2D correlation spectra (the x,x,x,x polarization geometry) of H2O confined in 10 nm diameter micelles at waiting times 0 ps, 0.1 ps, and 0.4 ps. The equilateral contours are drawn with the 8% step of the maximal amplitude.
Figure 1c shows the absorptive part of the 2D correlation spectra of H2O confined in d=10 nm micelles at different evolution times. A strong positive signal (red) along the diagonal represents ground state bleaching and stimulated emission at the |0>-|1> transition. The excited state |1>-|2> absorption is also visible as a negative (blue) signal shifted along the vertical axis for the anharmonicity value of ~200 cm-1. At waiting times longer than 0.1 ps the excited state absorption gradually disappears because of the ~0.2 ps population relaxation time of the stretch mode [4] and overlap with an additional bleaching induced by the overall heating of the sample. The diagonally-elongated shapes of the peaks are indicative of the inhomogeneous broadening of the absorption line. However, the antidiagonal width is also substantial showing significant dephasing during the coherence interval. At 0.1 ps inhomogeneity decreases dramatically while after 0.4 ps the 2D spectrum entirely loses any signs of correlation. Therefore, the water phase memory in the large micelles decays at ~0.1 ps timescale which is in a good agreement with the previous study on bulk H2O [4].
Fig. 2. Normalized experimental (a) and calculated (b) absorptive components of 2D correlation spectra (x,x,x,x polarization) of H2O confined in d=1 nm micelles at waiting times 0.1, 0.3, 0.6, 1.2 and 5 ps. (c) Correlation functions derived from the 2D IR spectra for d=1 nm (circles) and d=10 nm (triangles) micelles. Solid curves are exponential fits to data.
Figure 2a depicts absorptive 2D correlation spectra of H2O confined in the d=1 nm micelles. Similarly to Fig.1c, the spectra display the on-diagonal |0>-|1> bleaching peak and |1>-|2> induced absorption shifted by the anharmonicity value. However, in a striking contrast with the d=10 nm case, the correlation spectra are considerably narrower at short waiting times and retain the elliptical shape for much longer evolution times. Fortunately, even at waiting times longer than 1 ps the thermal response remains spectrally separated from the population peak which allows tracking the latter up to 5 and even 10 ps (not shown). The diagonal peak shapes observed in
515
the 2D spectra directly demonstrate the still-existing correlation between the excitation and probing. This implies that at 10 ps time scale there are no dynamical processes which wash out the structural variations in the hydrogen bond network and thereby scramble the OH bond stretching frequency.
Fig. 3. Absorptive components of the 2D spectra for the d=1 nm micelles obtained in the crosspolarization geometry (x,x,y,y). Waiting times are 0 ps, 0.1 ps, 0.3 ps and 1.2 ps.
The results of the eccentricity analysis [5] are presented in Fig.2c by symbols. Consistently with previous quantitative estimates, the correlation function for the large water droplets decays with the ~0.15 ps. In contrast, the correlation function for the small micelles, after a small decrease in the first 0.3 ps, levels off and stays invariable on a few ps timescale. With this correlation function, the 2D spectra are successfully modelled as a simple combination of the population (correlated) and thermal (uncorrelated) responses (Fig.2b). The long phase memory in the small micelles signifies frozen dynamics of the OH stretch vibrations in the interface-bounded water. This also shows that there are no intermolecular communications amongst the interface H2O molecules nor between the interface and bulk water because otherwise the correlation would be wiped out by intermolecular energy transfer [4]. The correlation spectra measured in the cross-polarization geometry (Fig.3) show a noticeable off-diagonal peak at ω1=3450, ω3=3540 cm-1 (the second symmetric peak is shadowed by the excited state absorption). The amplitude of this peak is non-zero at early evolution times which signifies a common ground state of both transitions. Furthermore, its amplitude rises with a ~200 fs time constant until it equalizes with the amplitude of the diagonal 3540 cm-1 peak. Such behavior is typical for a pair of coupled dipoles with orthogonal directions of the dipole moments such as the H2O symmetric and asymmetric stretch modes. The splitting originates from isolation of the water molecules near the interface and their double bonding to the membrane.
Conclusions We directly demonstrate that the phase memory of water near the lipid interface is retained for over 10’s of ps which is a factor of 100 slower than in bulk [4]. The observed splitting of the interfacial H2O stretch onto the symmetric and asymmetric modes signifies the weakening of the hydrogen bond network and dominant bonding of the water molecules to the lipid membrane. 1 2 3 4 5
516
J.A. McGuire, Y.R. Shen, Science 313, 1945, 2006. M. Smits, A. Ghosh, M. Sterrer, M. Muller, M. Bonn, Phys. Rev. Lett. 98 98302, 2007. A. Ghosh, M. Smits, J. Bredenbeck, M. Bonn, J. Am. Chem. Soc. 129, 9608, 2007. M.L. Cowan, B.D. Bruner, N. Huse, J.R. Dwyer, B. Chugh, E.T.J. Nibbering, T. Elsaesser, R. J. D. Miller, Nature 434, 199, 2005. K. Lazonder, M.S. Pshenichnikov, D.A. Wiersma, Optics Letters 31, 3354, 2006.
Ultrafast vibrational dynamics of interfacial water Avishek Ghosh1,2, Richard K. Campen1, Maria Sovago1, Jens Bredenbeck1 and Mischa Bonn1,2 1
FOM-Institute for Atomic and Molecular Physics AMOLF, Kruislaan 407, 1098 SJ Amsterdam, The Netherlands; E-mail: [email protected] 2 Chemistry Dept., Leiden University, P.O. Box 9502, 2300 RA Leiden, The Netherlands Abstract. We report investigations on the ultrafast vibrational dynamics of water molecules at model biological interfaces and the neat water/air interface, using a newly developed surfacespecific 4th-order femtosecond infrared pump-probe spectroscopic technique. The vibrational relaxation rates and mechanisms depend strongly on the nature of the interface. Whereas water at the neat water/air interface exchanges vibrational energy rapidly with the bulk, the water molecules at model biological interfaces are energetically decoupled from the bulk.
Introduction The interaction between water and lipid headgroups is very important in biology [1-3]. It is challenging, however, to investigate and characterize the structure and dynamics of the ~1 molecular layer of water interacting with the lipid headgroups. Vibrational sum frequency generation (SFG) has enabled the investigation of the vibrational spectrum of ~1 monolayer of water molecules directly interacting with the lipids, owing to its unique selection rules. In this spectroscopic technique, infrared light pulses resonant with the O-H stretch vibration are overlapped with visible pulses at the interface, which results in the generation of light with a frequency equal to the sum of the two incident frequencies. The nonlinear optical process of SFG is forbidden in centrosymmetric media but allowed wherever inversion symmetry is broken, for instance at interfaces, making this technique highly surface-sensitive. The static SFG spectrum of interfacial water in the O-H stretch frequency range has been intensely investigated the past decades. Similar to the water-air interface [4], the static SFG spectrum of the water-lipid interface [5] is characterized by two broad peaks in the hydrogen-bonded region between 3100 and 3500 cm-1, thus masking the true dynamics of interfacial, lipid-bound water molecules. Questions arise pertaining to the heterogeneity of water molecules and the timescales of exchange between possible subensembles. To address these issues, we have recently developed the technique of femtosecond time-resolved sum frequency generation spectroscopy (TR-SFG) that allows us to study interfacial water dynamics directly [6].
Experimental Methods In the TR-SFG experiments, the O-H stretch vibration of water molecules is excited by an intense infrared pump pulse. The vibrational relaxation of these excited O-H vibrations of interfacial water molecules is then monitored in real-time using an SFG probing scheme (see figure 1). The fact that real membranes are chemically heterogeneous with respect to lipid composition, has motivated us to perform time-
517
VIS
IR probe
IR pump
A lipids
B SFG
VIS
vO-H = 1
water
SFG
τpu-pr
vO-H = 0
Figure 1 (A) Schematic for the time-resolved pump-probe SFG technique applied to a waterlipid interface. (B) Energy levels involved in the IR pump-SFG probe technique. The SFG response is monitored as a function of the delay between the IR pump pulse and the SFG probe pulse pair (IR and visible).
resolved IR-pump-SFG-probe studies on water molecules at a variety of lipid interfaces in which the lipid head-group structure differs. TR-SFG experiments were performed on Langmuir monolayer films of 1,2-Dimyristoyl-Glycero-3-Phospho-LSerine (DMPS, Sodium salt, net negatively-charged headgroup), 1,2-Dipalmitoyl-3Trimethylammonium-Propane (DPTAP, Chloride salt, net positively-charged headgroup), 1,2-Dipalmitoyl-sn-Glycero-3-Phosphocholine (DPPC) and 1,2Dipalmitoyl-sn-Glycero-3-Phosphoethanolamine (DPPE, both zwitterionic headgroups), prepared on ultrapure Millipore water subphases.
Results and Discussion Using tr-SFG, it has been shown that water molecules at the neat water surface [6] and water-silica interfaces [7], exchange vibrational energy with the underlying bulk water molecules on sub-100 fs timescales. However, in the tr-SFG transients for water-DMPS no such ultrafast energy exchange with the bulk was observed. In fact, we have shown that the water molecules at this water/lipid interface are energetically decoupled from the bulk and that vibrational relaxation proceeds by coupling of the O-H oscillator to the hydrogen-bond mode [5]. In figure 2, a comparison of the vibrational dynamics at the air-water and lipid-water interfaces is shown. Water at the DPTAP interface exhibits similar dynamics to water at DMPS, indicating that the water behavior for charged lipids is independent of the type and the total number of charges on the headgroup. For both charged lipids, the vibrational relaxation process can be explained by efficient energy flow from the excited O-H oscillators into the hydrogen-bond mode: we find no evidence for significant intramolecular energy flow. Interestingly, the dynamics of water at DPPC and DPPE show several spectral features that suggest additional modes participate in vibrational relaxation. These observations are consistent with water molecules strongly interacting with lipid functional groups in the DMPS/DPTAP case and enjoying more structural flexibility in the case of DPPC/DPPE.
518
Normalized differential SFG
1.6 1.5
A
B
T1 = 190 fs, T therm = 500 fs
-1
-1
3500 cm , T1 = 570 fs
-1
3430 cm , T1 = 430 fs
3500 cm
1.4 1.3
-1
3430 cm
-1
1.2 1.1
3300 cm
-1
3200 cm
-1
1.0 0 500 1000 1500 pump-probe delay (fs)
3300 cm , T1 = 130 fs
-1
3200 cm , T1 < 100 fs
0 500 1000 1500 pump-probe delay (fs)
Figure 2. Pump-probe SFG transients for (A) the neat water-air interface and (B) the waterDMPS interface. The infrared-infrared-visible SFG crosscorrelation trace (lower left) determines the time-zero and illustrates the time resolution of the experiment. All traces are offset from 1.0 for clarity.
Conclusions In summary, we have investigated the local hydrogen-bonding environment of water surrounding a variety of lipid headgroups, using a novel time-resolved spectroscopic technique to probe water interfacial vibrational dynamics. Future work is aimed at answering some long-standing questions concerning the molecular description of lipid-water interaction at biological membranes. Acknowledgements. This work is part of the research program of the Stichting FOM with financial support from NWO. 1 Mulkidjanian, A. Y.; Heberle, J.; Cherepanov, D. A. Biochim. Biophys. ActaBioenergetics 2006, 1757, 913. 2 Poolman, B.; Spitzer, J. J.; Wood, J. A. Biochim. Biophys. Acta-Biomembranes 2004, 1666, 88. 3 Freites, J. A.; Tobias, D. J.; von Heijne, G.; White, S. H. Proc. Natl. Acad. Sci.USA 2005, 102, 15059. 4 Du, Q.; Superfine, R.; Freysz, E.; Shen, Y. R. Phys. Rev. Lett. 1993, 70, 2313. 5 Ghosh, A.; Smits, M.; Bredenbeck, J.; Bonn, M. J. Am. Chem. Soc. 2007, 129, 9608. 6 Smits, M.; Ghosh, A.; Sterrer, M.; Muller, M.; Bonn, M. Phys. Rev. Lett. 2007, 98, 098302. 7 McGuire, J. A.; Shen, Y. R. Science 2006, 313, 1945.
519
Ultrafast Dynamics at Liquid Interfaces Investigated with Femtosecond Time-Resolved Multiplex Electronic Sum-Frequency Generation (TR-ESFG) Spectroscopy Kentaro Sekiguchi, Shoichi Yamaguchi, and Tahei Tahara Molecular Spectroscopy Laboratory RIKEN (The Institute of Physical and Chemical Research), 2-1 Hirosawa, Wako, Saitama, 351-0198, Japan E-mail: [email protected] Abstract. We developed a new nonlinear spectroscopy, femtosecond time-resolved electronic sum-frequency generation (TR-ESFG) spectroscopy, to investigate ultrafast dynamics at liquid interfaces. Transient electronic spectra at the air/water interface were obtained for the first time.
Introduction Investigation of molecular properties at liquid interfaces, especially at aqueous solution interfaces, has been an important issue that has fundamental significance in chemical kinetics, biophysics, and atmospheric chemistry. Even-order nonlinear optical spectroscopy allows us to selectively obtain signals from molecules at interfaces. Especially for ultrafast dynamics of excited molecules, time-resolved second harmonic generation (TR-SHG) spectroscopy has been utilized to monitor interesting phenomena that happen only at the interfaces. In this study, we developed femtosecond time-resolved electronic sum-frequency generation (TR-ESFG) spectroscopy [1] based on the steady-state ESFG method that we have developed earlier [2,3]. The advantage of TR-ESFG spectroscopy over conventional TR-SHG is that it provides spectral information. We examined the ultrafast dynamics of a surface active dye, Rhodamine 800 (R800), at the air/water interface and successfully measured time-resolved interface-selective electronic spectra for the first time.
Experimental Methods The energy diagram and experimental configuration are schematically shown in Fig. 1. A femtosecond Ti:sapphire regenerative amplifier was used as the light source, and its output (800 nm, 1 kHz, 1 mJ) was divided into three. The first part was used for the excitation of an optical parametric amplifier to generate a signal output at 1380 nm. It was frequency-doubled to 690 nm, and used as the pump pulse, ω p . The second part of the regenerative amplifier output was used as the narrow-band ω1 probe pulse. The third part was focused into water to generate a white light continuum, which was used as the broad-band ω2 probe pulse. After the excitation of molecules with the pump pulse, sum-frequency signals generated by the two probe pulses were introduced to a single polychromator and detected by a CCD. Obtained spectra were normalized by the quartz standard spectrum to correct the spectral distortion due to the intensity distribution of the ω2 pulse. The spectra after this correction are called the “ESFG spectra” that correspond to |χ (2) |2 spectra. 520
Fig. 1. (a) Energy diagram and (b) experimental configurations for the TR-ESFG spectroscopy.
The ESFG spectrum was measured with and without the pump pulse alternately, and the difference spectrum (∆|χ (2) |2 spectrum) was obtained by subtracting the steadystate ESFG spectrum (which was observed without the pump irradiation) from the spectrum measured with the pump pulse. The time-resolution of TR-ESFG measurements was about 400 fs (FWHM). The ω p pulse was circularly polarized and does not induce in-plane anisotropy. The ω1 and ω2 pulses were linearly polarized and set at p-polarization. The pulse energies of ω p , ω1 , ω2 pulses were typically 7.5 µJ, 10 µJ, and 7.5 µJ, respectively.
Results and Discussion The obtained TR-ESFG spectra of R800 at the air/water interface are shown in Fig. 2. In addition to the negative signals due to the ground state bleaching, positive signals with a finite rise time were clearly observed. We carried out a global fitting analysis using exponential functions and found that all the temporal change can be represented with three time constants 0.32 ps, 6.4 ps, and 0.85 ns. By comparing the result with that obtained by transient absorption measurements for R800 in bulk water [4], we ascribed the 0.32-ps component to the lifetime of R800 dimers in the lowest excited singlet (S1 ) state. They dissociate and generate monomers 521
Fig. 2. TR-ESFG signals of R800 at the air/water interface. The time delays between the pump (ω p ) and probe (ω1 and ω2 ) pulses are shown at the side of each spectrum. The positive and negative signals are represented with the area painted grey and black, respectively. Expanded signals are also shown in the wavelength region of 425 – 460 nm. (ω p , ω1 , and ω2 were 690 nm, 800 nm, and 540 – 1200 nm, respectively.)
in the S1 state, and 0.85 ns dynamics is ascribed to the lifetime of the S1 monomer. The 6.4-ps component is attributed to an interface-specific deactivation process of the S1 monomers, because the corresponding dynamics is missing in the bulk. The present work demonstrated that TR-ESFG spectroscopy allows us to investigate ultrafast dynamics at liquid interfaces as thoroughly as we do for the dynamics in bulk solutions with conventional transient absorption spectroscopy. Acknowledgements. K.S. acknowledges the Special Postdoctoral Researchers Program of RlKEN. S.Y. acknowledges a Grant-in-Aid for Young Scientists (B) (No. 15750023) from the Ministry of Education, Culture, Sports, Science and Technology (MEXT) of Japan. This work is supported by a Grant-in-Aid for Scientific Research on Priority Areas “Molecular Science for Supra Functional Systems” (No. 19056009) from MEXT of Japan and a Grant-in-Aid for Scientific Research (A) (No. 19205005) from Japan Society for the Promotion of Science (JSPS). 1 2 3 4
K. Sekiguchi, S. Yamaguchi, and T. Tahara, J. Chern. Phys. 128, 114715 (2008). S. Yamaguchi, and T. Tahara, J. Phys. Chern. B 108, 19079 (2004). Yamaguchi, and T. Tahara, J. Chern. Phys. 125, 194711 (2006). Sekiguchi, and S. Yamaguchi, and T. Tahara, J. Phys. Chern. A 110, 2601 (2006).
522
Femtosecond spectral phase shaping for CARS spectroscopy and imaging Sytse Postma, Alexander C. W. van Rhijn, Jeroen P. Korterik, Jennifer L. Herek, and Herman L. Offerhaus Optical Sciences Group, Department of Science and Technology, MESA+ Institute for Nanotechnology, University of Twente, P.O. Box 217, 7500 AE Enschede, The Netherlands E-mail: [email protected] Abstract. Coherent Anti-Stokes Raman Scattering (CARS) is a third-order non-linear optical process that provides label-free, chemically selective microscopy by probing the internal vibrational structure of molecules. Due to the resonant enhancement of the CARS process, faster imaging is possible compared to Raman microscopy. CARS is unaffected by background fluorescence, but the inherent non-resonant background signal can overwhelm the resonant signal. We demonstrate how simple phase shapes on the pump (and probe) beam reduce the background signal and enhance the resonant signal. We demonstrate chemically selective microscopy using these shaped pulses on plastic beads.
Introduction Coherent anti-Stokes Raman scattering (CARS) has been used successfully in spectroscopy and microscopy since the development of (tunable) pulsed laser sources. In resonant CARS, molecular vibrations are coherently excited by a pump (ωp) and Stokes (ωs) pulse. Subsequently a probe (ωpr) pulse, which is often similar to the pump pulse, generates the anti-Stokes signal (ωc = ωp - ωs + ωpr). The resonant CARS signal is accompanied by an inherent non-resonant background. Here we demonstrate how spectral phase shaping strategies can amplify the resonant features in the spectrum to such an extent that spectroscopy and microscopy can be done at high spectral resolution, even on the integrated spectral response [1]. We use this technique for chemical selective imaging of polystyrene (PS) and polymethylmethacrylate (PMMA)
Setup We use a tunable Ti:Sapphire oscillator with a FWHM of 20 nm (80 MHz repetition rate). The liquid crystal device (LCD) of the reflective spectral phase shaper has 4096 pixels with a pixel size of 1 µm by 6 mm and a pitch of 1.8 µm. Effectively the spectral phase shaper has ~600 degrees of freedom for pattering. For the absolute positioning of phase profiles the complete number of pixels can be used, which implies a positioning precision of 14 GHz (0.5 cm-1). Further details of the spectral shaper setup can be found in an earlier publication [2]. The shaped Ti:Sapphire pulses are used as the pump and probe pulses in the CARS process. The Stokes pulse is generated by a 15 ps (1 cm-1) Nd:YVO laser. A reflective objective of 0.65 NA is used to focus the light on the sample. The collection objective is a 0.65 NA regular glass objective. The collected light is detected by a spectrometer or a photomultiplier tube. 523
Spectroscopy Our spectroscopic method is based on sweeping a π-phase step through the spectrum of the broadband (pump and probe) pulse and recording the CARS spectrum for each position of the step [1]. In this 2D plot the signal from vibrational resonances and the signal from the non-resonant background can be easily identified as distinct features. The difference between the positive and negative π-phase step rejects purely nonresonant features. Figure 1 shows the spectra for (a) a positive step sweep, (b) a negative step sweep, (c) the difference between (a) and (b) and (d) the integrated CARS signals. The horizontal axes represent the frequency of the phase step and the vertical axes represent the CARS spectra or the integrated signal. The intensities in the figures have been normalized to the unshaped CARS intensity. The strong resonance is clearly identified, the weaker resonances are lost in the noise.
Fig. 1. CARS spectra for a π-phase step sweep on acetone. a) Positive π-phase step sweep. b) Negative π-phase step sweep. c) The difference between the positive and negative sweeps. d) The integrated signal for the three cases.
The imaging of the plastic beads is based on the integrated difference signal. Figure 1(d) shows that, for a slightly red shifted frequency for the phase step in comparison with the location of the vibrational resonance, the difference signal is positive. For blue shifted frequencies the difference signal is negative. In a sample of mixed 4 μm PS and PMMA beads dried on a glass substrate, an ordinary transmission image can not resolve the different types of beads. Figure 2 shows integrated difference between the CARS signal for a positive and a negative phase profile. Figure 2(a) shows the CARS image for transform limited pulses (flat phase profile). Figure 2(b-d) show difference CARS images, with (b) the difference of two π-phase steps at 372.7 THz, (c) the difference of two π-phase steps at 369.0 THz, and (d) the difference of two phase structures with two phase steps. The images have been normalized by the maximum CARS signal for the transform limited pulse.
524
Fig. 2. Spectral phase profiles applied to the pump and probe pulses, which result in chemical contrast as a result of the applied phase profile. a) flat phase profile. b) PS enhanced phase profile. c) PMMA enhanced phase profile. d) PMMA enhanced and PS decreased phase profile.
In this particular case the transform limited CARS image also shows a difference between the PS and PMMA beads, because PS has a larger cross-section than PMMA for the chosen spectra of the pump and probe and Stokes. Figure 2(b) shows a CARS image for an applied phase profile that enhances the resonant signal of the PS beads and it results in a net negative result for the PMMA beads. Figure 2(c) shows a CARS image for an applied phase profile that enhances the resonant signal of the PMMA beads, which results in a similar net result for the PS beads. Figure 2(d) shows the same image for an applied phase profile that enhances the PMMA beads and at the same time results in a net negative result for the PS beads. The subtraction of opposite phase profiles results in a removal of all pure nonresonant contributions. The concurrent phase step in the probe pulse results in less (resonant) CARS signal, which reduces the signal to noise ratio. The low signal to noise ratio is especially visible in the case for the double step image.
Conclusions We demonstrate a method for chemically selective imaging based on detection of the integrated CARS signal, by applying simple phase profiles to the pump and probe pulses of the CARS process. For the future we are planning to expand these techniques for use in the fingerprint region for biological purposes. [1] S. Postma, A. C. W. van Rhijn, J. P. Korterik, P. Gross, J. L. Herek, and H. L. Offerhaus, Opt. Express 16, 7985 (2008). [2] S. Postma, P. van der Walle, H. L. Offerhaus, and N.F. van Hulst, Rev. Sci. Instrum. 76, 123105 (2005).
525
Part VII
Biological Systems, Molecular Light Harvesting and Charge-Transfer Complexes
Energy transfer along a poly(Pro) - peptide Wolfgang Zinth1, Wolfgang J. Schreier1, Tobias E. Schrader1, Florian O. Koller1, Markus Löweneck3, Hans-Jürgen Musiol2 and Luis Moroder2 1
LS für BioMolekulare Optik, LMU München, Oettingenstr. 67, D-80538 Munich, Germany Max-Planck Institut für Biochemie, Am Klopferplatz 18 A, D-82152 Martinsried, Germany 3 Senn Chemicals AG, Guido Senn Strasse 1,CH-8157 Dielsdorf, Switzerland E-mail: [email protected] 2
Abstract. Using a novel molecular thermometer, p-nitro-phenylalanine, we investigate the transport of vibrational excess energy along a poly(Pro) sequence. Time resolved IRspectroscopy reveals that heat transfer proceeds at a speed of several Å per picosecond.
Introduction Many biological processes rely on the directed supply of energy to the molecular reaction sites. In this respect the speed of the energy transport, its efficiency, pathways of preferred energy flow and even solitonic energy transport are widely discussed. A number of experiments has shown, that vibrational excess energy released in a chromophore in solution is dissipated within about 10 ps to the surrounding solvent. Similar times for vibrational cooling have been found for chromopeptides or chromoproteins. Very little is known about heat diffusion over longer distances in proteins. Recently the transport of vibrational energy (heat flow) released after photoexcitation/photoisomerization of an azobenzene was studied in a 310-helix, covalently attached to the chromophore [1]. Isotopic labelling was used to monitor the heat flow. In the present contribution we apply an alternative approach to study the transport of excess energy in a peptide. We introduce a modified amino acid containing a NO2-group as a local sensor (molecular thermometer) for vibrational excess population or local temperature. We study the heat transport along a rigid poly-Pro sequence with time resolved IR-spectroscopy on the picosecond time range and discuss the results using a model for directed and isotropic heat transport.
Materials and Methods The azobenzene peptides, (Ac-AMPB-(Pro)n-Phe(NO2)-(Pro)7-n-Asp-NH2, with n = 1, 5 or 7) have been synthesized following analogous procedures as reported for other azobenzene peptides [3] and are dissolved in DMSO-d6. The amino acid containing the sensor - a p-nitro-phenylalanine, Phe(NO2) - is placed at different positions (see Fig. 1) along the poly(Pro) sequence. The strong oscillatory strength of the NO2stretching vibrations and the clear separation from other bands in the peptide allow sensitive detection of the local temperature change. It has been shown recently that the NO2-group is well coupled to the peptide backbone: experiments on FmocPhe(NO2) indicate that vibrational excitation of the backbone CO vibration is coupled to the reporting NO2-group within 2 ps. Transient absorption experiment are performed by the pump-probe technique using a Ti-sapphire laser system (operated at 1 kHz) with excitation pulses at 405 nm in the nπ*-band of the AMPB-chromophore. Probing in the mid-IR is performed as
529
described recently [2]. After photo-excitation and ultrafast internal conversion the AMPB-moiety acts as a heat source, that is covalently linked with the peptide chain.
Fig. 1. Schematic representation of the energy-marker molecule (a) and the AMPB-peptide (b). Three sequences have been synthesized with the Phe(NO2)-amino acid incorporated at the indicated positions along the peptide chain. (c) Infrared absorption spectra of a peptide with the marker at Position 2 and a corresponding peptide missing the NO2-group.
Results and Discussion In Fig. 2 we show the transient absorption data recorded in the range of the symmetric NO2-mode for three positions of Phe(NO2). When the marker molecule Phe(NO2) or its directly coupled surrounding undergoes a temperature rise, the coupling of the NO2-stretching modes to thermally excited low frequency vibrations changes the NO2 absorption band. From steady state experiments performed at different temperatures we know that the band is red-shifted resulting in an absorption decrease at the original band position and an increase at lower frequencies. This explains the signature (absorption decrease) observed in Fig. 2 taken nearby the peak of the original NO2 absorption band. When the Phe(NO2) is closer to the AMPBchromophore (Pos. 2), the signal rise occurs earlier than for more distant locations. As illustrated in Fig. 2 the maximum excess heat arrives at position 2 after about 3.5 ps. At positions 6 and 8 the excess heat is further delayed by about 2.5 and 3.5 ps respectively. At later times the signal decays due to energy transfer along the residual peptide and to the surrounding solvent on a time scale of 10 – 20 ps. In addition it can be seen, that the measured absorption change corresponding to the maximum temperature increase at the Phe(NO2) location is smaller for Pos. 6 and 8. At very late times the samples show similar absorption changes as expected for a thermal equilibrium where most of the excess energy resides in the surrounding solvent. A qualitative model of the heat transport would suggest, that the excess energy is released from the chromophore by isomerization and internal conversion to the attached peptide and to the surrounding solvent. This process happens on the 1 ps time scale. We assume that the energy is essentially thermalized within the vibrational system and that the observed absorption changes originate from the population of low lying frequency modes coupling anharmonically to the monitor bands. With increasing distances from the heat source (e.g. along the peptide) the heat wave arrives more and more delayed. Simultaneously the peak temperature at the sensor will decrease for larger distances since the heating energy is distributed over a larger volume. From the results shown in figure 2 one may estimate the speed of the heat transfer from Pos. 2 to Pos. 6 to be in the order of a few Å per picosecond.
530
Fig. 2. Comparison of the transient absorption data recorded in the range of the symmetric NO2-mode for three positions of the Phe(NO2)-molecule in the peptide chain. The data were corrected for the induced absorption caused by neighboring hot proline.
To model the heat transport process we assume that azobenzene acts as an ultrafast local heat source and consider two situations: (i) In a 3-d heat diffusion model the azobenzene heat source feeds the excess energy isotropically into its surroundings (solvent and peptide). There is a fast distribution of the excess energy over a considerable volume and the peak temperature drops rapidly with distance from the heat source. For large distances from the heat source, the simulation does not fit to the observations. (ii) In a 1d-model heat diffuses predominantly along the peptide chain (we assume a rod-like structure and a heat conductivity as found in water), before it is conducted to the surrounding solvent. This model qualitatively describes the experimental observations. Along the peptide the decrease of peak temperature is slower as compared to the 3d-case. Within this model it is not surprising that the excess heat has been observed as far away from the source as 25 Å.
Conclusions Heat transfer in molecular systems of biological relevance can occur within few picoseconds over distances of several nanometers. The experimental observations together with simulations within a model of 1d-heat conduction along the poly(Pro) peptides suggest that heat transfer may occur preferentially along defined molecular structures. Acknowledgements. We would like to thank the German Science Foundation DFG for supporting the studies via SFB 533 and SFB 749. 1 2 3
V. Botan, E. H. G. Backus, R. Pfister, A. Moretto, M. Crisma, C. Toniolo, P. H. Nguyen, G. Stock, and P. Hamm in Proceedings of the National Academy of Sciences U. S. A. 104, 12749, 2007. T. Schrader, A. Sieg, F. Koller, W. Schreier, Q. An, W. Zinth, and P. Gilch in Chemical Physics Letters 392, 358, 2004. C. Renner, J. Cramer, R. Behrendt, and L. Moroder in Biopolymers 54, 501, 2000.
531
Energy transport in peptide helices around the glass transition Ellen H.G. Backus1, Phuong H. Nguyen2, Virgiliu Botan1, Rolf Pfister1, Alessandro Moretto3, Marco Crisma3, Claudio Toniolo3, Gerhard Stock2, and Peter Hamm1 1
Physikalisch-Chemisches Institut, Universität Zürich, Winterthurerstrasse 190, CH-8057 Zürich, Switzerland E-mail: [email protected] 2 Institut für Physikalische und Theoretische Chemie, J.W. Goethe Universität, Max-von-LaueStrasse 7, D-60438 Frankfurt, Germany 3 Institute of Biomolecular Chemistry, Padova Unit, CNR, Department of Chemistry, University of Padova, via Marzola 1, I-35131 Padova, Italy Abstract. The energy transport through a small helical peptide has been studied as
function of temperature. Diffusive transport dominates at high temperature, while ballistic transport seems to be important at low temperature.
Introduction Proteins are molecular machines which need energy to function, but also only function in narrow temperature ranges. Apparently, Nature has invented a way to transport efficiently energy to and from the active site of a protein. To study part of this process, we have investigated energy transport in a small helix, a dominant structural element of a protein. Our model system consists of 8 amino acids (one of them labelled with 13 C=O) attached to an azobenzene photoswitch, which is used to deposit energy in the molecule (Fig. 1 top). Heat in the molecule is detected with infrared spectroscopy, making use of the shift of vibrational bands upon local heating [1]. The isotope labelled amino acid is either the second or the fourth residue counted from the N-terminal azobenzene, resulting in the molecules Aib16 and Aib34. Isotope labelling shifts the frequency of the Amide I band to lower frequency, well-separated from the other Amide I bands, resulting in a locally specific thermometer (Fig. 1 middle). Combining this isotope labelling with femtosecond pump-probe spectroscopy, spatial and temporal resolution can be obtained.
Results and Discussion The bottom panel of Fig. 1 depicts pump-probe data at four different times after excitation of the azobenzene group at two different temperatures. At time zero the azobenzene moiety isomerizes from cis to trans depositing a huge amount of heat (estimated temperature close to 1000 K) in the molecule. Clearly, band 1 and band 3 respond immediately to the heat in the azobenzene, because already a signal is observed at time zero. For band 5 a signal is observed only after 3 ps at high temperature. In Fig. 2 (top) the bleach intensity for bands 1, 3, and 5 is plotted as a function of time for 220 and 303 K. At 303 K the maxima of band 1, 3, and 5 are delayed with respect to each other, showing the propagation of heat through the molecule. We can describe the curves very well with a diffusive model depicted in the inset. In this
532
model [2] we need two time constants, one for the heat transport between neighbouring C=O groups (1.5 ps) and one for cooling to the solvent (7 ps). In contrast, at 220 K the diffusive model does no longer work. The intensity of band 3 can be modelled correctly by reducing the propagation rate, but then the maximum comes too late (inset of Fig. 2). Apparently, the energy transport is no longer diffusive.
1
0.04
main
243 K 1 9
3/5
303 K 1 9
3/5
0.08 0.04
0.00 0.00 1600 1640 1680 1720 1760 1600 1640 1680 1720 1760 -1
frequency (cm ) 0.2
Aib16
220 K 9
0.0 -0.2 -0.4 -0.6 0.2
1 213 K 9
0.0 -0.2 -0.4
5
main Aib34
0.0 -0.4
1
-0.6
303 K 9
0.2 0.0 -0.2
5 1
0.2 -0.2
3
3 main Aib34
303 K 9
Aib16
main
1
-0.4
-0.6 -0.6 main 1600 1640 1680 1720 1760 1600 1640 1680 1720 1760
absorbance change (mOD)
absorbance change (mOD)
main
0.08
9
5
absorbance (OD)
absorbance (OD)
3
-1
frequency (cm )
Fig. 1. Top) X-ray diffraction structure of the backbone of the molecule showing the azobenzene, the ester connection group, the helix with the 8 residues and the OMe endgroup. Middle) FTIR spectra of Aib16 at 243 and 303 K (‘main’ means all C=O groups except number 1, 9, and the labelled one). Bottom) Transient pump-probe spectra of Aib16 and Aib34 at two different temperatures at 1, 3, 11 and 40 ps after excitation of the azobenzene group with 420 nm. The numbers 1 to 9 refer to C=O groups in the molecule counted from the azobenzene.
To get a model independent measurement for the energy transport, we plot in the bottom right panel of Fig. 2 the ratio of the increase of band 3 (from t=0 to t=3 ps) and the intensity of band 1 (t=0) for six different temperatures. This plot suggests that the heat transport is more or less constant from 220 to 270 K after which it suddenly increases with temperature. Also the vibrational frequency of the main band shows a discontinuity around this temperature as shown in the bottom left panel. Below and above 280 K the vibrational frequency shift linear as a function of temperature, but the slope is different in these two temperature regimes. With NMR spectroscopy we see a similar effect on the chemical shift of the NH protons. In combination with MD simulations we conclude that the molecule is less flexible at low temperature and becomes more flexible around 270 K. Similar sudden changes in the behaviour of
533
220 K
0.6
1
0.4
0.10
1
0.05
3
0.00 0
0.2
5
10
0
5
CO1 ks
kp CO2 ks
CO3 ks
303 K
3
15
3 5
0.0
kp PAZ ks
5 10
15
20
time (ps)
0
-1
0.2
10
15
20
time (ps) 0.16 0.14
main
0.12
1664
0.10 0.08
1662 220 240 260 280 300 320
220 240 260 280 300 320
temperature (K)
temperature (K)
ratio (a.u.)
frequency (cm )
0.4
0.0 5
1668 1666
0.6
bleach intensity (mOD)
bleach intensity (mOD)
molecules have been observed with for example x-ray crystallography and infrared spectroscopy for proteins in aqueous solution around ~200 K, the so called glass transition [3].
Fig. 2. Top) Time dependence of the bleach intensity (opposite sign as Fig. 1) for C=O 1, 3, and 5 as a function of time at 220 and 303 K. Bottom) Frequency of the main band (left) and ratio of the increase of band 3 and the intensity of band 1 (right) as a function of temperature.
Conclusions Low-frequency vibrational modes are delocalized over large parts of (bio)polymers, while high-frequency modes are not [4]. Therefore, vibrational energy in these delocalized low-frequency modes dominates the energy transport. After excitation, part of the energy is in this type of modes and will be transported through the chain in a ballistic-like manner. The part in higher frequency modes can participate in the transport after it relaxes to low-frequency modes. At low temperatures, the molecule is rigid, and the vibrational energy re-distribution is inefficient. As the temperature increases above the glass transition, the molecule is flexible enough to get energy out of higher-frequency modes into the low-frequency modes. This re-feeding makes the transport more efficient and diffusive-like, in the sense that it fits a rate equation model. Apparently, the energy transport properties of molecular devices can be regulated by engineering the flexibility of a molecule. 1 2 3 4
534
P. Hamm, S.M. Ohline, and W. Zinth, J. Chem. Phys. 106, 519 (1997). V. Botan, E.H.G. Backus, R. Pfister, A. Moretto, M. Crisma, C. Toniolo, P.H. Nguyen, G. Stock, and P. Hamm, Proc. Natl. Acad. Sci. U.S.A. 104, 12749 (2007). D. Ringe and G.A. Petsko, Biophys. Chem. 105, 667 (2003). X. Yu and D.M. Leitner, J. Phys. Chem. B 107, 1698 (2003).
Ultrafast Vibrational Dynamics of AdenineThymine Base Pairs in Hydrated DNA J. R. Dwyer1,2, Ł. Szyc1. E. T. J. Nibbering1, and T. Elsaesser1 1
Max-Born-Institut für Nichtlineare Optik und Kurzzeitspektroskopie, D-12489 Berlin, Germany, E-mail: [email protected] 2 Department of Physics and Astronomy, University of British Columbia, Vancouver, B.C., Canada, V6T 1Z1 Abstract. We report femtosecond two-color pump-probe studies of the congested N-H/O-H stretching absorption of high-quality thin films of DNA oligomers in a broad hydration range. Different vibrational excitations are separated and their characteristic relaxation times identified.
Introduction Hydrogen bonds play a key role for the structure of DNA and its interaction with an aqueous environment. Intermolecular hydrogen bonds define the planar WatsonCrick geometry of adenine-thymine (A-T) and guanine-cytosine base pairs in the double helix [1]. Moreover, the interaction of water molecules with different parts of the DNA structure is mediated through hydrogen bonding [2]. Linear vibrational spectroscopy has been applied to identify particular functional groups and characterize different DNA structures. Such work has led to conflicting conclusions as the vibrational spectra of DNA are highly congested, both in the fingerprint range and in the range between 3000 and 3600 cm-1 where different N-H and O-H stretching bands display a strong overlap. Ultrafast vibrational spectroscopy allows for a much more specific insight as different types of excitations are separated via their intrinsic dynamics and vibrational couplings are determined in a quantitative way [3,4]. So far, ultrafast vibrational spectroscopy has mainly concentrated on fingerprint vibrations [4]. Here, we present femtosecond two-color pump-probe studies of highfrequency N-H and O-H stretching excitations in artificial DNA oligomers containing 23 A-T pairs. Transient vibrational spectra, vibrational relaxation and anisotropy decay times are measured for different levels of DNA hydration. We discern N-H stretching excitations of the A-T base pairs from O-H stretching excitations of water molecules even at a high hydration level and determine the ultrafast dynamic properties of these N-H stretching excitations.
Experimental Methods In our experiments, we study artificial DNA oligomers containing 23 alternating A-T pairs. Thin supramolecular DNA films of approximately 10 micron thickness and high structural quality were prepared on 500 nm thick Si3N4 substrates. The sample preparation preserves the double helix structure and allows changes in water content to drive conformational changes well-known from native DNA. The substrate is fully transparent in the frequency range studied here and makes a negligible contribution to the nonlinear response measured in the ultrafast experiments. The DNA sample was part of a closed sample cell in which a well-defined moisture level was maintained. 535
Independently tunable pump and probe pulses were generated in two parametric frequency converters driven by amplified pulses from a Ti:sapphire laser (repetition rate 1 kHz). The energy of the pump pulses of 200 cm-1 bandwidth was ~2 µJ, the temporal width of the cross-correlation with the probe pulses was 150 fs. After interaction with the sample, the probe pulses were dispersed in a monochromator and detected with a 16-element HgCdTe detector array.
Fig. 1. (a) Infrared absorption spectra of DNA oligomers containing 23 adenine-thymine base pairs for 0% and 93% relative humidity (r.h., solid lines). Dash-dotted lines: spectra of the femtosecond pump pulses. (b,c) Transient infrared spectra at 0% r.h. for pump-probe delays of 200 fs (solid circles) and 1 ps (open circles) after excitation centered at Eex=3250 cm-1 and Eex=3550 cm-1. The change of absorbance is plotted as a function of probe frequency. (d) Transient infrared spectra at 93% r.h. for the same pump-probe delays after excitation centered at Eex=3550 cm-1.
536
Results and Discussion The (linear) vibrational absorption spectra of the DNA films shown in Fig. 1 (a) (solid lines) display a broad absorption band between 3050 and 3700 cm-1, i.e., in the range of N-H and O-H stretching absorption. With increasing humidity, both a reshaping of the spectral envelope with enhanced high-frequency components and an increase of the overall absorption strength occur. The (nonlinear) transient spectra in Figs. 1 (b,c) reveal different vibrational transitions contributing to the overall absorption. For 0% relative humidity (r.h.), there is a pronounced bleaching on vibrational fundamentals (v=0 to 1 transitions) with maxima at 3200 cm-1, 3350 cm-1 (Fig. 1 b), and 3500 cm-1 (Fig. 1 c). The enhanced absorption at low frequencies is due to the v=1 to 2 transitions of the different oscillators and decays with the vibrational lifetimes of ~0.5 ps (not shown). The fundamental transitions display a negligible spectral diffusion up to pump-probe delays of 10 ps. This behavior is attributed to both the absence of structural fluctuations of the hydrogen-bonded dimers and a limited Coulomb interaction with the spatially separated ionic phosphate groups of the DNA backbone that undergoes low-frequency (sub-20 cm-1) fluctuating motions. Taking into account the vibrational spectra of isolated A-T pairs in the gas phase [5], we assign the band around 3200 cm-1 to a superposition of the symmetric NH2 stretching vibration of adenine and the stretching vibration of the hydrogen bonded NH group of thymine. The peaks at 3350 cm-1 and 3500 cm-1 are attributed to the asymmetric NH2 stretching vibration of adenine and the O-H stretching absorption of the residual H2O molecules (~2 water molecules per base pair), respectively. The O-H stretching absorption occurs at higher frequencies than in bulk water [6] due to the strongly modified hydrogen bonding with the phosphate groups of DNA. This behavior is similar to water in micelles where water molecules interact with the micelle's polar head groups [7,8]. Time-resolved pump-probe transients at fixed probe frequencies of 3200 cm-1 and 3335 cm-1 (not shown) demonstrate a femtosecond decay of polarization anisotropy to a constant residual value of ~0.18, pointing to a pronounced coupling of the different N-H stretching excitations. In contrast, the anisotropy of the O-H streching component has a constant value of 0.4. For 93% r.h., the broad O-H stretching component is well-pronounced, in particular at a 200 fs delay (Fig. 1 d). In contrast to 0% r.h., this band exhibits a distinct spectral diffusion towards smaller frequencies. The subpicosecond time scale of such spectral evolution is again much slower than in bulk water [6], and in line with the comparably slow spectral diffusion of water in small micelles [8]. 1 2 3 4 5 6 7 8
J. D. Watson and F. H. C. Crick, Nature 171, 737, 1953. M. Ouali, H. Gousset, F. Geinguenaud, J. Liquier, J. Gabarro-Arpa, M. Le Bret, and E. Tallandier, Nucleic Acids Res. 25, 4816, 1997. E. T. J. Nibbering and T. Elsaesser, Chem. Rev. 104, 1887, 2004. A. T. Krummel, P. Mukherjee, and M. T. Zanni, J. Phys. Chem. B 107, 9165, 2003. C. Pluetzer, I. Huenig, K. Kleinermanns, E. Nir, and M. S. de Vries, ChemPhysChem 4, 838, 2003. M. L. Cowan, B. D. Bruner, N. Huse, J. R. Dwyer, B. Chugh, E. T. J. Nibbering, T. Elsaesser, R. J. D. Miller, Nature 434, 199, 2005. D. Cringus, A. Bakulin, J. Lindner, P. Vöhringer, M. S. Pshenichnikov, and D. A. Wiersma, J. Phys. Chem B 111, 14193, 2007. H. S. Tan, I. R. Piletic, R. E. Riter, N. E. Levinger, and M. D. Fayer, Phys. Rev. Lett. 94, 057405, 2005.
537
Ultrafast Vibrational Dynamics in the AppA Blue Light Sensing Protein Allison Stelling,2 Minako Kondo,1 Kate L. Ronayne,3 Peter J. Tonge2 and Stephen R. Meech1 1
School of Chemical Sciences and Pharmacy, University of East Anglia, Norwich NR4 7TK, UK E-mail: [email protected] 2 Department of Chemistry, Stony Brook University, Stony Brook, New York 11794-3400, USA E-mail: [email protected] 3 Central Laser Facility, Harwell Science and Innovation Campus, Didcot, Oxon OX11 0QX, UK Abstract. The mechanism of blue light sensing in the photoactive protein AppA is
investigated by transient infra-red spectroscopy. Modes associated with the flavin excited state and perturbation of the protein are detected.
Introduction AppA is one of a number of recently characterised blue light sensing proteins. These have important roles in, for example, regulating phototropism, circadian rhythms and photosystem biosynthesis. Specifically AppA is a transcriptional anti-repressor found in photosynthetic bacteria.[1,2] Under low light and oxygen conditions it is bound to the transcriptional repressor PspR, preventing PspR from binding to DNA. However, under strong light (or high oxygen) the AppA-PspR complex dissociates, allowing PpsR to repress the synthesis of photosynthetic proteins. AppA is bifunctional, being sensitive to light and oxygen levels, but here we are concerned with the chromophore containing BLUF (blue light sensing using flavin adenine dinucleotide (FAD)) domain, the portion of the protein responsible for sensing light levels. In many photoactive proteins, such as the rhodopsins and photoactive yellow protein the driving force for formation of the signaling state is a large scale structural change in the excited electronic state of the chromophore. Intriguingly in the planar FAD chromophore no such possibility exists. Moreover, in AppA, unlike the flavin based LOV domain and cryptochrome light sensing proteins, no excited state photochemistry is observed. Instead the only spectroscopic signature of formation of the signaling state is a small red shift in the electronic absorption spectrum between the dark(d) and light(l) adapted states. Recent ultrafast time resolved studies in the visible region showed that the red shift occurs on a 1 ns time scale.[3] It was proposed that the primary step was an electron and proton transfer between the FAD chromophore and a nearby highly conserved tyrosine residue. Importantly, structural information for both l and dAppA exist. Most agree that an important structural change is the reorientation of the Q63 residue adjacent to the flavin ring.[4] Such a reorientation (illustrated in Figure 1) might be a result of changes in the extensive network of protein – chromophore H-bonds in AppA. Ultrafast vibrational spectroscopy is a powerful tool for studying changes in the proton transfer network following photoexcitation of chromoproteins [5,6] and this technique has now been applied to FAD and AppA.[7,8]. 538
Fig. 1. An illustration of the possible transformation in Q63 orientation between dark (dAppA) and light (lAppA) forms of AppA. Possible H-bond interactions are shown as dash lines. Data for this figure were based on the structure presented in [4].
Experimental Methods Solutions of AppA and its mutants were prepared in D2O solution at 0.1 – 1 mM concentrations, and therefore probably exist as dimers or higher aggregates. dAppA photodynamics were measured in a flow cell under low irradiation conditions such that negligible photoconversion occurred during the experiment. Mutants were studied in a raster scanned cell, and the photoactive mutant was exposed to radiation for less than 1 minute before being replaced and allowed to recover in the dark. Excitation was at 400 nm and transient IR difference spectra were measured between 1600 and 1750 cm-1.
Results and Discussion The transient IR difference spectra have been measured between 1 ps and 1 ns after excitation for dAppA, lAppA, the photoactive mutant W104F and the photoinactive mutant Q63L.(7) Both dAppA and W104F are observed to exhibit the characteristic red shift under continued irradiation at 400 nm, while the absorption spectra for lAppA and Q63L are independent of irradiation The transient IR data for the light and dark adapted forms are shown in Figure 2 2.0E-04 2.E-04
1.0E-04
Delta OD
Delta OD
5.E-05
0.0E+00 -1.0E-04
-2.E-04
-2.0E-04 -3.0E-04 1550
-5.E-05
-3.E-04
1600
1650 Wavenumber
1700
1750
-4.E-04 1550
1600
1650
1700
1750
Wavenumber
Fig 2 (a) Transient IR spectra for dAppA recorded between 1 ps and 2 ns after excitation, characterized by an instantaneous bleach and slow recovery. (b) The same for lAppA between 1ps and 200 ps after excitation.
539
The four bleach bands observed in both cases are readily assigned to modes of the isoalloxazine ring of the flavin chromophore on the basis of our previous study [8]: the two highest frequency bands are mainly localised on the carbonyl groups; the two bleach modes below 1600 cm-1 are largely due to ring CN stretches; the transient absorption around 1600 cm-1 is ascribed to carbonyl modes in the excited state. There are three significant differences between the spectra of the light and dark adapted states. First the highest frequency C=O mode evolves into a doublet under irradiation. Second the ground state recovery is more rapid in the light adapted form, suggesting quenching of the excited state. Finally the dark adapted form has a transient absorption at 1666 cm-1 which is absent in lAppA. In a study of the photoactive W104F mutant the same three changes were observed on irradiation.[7] Significantly in neither the dark nor the light adapted transient spectra is there any evolution in the shape of the spectrum as a function of time during at least 1 ns. The doubling of the carbonyl mode and the quenching are provisionally ascribed to structural disorder induced by irradiation. The doubling of the carbonyl mode suggests that FAD in lAppA occupies at least two distinct sites. The observation of quenching of the flavin excited state in the light adapted form may suggest an enhanced electron transfer rate. One possibility is that the separation between FAD and a tyrosine or tryptophan residue is reduced as a result of disorder. Alternatively disorder may allow mobile water molecules into the FAD binding site; rapid reorientation of water molecules could contribute to enhanced electron transfer rates. The absence of any observable intermediates being formed in dAppA during the one nanosecond time window of our experiments is surprising. It was previously proposed that the primary step involved electron and proton transfer from the nearby tyrosine residue (Y21) to the flavin ring. Such an electron transfer reaction is feasible on both structural and energetic grounds. The absence of a spectrum assignable to a radical in the IR may simply reflect the kinetics or a low oscillator strength in the spectral region investigated, or suggest an alternative mechanism; model studies are needed The most interesting observation concerns the instantly formed (295 nm, with wavelength-independent kinetic traces in the range 305-320 nm (see Fig. 2-left-top). Figure 2-right displays the kinetic trace of the Trp86 contribution spectrally averaged from 305 to 320 nm. A biexponential fit yields the well-known retinal isomerization time 450 +/- 50 fs and a second time constant much longer than
545
100 fs 2 to 4 ps
1 0 -1
100 fs 2 to 4 ps
1
∆A [mOD]
∆A (mOD)
∆A (mOD)
our observation time window (up to 20 ps, not shown here). The amplitude associated wih the fast decay component is 2.2 +/- 0.3 times larger than that of the slow one. From the above interacting-dipole model, it appears that the positive induced absorption shown in Fig. 2-right is a measurement of the transient Stark shift of the Trp86 absorption band induced by the dipole moment change of PSBR. The latter decreases while isomerization proceeds as observed by the 450 fs decay time. The remaining positive absorption at later times suggests that the S0 dipole moment of J and K states remains larger than that of all-trans bR, but other scenarios can be envisaged as well. In addition, the signal rises within the instrument response function indicating that a possible twist-induced charge translocation [4] has to occur on a sub80 fs time scale. This conclusion now based upon a sound assignment of the Trp86 contribution rectifies results from our previous work.
0.5
0
0.0
-1 260
270
280
290
300
310
Wavelength (nm)
320
0
1 2 3 Time delay (ps)
4
Fig. 2. Left: Individual spectral contributions of Trp86, and Trp182 obtained by linear superposition of the raw data displayed in Figure 1: Trp86 = wt bR W86F, Trp182 = wt bR - W182F. Right: transient contribution of the Trp86 residue in the range of 305-320 nm. Overlaid is a fit to a two-component exponential decay convolved with a Gaussian response function.
This new systematic measurement and comparison of the transient UV absorption spectroscopy of wt bR and the two mutants W86F and W182F allows us to isolate the spectral contribution of the Trp’s interacting with the PSBR. The results demonstrate that transient absorption spectroscopy of Trp’s can reveal ultrafast changes of dipole moments on nearby moieties. The strong distance dependence of excitonic interactions introduces a natural selectivity among the many Trp's that the protein may contain (8 in bR). A further refined analysis of the excitonic interaction between Trp’s and PSBR is in progress to explore the possibility to turn this approach into a quantitative measurement of the time-dependent ∆µ. Acknowledgements. J.L. acknowledges support from the DYNA programme of the European Science Foundation. 1 2 3 4
546
A. Colonna, G. I. Groma, J.-L. Martin, M. Joffre, M. H. Vos, J. Phys. Chem. B, 111, 2707 -2710 (2007). S. Schenkl, F. van Mourik, G. van der Zwan, S. Haacke, M. Chergui, Science, 309, 917 (2005). S. Schenkl, F. van Mourik, N. Friedman, M. Sheves, R. Schlesinger, S. Haacke, and M. Chergui, PNAS 103, 4101 (2006). R. Gonzalez-Luque, M. Garavelli, F. Bernardi, M. Merchán, M. Robb, M. Olivucci, PNAS, 97, 9379 (2000).
Interrogating Fiber Formation Kinetics with Automated 2D-IR Spectroscopy David B. Strasfeld1, Yun L. Ling1, Sang-Hee Shim1 and Martin T. Zanni1 1
Department of Chemistryt of Chemistry, University of Wisconsin-Madison, 1101 University Ave. Madison, Wisconsin 53706-1322 E-mail: [email protected]
Abstract. A new method for collecting 2D-IR spectra that utilizes both a pump-probe beam geometry and a mid-IR pulse shaper is used to gain a fuller understanding of fiber formation in the human islet amyloid polypeptide (hIAPP). We extract structural kinetics in order to better understand aggregation in hIAPP, the protein component of the amyloid fibers found to inhibit insulin production in type II diabetes patients.
1.
Introduction
The formation of amyloid fibers by the human islet amyloid polypeptide (hIAPP) has been indicated as a primary cause of β-cell death in type II diabetes patients. The structural kinetics that dictate this transformation remain obscure due to interrogation primarily by circular dichroism (CD) and a fluorescence shift in protein bound dyes [1,2]. These two spectroscopies cannot monitor different secondary structural confirmations independently and simultaneously. Two dimensional infrared (2D-IR) spectroscopy offers an improved capacity to resolve protein secondary structures and added structural insight from coupling indicative cross peaks. The monitoring of amyloid formation with 2D-IR spectroscopy is hindered by two inherent difficulties: the fact that the kinetics of amyloid fiber formation are not perfectly reproducible and the enormous background noise attributable to light scattering by the amyloid fibers. We have learned to overcome these impediments by automating data collection [3] with a mid-IR pulse shaper. The issue of a scattering background is resolved by rotating the phases of subsequent pulses in the pump-pulse train, which, in turn, shifts the scatter frequency away from desired signal elements. Replacing mechanical stages with a pulse shaper greatly reduces the time necessary to generate a 2D spectrum, allowing us to take a single scan in 2 ps |Etr (t)| > |Ein (t)|, i.e., the sample shows stimulated emission. Absorption is characterized by a phase of φ = π between the emitted field Eem (t) and the local driving field Etr (t), whereas for gain we find φ = 0. The time dependent energy transfer rate between the THz field and the sample is given by: ¤ £ (1) Iabs (t) = ε0 c Ein (t)2 − Ere (t)2 − Etr (t)2 = 2 ε0 c Eem (t) Etr (t).
The energy that is transiently absorbed per electron is the time integral of Iabs (t): Λ(t) =
1 ND d
Z t
−∞
Iabs (t 0 ) dt 0 .
(2)
Iabs (t) and Λ(t) are shown in Figs. 1(b) – (g) as the solid lines for different electric field amplitudes together with the results from a theoretical calculation (dashed lines).
658
In the linear case [Figs. 1(b) and (e)] Iabs (t) oscillates around zero due to the slightly off-resonant center frequency of the driving pulse (2 THz) compared to the 1S – 2P transition frequency (1.5 THz). Hence, the transiently absorbed energy Λ(t) increases with strong oscillations as a function of time. For higher field amplitudes one finds a change between absorption and stimulated emission [Figs. 1(b) and (g)], a typical signature of Rabi oscillations. We model our result with the Maxwell-Bloch equations for a two-level system [4,6]. For strong radiative coupling which results from the two-dimensional sample geometry and the high doping, the local driving field acting on the sample is given by Eloc (t) = Etr (t) = Ein (t) + Eem (t). The strong radiative coupling leads to a collective response of the ensemble of impurity transitions and therefore to a much longer decoherence time compared to the individual two-level system. The result of the calculation is shown as dashed lines in Figs. 1(b) – (g). Since in the case of strong radiative coupling the energy and phase relaxation times are of minor relevance, the model does not contain any fitting parameters. The quantitative agreement between experiment and theory is good for field amplitudes up to 5 kV/cm. For higher amplitudes [Figs. 1(d) and (g)] the two-level approximation breaks down. This is seen in Fig. 1(g) where at t = 1.25 ps the energy per carrier Λ(t) is much larger than the maximum energy h¯ ω2P−1S = 5.5 meV possible in the two-level model [shown as the dotted line in Figs. 1(e) – (g)], pointing to a significant population of other levels during the driving pulse.
Conclusion We studied the THz response of a thin n-type GaAs layer at low temperatures. The emitted THz radiation is directly measured in phase-resolved nonlinear propagation experiments and demonstrates carrier-wave Rabi oscillations between bound levels of shallow impurities. The Rabi oscillation picture holds for driving fields up to 5 kV/cm. For higher driving fields the two-level approach breaks down. 1 2 3 4 5 6
B. E. Cole, J. B. Williams, B. T. King, M. S. Sherwin, and C. R. Stanley, Nature 410, 60, 2001. M. F. Doty, B. T. King, M. S. Sherwin, and C. R. Stanley, Physical Review B 71, 201201(R), 2005. P. Gaal, K. Reimann, M. Woerner, T. Elsaesser, R. Hey, and K. H. Ploog, Physical Review Letters 96, 187402, 2006. T. Stroucken, A. Knorr, P. Thomas, and S. W. Koch, Physical Review B 53, 2026, 1996. S. Hughes, Physical Review Letters 81, 3363, 1998. C. W. Luo, K. Reimann, M. Woerner, T. Elsaesser, R. Hey, and K. H. Ploog, Physical Review Letters 92, 047402, 2004.
659
Nonlinear optical effects in germanium in the THz range: THz-pump - THz-probe measurement of carrier dynamics János Hebling1,2,, Matthias C. Hoffmann1, Harold Y. Hwang1, Ka-Lo Yeh1, Keith A. Nelson1 1
Dept. of Chemistry, Massachusetts Institute of Technology, Cambridge, MA 02139, USA E-mail: [email protected] 2 Dept. of Experimental Physics, University of Pécs, 7624 Hungary E-mail: [email protected] Abstract. Nonlinear high intensity THz transmission measurements were made on Ge, showing nonlinear optical effects in the THz range including self-phase modulation and absorption saturation. THz-pump - THz-probe measurements were performed on n-type Ge, GaAs, and Si to follow the ultrafast free carrier dynamics that underlie the nonlinear effects seen in the transmission measurements. Complex relaxation behavior was observed, arising from inter- and intra-valley scattering in the conduction bands of these semiconductors.
Introduction Although there has been extensive investigation of nonlinear terahertz effects in semiconductors with long pulses or continuous wave radiation [1], time-resolved studies on the ps timescale using ultrashort THz pulses are rare [2-4]. These recent investigations explore bound systems like excitons [2], impurity states [3] or polarons [4]. Nonlinear optical effects of free carriers in semiconductors were observed earlier in the THz range using a far-infrared molecular laser [5]. Although, in accordance with theory, a relatively strong third-order nonlinearity was observed, a detailed analysis of the nonlinear optical phenomena requires high intensity THz sources with temporal resolution. Using tilted pulse front excitation [6] we are able to routinely produce near-single-cycle or few-cycle THz pulses with intensities up to 300 MW/cm² [7,8]. Here we describe the use of these pulses for observation of self-phase modulation (SPM) and absorption saturation dynamics in n-type Ge. THz pumpprobe studies on n-type Si and GaAs as well as Ge are presented.
Experimental Methods For nonlinear THz transmission measurements, high intensity THz pulses were generated by optical rectification of 100 fs duration 5.6 mJ pulses from a 1 kHz Ti:sapphire laser system (Coherent LEGEND) in lithium niobate using the tilted pulse front (TPF) setup [7]. THz pulses with up to 3 μJ energy were generated in this way. These pulses were focused into the sample by two parabolic mirrors. The THz intensity could be attenuated continuously by rotating one of two wiregrid polarizers inserted before the sample. The transmitted THz radiation was collimated and lightly focused into a sandwiched ZnTe crystal with 1 mm total thickness and 0.1 mm sensitive ([110] surface) thickness. The THz field with and without a sample in the THz focus was recorded for different THz intensities by electro-optic sampling. The 660
maximum THz intensity (for 2.6 μJ THz pulse energy at the sample position) was about 250 MW/cm2. For comparison, a linear transmission measurement was carried out using a THz-TD spectrometer based on photoconductive switches with THz pulse energies of only several femto-Joules. THz-pump - THz-probe measurements were performed in a collinear geometry with a similar setup. The optical pump for THz generation was split into pump and probe arms using a beamsplitter, then recombined for TPF excitation in the LN generation crystal. A delay stage in the pump arm allowed for arbitrary separation of the THz pump and probe pulses in time.
Results and Discussion Free carrier absorption in the THz range is readily observed in bulk doped semiconductors, and can be described by the Drude model [5]. Linear THz-TDS measurements for Ge (Figure 1b) show standard Drude behavior as expected. At high THz pulse energies, THz absorption by free carriers saturates, corresponding to a dramatic phase shift in the THz field temporal profile, indicating self-phase modulation in Ge. As the THz intensity is attenuated, absorption returns to linear Drude behavior (a)
(b)
Fig. 1. (a) Electro-optically measured temporal profile of transmitted THz pulse in nonlinear transmission measurements. (b) Corresponding absorption spectra for different THz pulse energies. Linear response is plotted for comparison.
Time and frequency resolved THz-pump - THz-probe measurements were employed to follow the dynamics of free carrier absorption saturation in Ge (Figure 2a). At zero probe delay, absorption saturation occurs as expected from the high intensity transmission measurements. At positive probe delays, absorption recovery on a several ps timescale is observed (Figure 2b). The complex behavior in absorption is related to the mobility of carriers in the conduction band. The pump THz pulse excites the free carriers to states of higher energy in the initial conduction band valley from which they may scatter into side valleys. For Ge, electrons may be scattered from the L valley to the Γ and X valleys. Initial scattering into side valleys causes absorption saturation due to the lower mobility and higher effective mass of the electrons in the X side valley. Different relaxation times are expected for interand intra-valley relaxation. Half of the initially lost absorption is recovered after 2.5 ps in Ge, 1.5 ps in Si, and 2 ps in GaAs. All three semiconductors exhibit similar
661
absorption saturation and recovery. The details of the absorption recovery are intimately related to the conduction band structure of each sample. (a)
(b)
Fig. 2. (a) Nonlinear absorption spectra in Ge for different probe delay times. Linear THz-TDS measurement and measurement without pump pulse are shown for comparison. (b) Spectrally integrated absorption coefficient of Ge from THz-pump - THz-probe measurements.
Conclusions Self-phase modulation and absorption saturation in the THz range were observed in Ge from THz transmission measurements. The origin of these effects has been studied with time and frequency resolved THz-pump - THz-probe spectroscopy. High intensity THz pulses excite free carriers that may scatter into higher energy valleys in the conduction band. This leads to a change in absorption and refractive index, underlying the self-phase modulation mechanism. The absorption saturation originates from carrier scattering into regions of lower mobility. Subsequent relaxation back to the bottom of the initial conduction band valley occurs on the timescale of a few ps. Multiple relaxation pathways are expected, including electronphonon and electron-electron scattering from different regions of the conduction band. Measurements have been carried out on n-type Si and GaAs yielding similar results. Acknowledgements. This work was supported in part by Office of Naval Research Grant N00014-06-1-0459 1 2 3 4 5 6 7 8
662
S. D. Ganichev, and W. Prettl, in Intense Terahertz Excitation of Semiconductors, Oxford University Press Inc., New York, 2006. R. Huber et al., Phys. Rev. Lett. 96, 017402, 2006. P. Gaal et al., Phys. Rev. Lett. 96, 187402, 2006. P. Gaal, et al., Nature 450, 1210, 2007. Mayer A., and F. Keilmann, Phys. Rev. B 33, 6962, 1986. J. Hebling, G. Almási, I. Z. Kozma, and J. Kuhl, Opt. Express 10, 1161, 2002. K.-L. Yeh, J. Hebling, M. C. Hoffmann, and K. A. Nelson, Optics Commun. 281, 3567, 2008. K.-L. Yeh, M. C. Hoffmann, J. Hebling, and K. A. Nelson, Appl. Phys. Lett. 90, 171121, 2007.
Terahertz Nonlinear Response and Coherent Population Control of Dark Excitons in Cu2O T. Kampfrath1, S. Leinß2, K. v. Volkmann1, M. Wolf1, D. Fröhlich3, A. Leitenstorfer2, and R. Huber2 1
Fachbereich Physik, Freie Universität Berlin, Arnimallee 14, 14195 Berlin, Germany Fachbereich Physik, Universität Konstanz, Universitätsstrasse 10, 78464 Konstanz, Germany Email: [email protected] 3 Fachbereich Physik, Universität Dortmund, 44221 Dortmund, Germany 2
Abstract. An optically dark, dense, and cold 1s para exciton gas is prepared by two-photon generation of electron-hole pairs and subsequent phonon cooling. Intense multi-terahertz fields of order MV/cm coherently promote 70% of the quasiparticles from the 1s to the 2p state via a partial internal Rabi oscillation. Electro-optic sampling monitors the Larmor precession of the Bloch vector in real time.
Introduction Sophisticated quantum optical protocols had been a prerequisite for the first observation of Bose-Einstein condensation (BEC) of atomic gases [1]. Excitons – bosonic Coulomb pairs of one electron and one hole – have been envisaged as another potential candidate for BEC [2]. Yet optical control of promising excitons for BEC by intense light fields has been a challenge since relevant systems often exhibit weak if any radiative interband coupling. 1s para-excitons in Cu2O are a prime example [3]. Terahertz (THz) pulses, in contrast, couple resonantly to the internal hydrogen-like fine structure, irrespective of interband selection rules. This concept has provided novel insight into formation dynamics, fine structure, density, and temperature of excitons [4-6]. The observation of stimulated THz emission from intra-excitonic transitions [7] has raised the hope for future coherent manipulation of dark exciton ensembles similar to atomic quantum optics. Here, we exploit intense THz fields of the order of MV/cm to systematically study the nonlinear response of optically dark, dense, and cold 1s-para excitons in Cu2O. A partial Rabi flop of the 1s-2p transition allows us to control the internal quantum state on a sub-ps time scale.
Formation and cooling dynamics of 1s para excitons in Cu2O The sample is a 334-µm thick, naturally grown single crystal of Cu2O kept at a lattice temperature TL = 5 K. Near-infrared femtosecond pulses centered at an energy of 1.5 eV are absorbed via two-photon transitions to generate unbound electron-hole pairs with homogeneous density throughout the crystal length. We first trace the ultrafast formation and cooling dynamics of 1s excitons via time-delayed multi-THz transients probing the internal 1s-2p absorption. Strong exchange interaction splits the n = 1 state into a triplet ortho-exciton and a lower-lying, optically inaccessible singlet para-exciton. The respective 1s-2p Lyman absorption lines are located at photon energies of 116 meV and 129 meV [6]. Fig. 1(a) depicts the pump-induced absorption changes in this frequency window for various delay times tD after
663
(a)
para
ortho
generation of e-h pairs. While the spectrally broad THz absorption at tD = 100 fs indicates unbound e-h pairs, the hallmark 1s-2p lines are already discernable at tD = 11 ps. Within 100 ps, the lines narrow and shift slightly towards lower frequencies while the ratio of para- versus ortho-exciton absorption increases. The subsequent decay of exciton populations follows a complex non-exponential dynamics (not shown). The strength of THz absorption is an absolute measure of the densities N1s,para and N1s,ortho [5]. Owing to the vastly different effective masses of 1s and 2p excitons [3] a detailed analysis of the THz line shape reflects the temperature T1s of the ensemble [8]. For tD = 100 ps [dots in Fig. 1(a)], we find best agreement with the experiment for a 1s-para exciton density N1s,para = 2 × 1016 cm-3 and a temperature T1s = 8 ± 2 K, close to the phonon bath. To our knowledge, this density is among the highest directly measured for a cold exciton gas in Cu2O.
(b) 0.5 (b)
tD =
∆α (cm )
100 fs -1
3 ps 11 ps 25 ps
10
100 ps
0
1.4 ns 110 120 130
Energy (meV)
Electric Field (norm.)
0 fs
0.0 -0.5
Elow ∆Elow ∆ρ
-1.0
Py
1.0 (c) (c) 0.5
Px
0.0 -0.5 -1.0
∆Ehigh
-200 0 200 400 600 800 Delay Time (fs)
Fig. 1. (a) Pump-induced changes ∆α of the mid-infrared absorption for various delay times tD after two-photon absorption of 12-fs near-infrared pulses; vertical lines: energies of the 1s-2p resonances in the ortho and para system at vanishing momentum. Circles: Numerical line fit (see text). (b) Field-resolved polarization response (black curve) of 1s-para excitons to an external THz field (gray dashed curve, peak electric field: 0.1 MV/cm) resonant to the 1s-2p transition (tD = 1 ns). (c) Reemitted field for peak values of the driving field of 0.5 MV/cm. Inset: schematic of the Bloch sphere of the 1s-2p two-level system.
Terahertz nonlinear control of the internal quantum state of excitons We next demonstrate how to control the internal quantum state of the dark, dense, and cold para-exciton gas formed at tD = 1 ns. Intense THz transients with peak fields as high as 0.5 MV/cm are generated via optical rectification of the output of a highpower Ti:sapphire laser amplifier. An acousto-optic phase and amplitude modulator is used to prepare optimally shaped 0.1-mJ pulses for efficient phase-matched rectification in a GaSe emitter of a thickness of 200 µm.
664
Fig. 1(b) depicts a THz pulse of moderate intensity (Epeak = 0.1 MV/cm) on resonance with the two-level system consisting of the 1s- and 2p-para states. The oscillating electric field of the reemitted light is directly measured in real time via ultrabroadband electro-optic sampling, on a femtosecond scale (black curve). The response is out of phase with the driving field by 180° characteristic of an absorptive transition in a resonant two-level system. The free induction decay following the external THz pulse exhibits a decoherence time of 0.8 ps. Fig. 1(c) compares this response with the polarization induced by a THz pulse of the same temporal profile, but a peak electric field scaled by a factor of 5. Surprisingly, the system polarization is not a linearly scaled version of the low-field response. Rather, the re-emitted field initially rises rapidly within the first seven cycles, reaches its maximum amplitude, and decreases within the decoherence time of the internal transition. This dynamics is a manifestation of a partial Rabi oscillation explained in a Bloch picture of the two-level system (Fig. 1, inset). The diagonal (population difference ρ22-ρ11) and off-diagonal (polarizations Px and Py) elements of the density matrix are depicted as the vertical and the two horizontal coordinates. For low driving fields, the Bloch vector performs a Larmor precession in the vicinity of the south pole of the Bloch sphere. Experimentally, the projection of this trajectory onto the polarization axis is directly mapped out in real time [Fig.1(b)]. With increasing THz field, the state vector may be driven towards the north pole inducing strong population inversion. During this partial Rabi cycle, the projection onto the polarization axis reaches a maximum at the equator and decreases from there on. The real-time data of the polarization [Fig.1(c)] directly reflect this dynamics and allow us to reconstruct the actual motion of the Bloch vector. From a comparison of the data with a numerical solution of the Maxwell-Bloch equations, we find that up to 70% of the optically dark states are promoted from the 1s into the 2p orbital. For yet larger driving pulse areas, the reemitted field exhibits an oscillatory dependence on increasing field strength, indicative of up to 1.5 Rabi cycles (not shown). Quantitative modelling of the nonlinear THz data with state-of-the-art microscopic semiconductor theory, accounting for ponderomotive contributions as well as highly excited and ionized exciton states is under way [9, 10]. In conclusion, our results point out a novel route towards ultrafast nonlinear control of optically dark exciton states. Advanced protocols known from atomic systems may now open new perspectives for preparing ultracold and dense exciton gases via high-field THz transients. 1
M. H. Anderson, et al., Science 269, 198 (1995); K. B. Davis et al., Phys. Rev. Lett. 75, 3969 (1995). 2 J. Kasprzak et al., Nature 443, 409 (2006). 3 J. Brandt et al., Phys. Rev. Lett. 99, 217403 (2007) and references therein. 4 R. A. Kaindl et al., Nature 423, 734 (2003). 5 R. Huber et al., Phys. Rev. B Rapid 72, 161314(R) (2005). 6 M. Kubouchi et al., Phys. Rev. Lett. 94, 016403 (2005). 7 R. Huber et al., Phys. Rev. Lett. 96, 017402 (2006). 8 K. Johnsen et al., Phys. Rev. Lett. 86, 858 (2001). 9 J.R. Danielson et al., Phys. Rev. Lett. 99, 237401 (2007). 10 M. Kira et al. in Progress in Quantum Electronics 30, 155 – 296 (2006), G. Eden ed. (Elsevier, Amsterdam, 2006).
665
Impact Ionization in InSb studied by THz-PumpTHz-probe spectroscopy Matthias C. Hoffmann1,*, János Hebling1,2, Harold Y. Hwang1, Ka-Lo Yeh1and Keith A. Nelson1 1
Department of Chemistry, Massachusetts Institute of Technology, Cambridge, Massachusetts 02139, USA *E-mail: [email protected] 2 Department of Experimental Physics, University of Pécs, 7624 Hungary Abstract. We observe impact ionization and saturation of free carrier absorption in indium antimonide at 200K and 80K. We employ a novel THz-pump-THz-probe scheme with pump fields up to 100 kV/cm amplitude and 100 MW/cm2 intensity.
Introduction Indium antimonide has the highest electron mobility and saturation velocity of all known semiconductors. This makes transistors with extremely high switching speed possible [1]. The carrier dynamics on the ultrashort timescale are hence of great technological as well as fundamental interest. The material is a direct semiconductor with a band gap of 170 meV at room temperature, making it well suited for applications in infrared sensors covering the wavelength range from 1 − 5 µm[2]. Impact ionization by high electric fields is a well known phenomenon in InSb [3,4]. Strong THz fields can directly achieve impact ionization on the picosecond time scale, avoiding additional experimental complications by using photon energies well below the band gap. An intensity dependent THz pump experiment on InSb has been reported recently by Lindenberg [5]. Due to their strong interaction with free carriers, THz pulses can also be used as a very sensitive probing tool to monitor the subsequent carrier dynamics after THz excitation. In this article we demonstrate THz-pump-THz-probe spectroscopy and apply it, with frequency as well as time resolution, to the study of impact ionization and hot carrier effects caused by THz fields greater than 100 kV/cm in InSb.
Experimental Methods Our THz pump-probe method was based on the use of intense THz pulses generated by tilted pulse front excitation in LiNbO3 [6]. A 6-mJ optical pulse from a kHz repetition rate Ti:sapphire amplifier was split into two parts using a 10:90 beam splitter and the more intense pulse was variably delayed. The optical pulse fronts were tilted with a grating-lens combination and directed to a common region of a LiNbO3 crystal to generate collinear THz pump and probe pulses that could be variably delayed. The single-cycle THz pulses were focused to the sample inside a cryostat using a set of off-axis parabolic mirrors resulting in a focus size of 1 mm, as verified by a razor blade scan. The THz pulse energy at the sample spot was 2 µJ. A second pair of parabolic mirrors was used to relay-image the THz field onto an electro-optic (EO) sampling setup. A ZnTe crystal with 0.1 mm active and 1 mm total thickness was used to keep the EO signal in the linear range. EO traces of the THz pulses were recorded with and without the sample in place at every pump step 666
covering a time window of 45 ps. Chopping only the laser beam used to generate the probe THz pulse ensured efficient suppression of the pump THz field. The sample was n-type InSb:Te (110) with a thickness of 450 um and a carrier concentration of 2.0×1015 cm-3 at 77 K, and a manufacturer-specified mobility of 2.5×105cm2/Vs at room temperature.
Results and Discussion From the sample and reference data we can calculate a frequency averaged effective absorption coefficient 2 ⎛ dtEsam (t ) ⎞ ⎟ 2 ⎜ ∫ dtEref (t ) ⎟ ⎝ ⎠
α eff = − ln ⎜ ∫ 1 d
where d is the sample thickness and E(t) is the electro-optically measured probe field temporal profile. Figure 1 shows our pump-probe results at 200K and 80K. At 200 K, the initial effect of the pump pulse is to accelerate the preexisting carriers, providing access to higher-energy states (including those in neighboring valleys) with lower mobility and higher effective mass. This results in reduced THz probe pulse absorption, consistent with our observations in THz pump-probe measurements of Ge, Si, and GaAs as well as observations of THz absorption saturation in nonlinear transmission measurements of these materials and InSb (reported elsewhere in this volume). The THz probe absorption then increases and reaches levels far higher than its initial value, indicating the presence of additional carriers produced through impact ionization. This behavior was not observed in the other semiconductors studied, whose bandgaps are too wide for impact ionization to occur under these conditions. Carrier cooling during the next 10 ps leads to continued increase in THz probe absorption. Partial reflection of the THz pump pulse within the sample leads to a second weaker response starting at around 10 ps that adds to the rising signal response already under way. At 80 K there are fewer preexisting carriers and their initially reduced absorption appears to be outweighed by the increase in carrier population at all probe delay times. We observe an overall sevenfold increase in the absorption after 30 ps due to impact ionization. Assuming proportionality between the effective absorption and the carrier concentration, we estimate that 1.4x1016 cm-3 additional carriers were produced, using the ratio of the absorption coefficient before and after the THz pump and the known carrier concentration at 80K. 90
−1
effective absorption α (cm )
180 160
200K
80
140
70
120
60
100
50
80
40
60
30
40
20
80K
10
20 -5
0
5
10
15
probe delay (ps)
20
25
30
-10
-5
0
5
10
15
20
25
30
probe delay (ps)
Fig. 1. The effective absorption coefficient observed before and after excitation with a 2 µJ THz pulse at sample temperatures of 200K and 80K. The dashed lines show the shape of the pump THz pulse.
667
Figure 2 shows the frequency resolved absorption spectra at 200K and 80K for selected time points. At negative time delays (i.e. before THz excitation) the absorption follows essentially a Drude model. At 200 K the absorption at positive delays also can be described by a Drude model. At 80K the behaviour is remarkably different. In the absorption spectrum we observe a broad feature around 1.1 THz that might be related to interaction with phonon modes [7]. We did not observe this absorption peak in nonlinear transmission measurements at any temperature or THz field strength.
200K
80 K THz-Signal (mV)
200 -1
absorption α (cm )
200
40
150
100
20 0
30 ps 7ps
150
-20
THz-Signal (mV)
250
50 0 -50 -100
-40 6
8
10 time (ps)
12
14
5
100
6
7
8
9 10 11 12 13 14 15 time (ps)
100
0
0.2
50
-7 ps 7.4 ps 34 ps
50
0.4
0.6
0.8
1.0
1.2
1.4
1.6
1.8
2.0
0
0.4
0.6
0.8
1.0
1.2
1.4
1.6
1.8
2.0
frequency (THz) frequency (THz) Fig. 2. The absorption coefficient as function of frequency for selected pump-probe delays. The insets show the corresponding transmitted THz probe fields.
Conclusions The newly developed THz pump/THz probe technique is a sensitive method that monitors carrier dynamics in semiconductors on ultrafast timescales. In InSb at 200K, we observed an initial drop in THz probe absorption due to electron heating, then an increase in absorption beyond its initial level due the generation of new carriers through impact ionization. At 80 K the number of carriers increases by a factor of 7 and the absorption spectrum shows a long-lived non-Drude peak at 1.1 THz. Acknowledgements. This work was supported in part by ONR grant no. N00014-061-0463. 1 2 3 4 5 6 7
668
T. Ashley et al,. Proc. 7th Intl. Solid State and Integrated Circuits Tech. Conf., Beijing, China, pp. 2253-2256, (2004). D. G. Avery, D. W. Goodwin and Miss A. E. Rennie , J. Sci. Instr. 34, 394 (1957). C. L. Dick. and B. Ancker-Johnson, Phys. Rev. B. 5, 526 (1972). A. Lobad and L. A. Schlie, J. Appl. Phys. 95, 97 (2004). A.M Lindenberg, H. Wen, E. Szilagyi, CLEO 2008. K. Yeh, J. Hebling, M. C. Hoffmann and Keith. A. Nelson, Opt. Commun. 281, 3567 (2008). D.L. Price and J.M. Rowe, Phys. Rev. B 3, 1268 (1971).
Single Shot Linear Detection of THz Electromagnetic Fields on the Fs to Ps Scale Uli Schmidhammer, Vincent De Waele, and Mehran Mostafavi Laboratoire de Chimie Physique - ELYSE, UMR8000 CNRS-Université Paris Sud 91405 Orsay, France E-mail: [email protected] Abstract. We present single shot electro-optic sampling based on spectral decoding with a chirped supercontinuum: The wavelength dependent polarization state of the probe is analyzed in polychromatic balanced detection. The frequency bandwidth of over 300 THz and the linear, normalizing detection allow widely tunable and broadband diagnostic of THz waveforms without signal distortion. The technique is compatible to state of the art fs to ps laser.
Electro-Optic Sampling for THz Spectroscopy and e- Bunch Monitoring While the last decade the diagnostic of THz electromagnetic fields via electro-optic (EO) sampling techniques has been a field of greatly increasing interest. One motivation is the variety of applications in THz spectroscopy and imaging with the possibility of non-destructive measuring. On the other hand there is the development of ultrafast electron accelerators used for the generation of new brilliant radiation sources. The nearly transverse Coulomb field of such relativistic electron pulses is correlated to their longitudinal shape allowing for non-invasive characterisation in a nearby placed electro-optic crystal. Here the need for synchronisation of the bunch with accelerating fields or external experiments is inherently linked to the single shot capability of the diagnostic. The fastest possible diagnostic is also desired for the THz characterisation of moving objects, dynamic processes or experiments with low repetition rate. In general there are two ways to substitute the delay line for conventional, repetitive EO scanning: the birefringence induced in the EO crystal by the THz electric field is encoded as polarization state modulation either spatially to the wavefront of the optical probe pulse or to its temporally dispersed spectrum. The existing techniques are compared in a more exhaustive manner in [1]. The often called spatial encoding techniques are well suited for the sub-ps regime but usually restricted to some ps. One dimensional imaging of the THz waveform in the EO-material is not possible without lateral averaging. In contrast, the spectral encoding can be applied in a stable and tunable manner on the ps scale [2]. Temporal resolution and detection window are in this case connected in analogy to the energy-time uncertainty of the Heisenberg principle via the bandwidth of the probe pulse [3]. This effect limits the applicability especially towards the short time scale in the way that only for the state of the art sources delivering ultrabroad spectra a reasonable operating range can be chosen. Moreover, the single shot EO techniques presented in the literature use a detection scheme based on crossed polarizer operated at (near) optical bias. In contrast to the state of the art high sensitive delay scan methods this configuration needs recording of only one polarization state. However its signal response is nonlinear with different terms that depend on the intensity distributions of the probe, the optical bias, scattering contributions and the amplitude of the electromagnetic field to be detected. Signal distortions and possible artefacts are the consequence [4,5].
669
EO Sampling by Supercontinuum Encoding with Balanced Detection These restrictions of single shot EO sampling have been overcome to a large extent in the following way: The temporal distribution of the electric field in the EO crystal is encoded to a supercontinuum (SC) whose wavelength dependent polarisation state is then analyzed with balanced detection (see Fig. 1). The generation of the SC not only delivers an ultrabroad bandwidth of the probe whose Fourier Tranform limit is < 5 fs. This strongly nonlinear process allows decoupling the characteristics of the diagnostic to a high degree from the initial laser source. So the required intensity to start SC generation can easily be achieved at different wavelengths, pulse durations and energies just by focusing an appropriate amount of energy smoothly into a sufficiently non-linear medium. In the presented setup we focus about 1 µJ of a Ti: Sapphire laser with a duration of about 200 fs into a Sapphire plate and obtain a highly stable single filament SC covering the visible spectral range.
Fig. 1. Scheme of the electro-optic single shot diagnostic based on supercontinuum decoding in polychromatic balanced detection. The configuration for the non-invasive monitoring of the longitudinale electron pulse shape at the laser triggered ps radiolysis facility ELYSE is shown.
In the following the spatial and temporal dispersion of the ultrabroad probe must be controlled. The inherent dispersion of the setup is kept on the sub-ps level up to the EO material. From this minimum detection window suited for highest temporal resolution the time scale can be adapted just by adding glass as pulse stretcher. The remaining dispersive optics must be chosen to be of high achromatic quality in order to avoid significant wavelength dependent beam deviations. So the used polarizer (P) and polarizing beam splitters (PBS) are based on the optical wiregrid technology. To optimize the balance of the two perpendicular polarization components the effective pathways from the separating PBS are realized symmetrically.
Kerr Effect with an Optical Pulse as Fs Reference To test the reliability and sensitivity on the fs scale we used the electro-optic Kerr effect with the Ti:Sapphire laser source as electric field. The interacting medium was 1 mm quartz. This experiment provides well defined conditions concerning the electromagnetic field to be studied: free space THz pulses in contrast have usually complicated waveforms that are difficult to analyze in an independent and reliable single shot experiment. The calculated correlation time to wavelength was verified experimentally by varying the optical delay (see inset figure 2a). Moreover, the influence of the stability of the laser source generating the SC on the diagnostic can be elucidated. One clearly observes the shot to shot jitter of the laser (see figure 2a). On the other hand the base line of the EO-measurements does not exhibit significant instabilities, i.e. the single shot EO sampling is to a high degree laser fluctuation free.
670
Fig. 2. Three consecutive single shot EO measurements: (a) On the fs scale with the Ti:Sapphire laser pulse as optical Kerr gate in a fused silica plate. The inset shows the calculated and measured wavelength dependence of the time scale. (b) On the ps scale for e-bunches whose electric field was probed in 0.5 mm ZnTe. The arrival time of a series of 20 e-bunches is shown in the inset exhibiting a jitter of ~ 1.5 ps.
Monitoring of Electron Bunches on the Ps Scale Adapting the dispersion by adding glass the identical set-up can be used on the large ps scale. Passing through 20 cm SF57 results in a 60 ps time window over the spectrum transmitted by the EO crystal, i.e. 0.5 mm ZnTe, and a resolution 12ps and x < 0.55mm) and the rest was transmitted trough the structure (t > 20ps and x > 1.5mm). The velocity of the transmitted THz wavepacket was higher than the bulk group velocity due to the lower effective index of refraction of the photonic structure. The strong temporal chirp of the transmitted signal indicates a strong dispersion. For a more quantitative analysis we performed a two dimensional Fourier transformation of the data shown in figure 1b) (transmitted part only, i.e. x > 1.5mm) with the result shown in figure 1c). The intensity plot shows the content of the THz wavepacket in kx -ω space which can be directly compared to the calculated band structure. Ideally, the maxima of the 2D Fourier transform should coincide with the dispersion functions of the different modes. The bandgap simulation were made in two dimensions, whereas the measured system is 2.5 dimensional. Nevertheless, the position of the band gaps should be predicted correctly as can be seen in figure 1c). All frequency components within the first bandgap are clearly missing in the transmitted waveform.
Conclusions We analyzed the propagation of THz phonon-polaritons in and close to 1D and 2D photonic crystal structures. The experimental results of the 1D systems agree well with simulations based on dielectric multilayer systems and with FDTD simulations. For a variety of 2D photonic structures the spectrum of the transmitted THz wavepacket reveals the locations and the widths of the lowest order band gaps and the results are in good agreement with the simulations. Acknowledgements. We would like to acknowledge the fruitful collaboration with E.R. Statz and K.A. Nelson and the Swiss National Science Foundation (project number 200021-111693) for funding. 1 2 3 4 5 6 7 8
S.C. Hagness and A. Taflove, Computational Electrodynamics, Artech House, 2000. J.N. Winn, J.D. Joannopoulos and R.D. Meade, Photonic Crystals: Molding the Flow of Light, Princeton University Press, 1995. A.M. Otter, J.P. Korterik, L. Kuipers, N.F. van Hulst, E. Flueck, and M. Hammer, in J. Lightwave Technology 21(1), 1, 2003. H. Gersen, T.J. Karle, R.J.P. Engelen, W. Bogaerts, J.P. Korterik, N.F. van Hulst, T.F. Krauss, and L. Kuipers, in Phys. Rev. Lett. 94(7), 073903-1-073903-4, 2005. T. Feurer, N.S. Stoyanov, D.W. Ward, J.C. Vaughan, E.R. Statz, and K.A. Nelson, in Annu. Rev. Mater. Res. 37, 317, 2007. A. Kleinman, D.H. Auston, in IEEE J. QE 20, 964, 1984. P. Peier, S. Pilz, F. M¨uller, K. A. Nelson, and T. Feurer, in JOSA B, Vol. 25, Issue 7, pp. B70-B75, 2008. D.W. Ward, E.R. Statz, and K.A. Nelson, in Appl. Phys. A 86, 49, 2007.
680
Terahertz wave from coherent LO phonon in a GaAs/AlAs multiple quantum well under an electric field K. Mizoguchi1, Y. Kanzawa2 , M. Nakayama2, S. Saito3 , K. Sakai3 1
2 3
Department of Physical Science, Osaka Prefecture University, Gakuen, Naka-ku, Sakai, 5998531, Japan E-mail: [email protected] Department of Applied Physics, Osaka City University, Sugimoto, Sumiyoshi-ku, Osaka, 5588585, Japan KARC, National Institute of Information and Communications Technology, Iwaoka, Nishi-ku, Kobe, 651-2492, Japan
Abstract. We report the terahertz wave from the coherent LO phonon in a GaAs/AlAs multiple quantum well by applying an electric field. It is found that the intensity of the THz wave from the coherent GaAs-like LO phonon is resonantly enhanced under the condition that the intersubband energy is tuned to the energy of the GaAs-like LO phonon.
Introduction The coherent generation and detection of terahertz (THz) electromagnetic waves generated with ultrashort pulse laser have been widely investigated for applications to spectroscopy, imaging, non-destructive sensing and communication [1]. Recently, the generation of the intense THz waves emitted from coherent GaAs-like longitudinal optical (LO) phonons in GaAs-well layers of GaAs/AlAs multiple quantum wells (MQWs) was reported [2,3]. It is well known that the GaAs-like LO phonon is confined in the GaAs layer and its symmetry is the infrared-active B2 mode. The LO-phonon confinement corresponds to the occurrence of translational symmetry breaking at each GaAs/AlAs interface. When the polarizations due to the coherent LO phonons in respective GaAs layers oscillate in phase, the THz waves generated in respective GaAs layers are constructively superimposed. These are the key points for the generation of the intense THz waves from the coherent LO phonons in GaAs/AlAs MQWs. On the other hand, the enhancement of the coherent LO phonon in GaAs/AlAs MQWs with use of intersubband energy tuning by application of an electric field has been investigated by using a reflection-type pump-probe technique [4]. In the case that the intersubband energy in GaAs/AlAs MQWs is tuned to the LO phonon energy of GaAs (ELO ) by the electric field owing to a quantum confined Stark effect (QCSE), the amplitude of the coherent LO phonon is intensively enhanced in comparison with the amplitude in a low electric field regime. We can expect the enhancement of the THz waves from the coherent LO phonons in GaAs/AlAs MQWs by tuning the intersubband energy to ELO . In this work, we have investigated the THz wave from the coherent LO phonon in a GaAs/AlAs MQW under an electric field.
Experimental Methods The sample used was an undoped GaAs/AlAs MQW embedded in a p-i-n structure
681
grown on a (001) n-GaAs substrate by molecular beam epitaxy, where the p and n layers consist of Al0.5 Ga0.5 As layers. The GaAs/AlAs MQW has periodic heterostructures of (GaAs)44 /(AlAs)16 with 20 periods, where the subscript denotes the number of the monolayers in the constituent layers (thickness of one monolayer = 0.283 nm). We performed photocurrent (PC) measurements at 10 K in order to estimate the transition energies including various higher subbands. The measurements of THz waves were carried out at 10 K by using Ti:sapphire laser pulses with a duration of 40 fs. The pump energy was tuned to 1.55 eV, which was located around the center energy between the E1HH1 and E1HH2 excitons at 110 kV/cm. Here, EiHH j indicates the transition between the electron subband with the quantum number m = i and the heavy-hole subband with m = j. The energy spacing between the E1HH1 and E1HH2 excitons around 110 kV/cm of the electric field becomes close to ELO owing to QCSE. The pump-power density was kept at approximately 0.4 µ J/cm2 . The samples were excited under 45 degree incidence by pump pulses. The THz waves emitted from the samples were collected with a pair of parabolic mirrors and detected by a photoconductive dipole antenna fabricated on a low-temperature-growth GaAs film, where the gap width is 5 µ m. The time-resolved waveforms of the THz waves were obtained by measuring the photocurrent in the dipole antenna and varying the time delay between the pump and gating pulses.
(b)
(GaAs)44/(AlAs)16 MQW
(GaAs)44/(AlAs)16 MQW Electric Field 190 kV/cm
170 kV/cm
Photocurrent (10
150 kV/cm 2 130 kV/cm 110 kV/cm 1
170 kV/cm 150 kV/cm 130 kV/cm 110 kV/cm
90 kV/cm
90 kV/cm
70 kV/cm
0 0
2
4
6
Time Delay (ps)
7
8
9
10
Frequency (THz)
60 50 200
40 LO
150
30
100
20 10
0 11
70
250
50
70 kV/cm 8
(GaAs)44/(AlAs)16 MQW
D
DFT Intensity (a. u.)
3
-10
A)
Exc. Energy:1.55 eV Electric Field Temp.: 10 K 190 kV/cm
(c)300
EE1HH2-E1HH1 (meV)
4
Integral Intensity (arb. units)
(a)
0 40
80
120
160
200
Electric Field (kV/cm)
Fig. 1. (a) Temporal THz waveforms from the (GaAs)44 /(AlAs)16 MQW at various electric fields. (b) Fourier transformed spectra of THz waves emitted from the coherent LO phonon at various electric fields. (c) Electric-field dependence of the integral intensity of THz signals from the coherent LO phonon (open circle). Closed circles indicate the energy spacing between E1HH1 and E1HH2 excitons in the (GaAs)44 /(AlAs)16 MQW plotted as a function of electric field. The broken line indicates the LO phonon energy.
Results and Discussion Figure 1(a) shows the temporal THz waveforms emitted from the (GaAs)44 /(AlAs)16 MQW, and Fig. 1(b) exhibits the Fourier-transformed spectra of the THz waveforms 682
in the whole time region. The THz signal around the time delay of 0 ps is due to the transient photocurrent. The oscillatory THz signals after 0 ps is assigned to the THz wave from the coherent GaAs-like LO phonon, because the observed frequency of the oscillatory THz signal corresponds to ELO . The amplitude and decay rate of the oscillatory THz wave of the coherent LO phonon clearly change with the electric field. Figure 1(b) also indicates that the intensity of the coherent LO phonon in the MQW is obviously enhanced around 110 kV/cm. In order to clarify the electric-field dependence of the THz wave from the coherent LO phonon, the integral intensity of the coherent LO phonon in the frequency region from 8 to 10 THz is plotted as a function of electric field in Fig. 1(c). The closed circles show the variation of the energy spacing between the E1HH1 and E1HH2 excitons (∆EE1HH2−E1HH1 ) estimated from the PC spectra as a function of electric field. The intensity of the THz wave from the coherent LO phonon dramatically changes with the electric field and is resonantly enhanced around 110 kV/cm under the condition that ∆EE1HH2−E1HH1 is tuned to ELO owing to the QCSE. The enhancement factor is over 200 as compared with the intensity of the THz wave in the low electric field region. If this enhancement will originate from the coupling of the coherent LO phonon to the impulsive excitonic interference between the E1HH1 and E1HH2 excitons, the pump-energy dependence of the THz wave from the coherent LO phonon should have a peak around the center energy between the E1HH1 and E1HH2 excitons [4]. However, the pump-energy dependence indicated that the intensity of the THz wave from the coherent LO phonon have a peak at the E1HH2 exciton energy (not shown here). The enhancement of the THz wave from the coherent LO phonon will originate from the double resonance in Raman scattering process[5]. When the incoming resonance at the E1HH2 exciton energy and the outgoing resonance at the E1HH1 exciton energy occur simultaneously, the intensity of the coherent LO phonon will be dramatically enhanced under the condition of ∆EE1HH2−E1HH1 ∼ ELO .
Conclusions We have investigated the enhancement of the THz wave from coherent GaAs-like LO phonon in a GaAs/AlAs MQW under an electric field. It is found that the intensity of the THz wave from the coherent GaAs-like LO phonon is intensively enhanced by a factor of 200 in comparison with that in a low electric field regime, under the condition that the intersubband energy is tuned to the energy of the GaAs-like LO phonon. Acknowledgements. This work was supported by the Ministry of Public Management, Home Affairs, Posts, and Telecommunications, Japan, and by Grant-in-Aid for the Scientific Research from Japan Society for the Promotion of Science. 1 K. Sakai, Terahertz Optoelectronics, (Springer-Verlag, Berlin Heidelberg 2005). 2 K. Mizoguchi, T. Furuichi, O. Kojima, M. Nakayama, S. Saito, A. Syouji, and K. Sakai, Appl. Phys. Lett., 87, 093102 (2005). 3 K. Mizoguchi, A. Mizumoto, M. Nakayama, S. Saito, A. Syouji, K. Sakai, N. Yamamoto, and K. Akahane, J. Appl. Phys., 100, 103527 (2006). 4 O. Kojima, K. Mizoguchi, and M. Nakayama, Phys. Rev. B70, 233306 (2004). 5 B. Jusserand and M. Cardona, in Light Scattering in Solids V, Edited by M. Cardona and G. G¨untherodt, (Springer-Verlag, Berlin Heidelberg, 1989), Chap. 3.
683
Improved Fast Scanning TeraHz Pulse System Bernhard Heinemann1, Colleen J. Fox2, Hermann Harde1 1
Helmut-Schmidt-Universitaet, Holstenhofweg 85, 22043 Hamburg, Germany E-mail: [email protected] 2 Thayer School of Engineering, Dartmouth College, 8000 Cummings Hall, Hanover, New Hampshire 03755-8001 E-mail: [email protected] Abstract. We demonstrate the operation of a fast scanning laser system that was modified to improve and to increase the time resolution as well as spectral width for femtosecond timeresolved optical pump-probe or THz time-domain spectroscopy.
Introduction Both time-resolved optical pump-probe spectroscopy as well as THz time-domain spectroscopy are typically using pump and probe pulses originating from a single laser. One pulse is employed to excite the sample or to generate the THz pulse, while the second pulse is delayed to probe the sample or the THz signal. Commonly the time delay is realized with an optical delay line consisting of a retro-reflector mounted on a mechanical translation stage or a vibrating shaker. These scanning mechanisms, however, have some significant limitations in their scan rates or the achievable time delay. Additional disadvantages may result from vibrations, some lateral shift of the beam or any spot size variation with increasing delay time. These problems are eliminated by asynchronous optical sampling (ASOPS), where two mode-locked lasers serve as pump and probe lasers and are linked to each other at a fixed repetition rate difference ∆f [1-4]. We have applied this ASOPS technique for fast scanning and data acquisition of femtosecond THz pulses.
Experimental Method and Results The central component of the experimental set-up, shown in Fig.1, is a commercial dual-laser system (Gigajet TWIN, Gigaoptics, Germany) which consists of two Ti:sapphire femtosecond oscillators acting at repetition rates f1 and f2 of approximately 1 GHz [2 - 4]. Each oscillator delivers an average output power of about 750 mW with pulses of 70 - 80 fs duration. While laser 2 drives a large area optoelectronic GaAs transmitting antenna to generate the THz pulses, laser 1 is utilized for electrooptic sampling of the THz signal. Both oscillators are linked to each other by an active feedback loop, which stabilizes the repetition rate difference ∆f = f2 – f1 to 10 kHz. To accomplish this, the pulse rate of laser 1, which acts as the master laser, is detected by a fast photodiode and the measured frequency up-shifted by 10 kHz via a frequency shifter. Simultaneously the pulse rate of laser 2 (slave) is registered by a second photodiode and phase locked to the frequency shifted signal of the master laser. As phase detector a double-balanced mixer (DBM) is used and the output of the mixer fed back to a piezoelectric transducer, by which the cavity length of laser 2 and therefore its repetition rate is controlled.
684
Fig.1. Experimental set-up of the fast scanning THz pulse System
The difference rate ∆f determines the scan rate and thus the time TS = 1/∆f = 100 µs which is required for a single scan over the temporal measurement window of 1 ns. Generally the time delay τ between the two pulse combs can be expressed by the relation τ(t) = ∆f/f2⋅t, where t represents the real-time scale. Therefore, any transients on a femtosecond scale are transformed to a nanosecond time scale by the factor f2/∆f = 105, and the other way round any bandwidth limited signals in real-time are converted to the time-delay scale by ∆f/f2 = 10-5. Also the minimum time shift from pulse to pulse and therefore the absolute time resolution ∆τ = ∆f/(f1⋅f2) = 10 fs is derived from this upper relation when choosing t as the pulse period T1 = 1/f1. The transmitted intensity of the probe pulses, which sample the THz signal in an electrooptic crystal, is monitored by a 125 MHz bandwidth photoreceiver and then stored by a fast data acquisition unit as a function of the delay time. Due to the limited bandwidth the receiver integrates over approximately 8 pulses and therefore causes an uncertainty on the time-delay scale of about 80 fs, quite similar to the pulse width of the lasers. The data acquisition unit works with 12,500 channels, so one channel also corresponds to a 80 fs time-delay width. In order to start the recording of data a trigger is generated from a second DBM, which uses the signals of the two fast photo diodes as input. This trigger signal appears periodically with the difference rate ∆f and, therefore, is used to integrate over many scans with the scan rate ∆f. A typical measurement derived with this set-up and averaged over 220 scans (in 400 s) is shown in Fig. 2a, which for clarity only displays the first 20 ps of the THz pulses which are propagating through dry air.
Fig. 2. a) Measured THz pulse, b) Fourier transform spectrum of pulses, measured with the standard set-up, c) Fourier transform spectrum obtained with the modified system (solid line) in comparison to the standard set-up (dashed line)
685
The respective Fourier transform spectrum depicted in Fig. 2b is significantly narrower than expected from the THz pulses and the time resolution of the measurement system. To characterize the system more accurately and to determine its true time resolution, we measured the cross-correlation signal between the two lasers by applying non-collinear sum-frequency generation in a barium-β-borate crystal. The cross-correlation signal is detected with a 200 MHz bandwidth photomultiplier and recorded by a digital sampling oscilloscope. The measured pulses are displayed in Fig. 3a on an expanded scale with the solid line representing a single scan and the dashed graph as the superposition of 512 scans. Obviously is the width of the single cross-correlation scan with 124 fs only determined by the bandwidth of the photomultiplier and the pulse durations, while averaging over many scans significantly broadens the overall pulse structure up to 590 fs.
Fig. 3. a) Cross-correlation signals between laser 1 and 2 shot with electronic trigger. b) Crosscorrelation signal recorded with optical trigger. Solid lines: single scan, dashed graphs: average over 512 scans. The real time scale is converted to delay-time scale by the factor 10-5
Analysis of these measurements immediately makes clear that the quality of the feedback electronics (DBM 1) and the trigger generator (DBM 2) strongly determine the true time delay between the pulses and their appearance with respect to the next scan. So, any jitter or temporal offset from the desired pulse rate causes deviations in the relative time position of the pulse combs, and phase changes of the electronic trigger produce a shift of the whole time scale from scan to scan. To overcome both these effects we replace the electronic trigger by an optical trigger which itself is derived from the cross-correlation signal. This trigger comes up with the right periodicity ∆f and inherently defines the origin of the time-delay scale. Therefore, any shifts of the pulses in their relative and absolute position to each other no longer affect the correct storage for subsequent scans. A cross-correlation signal measured with the optical trigger and averaged over 512 scans (dashed line) is shown in Fig. 3b. Compared to a single scan almost no additional broadening is observed. Using such optical trigger for a THz measurement, (see Fig.1 - lower part) we observe under otherwise similar conditions a distinctly broadened spectrum as represented by Fig. 2c. So, we have demonstrated the operation of a modified laser system that can be used for femtosecond time-resolved optical pump-probe and THz spectroscopy and that allows to take scans over one nanosecond time delay at 10 kHz scan rate with an improved time resolution and increased spectral width. 1 2 3 4
686
P. A. Elzinga et al., in Appl. Opt., Vol. 26, 4303, 1987. C. Janke et al., in Opt. Lett., Vol. 30, 1405, 2005. A. Bartels et al., in Appl. Phys. Lett., Vol. 88, 041117, 2006. A. Bartels et al., Optics Express, Vol. 430, 2006.
Ultrafast photoemission electron microscopy: imaging light with electrons on femto-nano scale Hrvoje Petek1,2 and Atsushi Kubo1,3
1
Department of Physics and Astronomy, University of Pittsburgh, Pittsburgh, PA 15260 USA Donostia International Physics Center, Donostia-San Sebastian 20018 Spain 3 PRESTO, Japan Science and Technology Agency, 4-1-8 Honcho Kawaguchi, Saitama, Japan E-mail: [email protected] 2
Abstract. Attosecond movies (330 as/frame) of surface plasmon polariton dynamics at a nanostructured silver/vacuum interface are recorded with a photoelectron emission microscope employing phase-locked pulse pair excitation. Examples of simple surface plasmon optical elements are given.
Introduction Multidimensional spectroscopic imaging combining laser excitation and electron detection makes it possible to study femtosecond dynamics on the nanometer spatial scale [1-3]. By combining 10 fs laser interferometric pump-probe pulse excitation with photoemission electron microscopy we achieve 330 as/frame, 10 µJ pulse energy for high field physics at multi-megahertz repetition rates T. Südmeyer1, S. V. Marchese1, C. R. E. Baer1, S. Hashimoto1, M. Golling1, A. G. Engqvist1, D. J. H. C. Maas1, G. Lépine2, G. Gingras2, B. Witzel2, and U. Keller1 1
Department of Physics, Institute of Quantum Electronics, ETH Zurich, 8093 Zurich, Switzerland E-mail: [email protected] 2 Centre d’optique, photonique et laser, Université Laval, Pav. d’optique-photonique Québec G1V 0A6, Canada Abstract. We present a modelocked femtosecond thin disk laser that generates pulse energies beyond the 10-µJ limit. We discuss the first photoelectron imaging spectroscopy measurements at multi-megahertz repetition rate, which is advantageous due to high signal-to-noise ratio and reduced measurement time. A maximum peak intensity of 6·1013 W/cm2 was achieved at 14 MHz repetition rate.
Introduction and motivation Numerous applications in science and industry require high energy pulses with durations in the femtosecond regime. The direct generation of such pulses from a multi-megahertz solid-state laser oscillator has significant advantages over more complex amplifier systems, which are typically operating at kilohertz repetition rates with average powers of only a few watts. We present an Yb:YAG thin disk laser generating pulses with a duration of 791 fs and an energy of 11 µJ. By performing photoelectron imaging spectroscopy (PEIS) in xenon and argon, we demonstrate the first application of a thin disk laser in high field science. Driving PEIS with a multimegahertz repetition rate laser results in high signal-to-noise ratio, short measurement time, and high accuracy.
Femtosecond thin disk laser The Yb:YAG thin disk laser head used in this experiment is described in detail in reference 1. Passive mode locking is started and stabilized using a semiconductor saturable absorber mirror (SESAM) 2 with a saturation fluence of 115 μJ/cm2 and a modulation depth of ≈ 0.6%. A set of GTI-type dispersive mirrors lead to a total negative dispersion of ≈ -19900 fs2 per cavity roundtrip, and a glass plate with a thickness of 1 mm inserted at Brewster’s angle ensures linear polarization of the laser output. Operation in helium atmosphere was used to eliminate the large contribution of the air to the nonlinearity inside the laser cavity.3 The increase in pulse energy presented in this paper was achieved by significantly extending the cavity length using a 4f-extension and a MPC.4 The former is realized with two equally curved mirrors, separated by their radius of curvature (R = 2f = 5 m), whilst the latter consists of two flat mirrors and one curved mirror (ROC = 10 m). The total beam path in the MPC is 23.4 m. To inject and extract the beam of the laser cavity into the MPC we use a single flat mirror in front of a flat MPC mirror (Fig. 1b).
747
23.4 m MPC
R = 10 m SESAM
4f-extension Yb:YAG thin disk laser head
GTI-type dispersive mirrors output coupler Brewster plate
AC signal (a.u.)
measured 2 sech -fit
ôp = 791 fs
-2
-1 0 1 time delay (ps)
spectral intensity (a.u.)
Fig. 1. Schematic of the 11-μJ laser setup including a multiple-pass cavity (MPC).
2
measured 2 sech -fit
1.56 nm
1028 1030 1032 wavelength (nm)
Fig. 2. Autocorrelation (left) and optical spectrum (right) of the output pulses. The dashed curves represent a sech2-fit with a duration of 791 fs and an optical bandwidth of 1.56 nm. The time bandwidth product is 0.35 (ideal: 0.315).
All these measures to increase the cavity length resulted in a 4-MHz resonator with an average output power of 45 W corresponding to a pulse energy of 11 μJ.5 The sech2-shaped pulses had a FWHM-duration of 791 fs and a spectral bandwidth of 1.56 nm (Fig. 2), resulting in a time-bandwidth product of 0.35 (Fourier limit: 0.315). A peak power of 12.5 MW was achieved. The beam quality was nearly diffraction limited with an M2 value of 1.1 (measured at 9.4 µJ).
Photoelectron imaging spectroscopy (PEIS) Photoelectron Imaging Spectroscopy (PEIS) is a sensitive method to determine the photoelectron momentum distribution after a multi-photon ionization process.6,7 By focusing a linearly polarized laser pulse with sufficient peak power into a gas target, atoms or molecules can be ionized. PEIS yields a detailed image pattern that gives insight into the excitation path of a photoelectron. Driving PEIS with a multimegahertz repetition rate laser results in a high signal-to-noise ratio, short measurement time, and high accuracy. Increasing the repetition rate by up to 4 orders of magnitude compared with amplified kilohertz systems allows a large number of photoelectrons to be accumulated, while operating at a low number of ionization events per pulse, which avoids space charge effects and results in a close to single atom response. For the PEIS experiments, the output pulses of a passively modelocked thin disk laser running at 14 MHz repetition rate were temporally compressed using a microstructured large mode area (LMA) fiber and a single prism pair as described in reference 8. The laser was operated in air atmosphere with 17 W of average power. Two different compressor setups were used to generate pulse durations of 35 fs and 79 fs respectively, with which we measured photoelectron images from ionization of 748
xenon and argon. An image is obtained by accumulating the electron signals from more than 109 laser pulses.
Fig. 3. Photoelectron imaging spectroscopy setup, and calibration of the peak intensity.
At low laser intensities the ionization is dominated by non-resonant multi-photon ionization (NRMPI) as well as ionization involving the population of excited electronic states of the atom.9 Due to the electric field in the laser focus, the linear ponderomotive shift of high lying Rydberg states enable enhanced multi-photon ionization (REMPI) at specific laser intensities.10 Their appearance in the measured spectra allows a calibration of the peak intensity in the laser focus as a function of the incident pulse energy. To perform this calibration, the appearance of the resonant (11+1)-photon ionization via the 5g state was used, as well as two calibration points obtained from channel closing of the non-resonant 11- and 12-photon ionization, and the zero-energy level.11 A maximum peak intensity of 6·1013 W/cm2 was achieved using the compressed 35-fs pulses.
Summary and conclusions We have demonstrated the first passively modelocked Yb:YAG thin disk laser generating pulse energies beyond the 10-μJ level. These lasers can strongly improve signal-to-noise ratio and reduce measurement time in high field physics applications, as we illustrated by the first PEIS experiments performed at multi-megahertz repetition rate. 1 2 3 4 5 6 7 8 9 10 11
Marchese, S. V. et al. Opt. Lett. 31 (18), 2728-2730 (2006). Keller, U. et al. IEEE J. Sel. Top. Quant. 2 (3), 435-453 (1996). Nibbering, E. T. J. et al. J. Opt. Soc. Am. B 14 (3), 650-660 (1997). Herriott, D. et al. Appl. Opt. 3 (4), 523-526 (1964). Marchese, S. V. et al. Opt. Express 16 (9), 6397-6407 (2008). Helm, H. et al. Phys. Rev. Lett. 70, 3221 (1993). Wiehle, R. et al. Phys. Rev. Lett. 89 (22), 223002 (2002). Südmeyer, T. et al. Opt. Lett. 28, 1951-1953 (2003). Helm, Hanspeter et al. Phys. Rev. A 49 (4), 2726-2733 (1994). Wiehle, R. et al. Phys. Rev. A 67, 063405 (2003). Schyja, V. et al. Phys. Rev. A 57 (5), 3692-3697 (1998).
749
Ultra-high intensity-High Contrast 300-TW laser at 0.1 Hz repetition rate. V. Yanovsky1, V. Chvykov1, G. Kalinchenko1, P. Rousseau1, T. Planchon1*, T. Matsuoka1, A. Maksimchuk1, J. Nees1, G. Cheriaux2, G. Mourou2, and K. Krushelnick1 1
FOCUS Center and Center for Ultrafast Optical Science, University of Michigan, Ann Arbor, Michigan 48109 2 Laboratoire d’Optique Appliqu´ee, UMR 7639 ENSTA,-CNRS-Ecole Polytechnique, F91761, Palaiseau Cedex, France *Current address Colorado School of Mines, Golden, CO 80401 Abstract: We demonstrate the highest intensity - 300 TW laser by developing booster amplifying stage to the 50-TW-Ti:sapphire laser (HERCULES). To our knowledge this is the first Petawatt-scale laser at 0.1 Hz repetition rate.
1. Introduction Recently, we demonstrated intensity as high as 1022 W/cm2 [1] by focusing a 50 TWlaser (HERCULES [2]) into a wavelength-limited spot. Later, the nanosecond-contrast of the laser was improved by 3 orders of magnitude to 1011 [2] . Significant increase of the intensity above 1022 W/cm2 can only be accomplished by increasing the laser energy as there is not much prospect in shortening high-power-pulse duration below 10 fs. Here we report the upgrade of the HERCULES laser to 300 TW output power at 0.1 Hz repetition rate. To our knowledge, this is the first Petawatt-scale laser at high repetition rate. By using adaptive optics and f/1 parabola we focused the output beam into a 1.3 µ focal spot corresponding to unprecedented intensity of ~ 2 1022 W/cm2 . 2. Laser design HERCULES laser design is based on chirped-pulse amplification with cleaning of amplified spontaneous emission (ASE) noise after the first amlifier. Output pulse of the short pulse oscillator of the HERCULES laser is preamlified in the two-pass preamlifier to the microjoule energy level. ASE added by the two-pass amplifier is removed by the cleaner based on cross-polarized-wave generation [2]. High-energy regenerative amplifier [3] and cryogenically cooled 4-pass amplifier bring the pulse energy to a joule-energy level with nearly diffraction-limited beam quality. Two sequential 2-pass-Ti:sapphire amplifiers of 1’ and 2” beam diameter respectively raise the output energy to a value approaching 20 J. We designed our own frequencydoubled Nd:glass pump laser [4] for pumping of the final two amplifiers of the HERCULES laser. The pump laser has two stages of amplification. The frequencydoubled output of the first stage is used for pumping of 1”-diameter T:sapphire amplifier, while the unconverted infrared light is injected into the second stage of the pump laser for further amplification. The frequency- doubled output of the second stage is used for pumping of the booster (2”-diameter) amplifier of the HERCULES laser. The pump laser has a quasi-flat-top beam profile that was achieved at 0.1 Hz
750
repetition rate through relay imaging and thermally-introduced birefringence compensation. The booster two-pass amplifier uses 11-cm-diameter Ti:sapphire crystal. Only a portion of this crystal is used to amplify 2” - diameter output beam of the HERCULES laser. In order to suppress parasitic oscillations the side surface of the crystal is covered with a thin layer of index-matching thermoplastic coating (Cargille Laboratories, Inc.) doped with organic dye absorbing at 800 nm.
a)
b)
Fig.1. Measurements at the output of the booster amplifier. a) Output beam profile of the HERCULES laser; b) Output energy of the booster amplifier in dependence on the pump energy, crosses- experimental data, solid line-Frantz-Nodvic calculations for 1.7 J input energy, dashed line- for 3J input energy
3. Experimental results Output beam profile (Fig. 1(a)) is quasi-flat-top as a result of using flat-top pump beams and of the image relaying of the amplified beam through the whole laser chain. Output energy of 17 J corresponding to 300 TW power after compression has been reached so far (Fig. 1(b)). The pump energy for the booster Ti:sapphire -amplifier (2”diameter) is controlled by changing the pumping level of the oscillator of the pump laser. Because the same oscillator provides seeding pulse for both stages of the pump laser, changing the oscillator energy changes the pump energy for the last two Ti:sapphire amplifiers. It means that not only the pumping energy of the booster amplifier changes but the input energy changes as well if the oscillator energy is changed. As a result, in predicting the performance of the booster amplifier we calculate several Frantz-Nodvic curves, corresponding to varying input energy. Two of them, corresponding to input energy 1.7 J and 3 J are shown in Fig. 1(b). The output pulse is compressed in a 4-grating compressor to ~30 fs. Because the beam size in the compressor is rather large (6”-diameter ) achromatic lenses are used in the final relays to prevent spatially varying group delay across the beam. The autocorrelator that is sensitive to spatial variation of the group delay (autocorrelator with inversion [5]) is used to control this effect. The pulse width is measured at full energy using beam leak through a mirror. The results of the measurements are shown in Fig. 2 (a,b) . The experimental spectrum profile (Fig. 2 (b)) is closely fitted by the Gaussian shape curve of 37 nm FWHM. The pulse-width-bandwidth product (~0.5) is close to the value of 0.44 for a transform-limited ~30 fs Gaussian pulse of 37 nm bandwidth. The final amplifier added no more than 0.2 to the estimated B-integral value of the laser chain. This value is too low to influence the
751
a)
b)
c)
Fig.2. Compressed-pulse measurements. a) Autocorrelation of 300 TW pulse, the experimental autocorrelation picture (insert) demonstrates that there is no amplitude front tilt or other spatial variations of the pulse arrival time; b) Output spectrum and Gaussian fit (FWHM=37nm); c) Focal spot focused by f/1 parabolic mirror
compression that is further evidenced by excellent quality of the compressed pulse (Fig. 2(a)). Although final amplifiers are only water-cooled, the thermal effects in them are minimal as the average absorbed pump power for them is quite modest ( 1 can be easily controlled by the waist diameter w of the input Gaussian beam. In Fig. 1 an example for sub-5 fs pulse generation by bandwidth-limited pulses is presented for fundamental transverse mode excitation only (w = 0.64d). The spectral maximum of the signal pulse is shifted to 153 nm (Fig. 1a) and the spectrum is extended from 142-174 nm (solid curve) with an almost linear increasing phase (dotted curve) in the main part of the spectrum. In Fig. 1b the pulse shapes of all three pulses are presented (note the different scale for the signal pulse). As seen the signal pulse (solid curves) splits into two parts with different velocities. The pulse moving with the linear group velocity at 153 nm has a duration of 11 fs. The other part is locked to the idler at 800 nm and is almost bandwidth-limited with a duration of only 2.5 fs. The spectrum of the free moving pulse is extended from 150 to 157 nm while the locked pulse has a spectrum from 150 to 175 nm. The later is spectrally broadened by cross-phase modulation from the idler pulse. 1
10
0
10
310
I1(t)/I1(0)
0.4
(b)
2.5fs
4
-10 0 10 0.2
2
-1
10
230 140
6
150
160 λ (nm)
170
0 -200 -150 -100 -50 t (fs)
0
50
105xI5(t)/I1(0)
390
φ(λ)
I(λ) (arb. units)
(a)
0
Fig. 1. FWM with bandwidth-limited input pulses. Spectrum (a) and pulse shapes (b) of the signal pulse for a 10 fs 8 µJ input idler at 800 nm, and 50 fs 2.4 µJ pump pulses at 267 nm with a fiber length z = 10 cm, a diameter d = 100 µm, and pressure p = 700 Torr. In (a) the spectrum is shown by the solid line and the phase by the dotted line. In (b), the intensity I(t) for the idler (dotted line) and signal (solid line) is presented. The insert shows the details of the VUV pulse in the vicinity of t = 0.
Let us consider now the case of chirped FWM with broadband streched idler input pulses at 800nm. In particular, we assumed for the idler broadband positively chirped 763
pulses with a pulse energy of 0.1mJ stretched from 3fs to 300 fs pulse duration. For the pump at 273nm unchirped 300 fs pulses with 0.1 mJ energy were assumed. In Fig. 2 one can see the results for VUV pulse generation by chirped FWM with the above described parameters after z = 10 cm propagation. The pulse energy now is 1.4 µJ and the spectrum in Fig.2a (solid curve) reaches from 147nm to 175nm, the spectral phase (dotted curve) shows a negative chirp opposite to the chirp of the idler. It allows compression of the output signal pulse (solid curve in Fig.2b) by a layer of normaldispersion material, here assuming a 1 mm layer made from MgF2 . The dotted curve in Fig. 2b represents the compressed pulse with a pulse duration of 6.0 fs.This method allows a further significant increase of the VUV pulse energy up to 100 µJ by using input idler pulses with about 1 mJ energy.
1 -1 0.1
-2
I5(t)/I1(0)x10-2
12
0 φ(λ)x102
I(λ) (arb. units)
14
(a)
6fs
(b)
10 8
1.6 1.2
-10 0 10
0.8
6 4
I5(t)/I1(0)
1 10
0.4
2 140
150
160 λ (nm)
170
180
0
-300 -200 -100 0 t (fs)
100 200
0
Fig. 2. High energy VUV pulse generation by chirped FWM. Spectrum (a) and pulse shape (b) of the VUV pulse after propagation in a fiber with length z = 10 cm and a diameter d = 100 µm at 1.3 atm for 300 fs 0.1 mJ input idler (800 nm) and pump (267 nm) pulses. The spectral width of the idler pulse corresponds to 3 fs duration. In (a) the spectral phase is shown by the dotted line and the intensity is shown by the solid line. In (b) the temporal intensity distribution of the chirped signal pulse is shown by the solid line and the compressed pulse by the dotted line. The insert shows the details of the compressed 6.0 fs VUV pulse.
Conclusions In conclusion, we studied the potential of four-wave mixing for VUV pulse generation in hollow waveguides with unprecedented short pulse durations. By using bandwidthlimited 10fs input idler pulses at 800nm the generation of 2.5fs VUV nJ pulses at 160 nm are predicted. The pulse energy can be significantly increased by using broadband chirped input idlers. As an example, we predict the generation of VUV pulses with 1.4 µJ energy and 6 fs duration for 300 fs, 0.1 mJ idler pulses. The VUV pulse energy can be increased up to the range of 100 µJ, the pulse duration in this case is about 8 fs. 1 2 3 4 5
P. Baum, S. Lochbrunner, and E. Riedle, “Tunable sub-10-fs ultraviolet pulses generated by achromatic frequency doubling,” Opt. Lett. 29, 1686 (2004). C. G. Durfee III et al., “Intense 8-fs pulse generation in the deep ultraviolet,” Opt. Lett. 24 697 (1999). P. Tzankov et al., “High-power fifth-harmonic generation of femtosecond pulses in the vacuum ultraviolet using a Ti:sapphire lasers,” Opt. Expr. 15, 6389 (2007). A. V. Husakou and J. Herrmann, “Supercontinuum Generation of Higher-Order Solitons by Fission in Photonic Crystal Fibers.” Phys. Rev. Lett. 87, 203901 (2001). I. Babushkin, A. Husakou, J. Herrmann, (to be published)
764
Generation of High Energy Pulses from a Fiberbased Femtosecond Oscillator Jungkwuen An1, Dongeon Kim1, Jay W. Dawson2, Michael J. Messerly2 and Christopher P. J. Barty2 1
Department of Physics, Pohang University of Science and Technology, Pohang 790-784, South Korea E-mail: [email protected] 2 Photon Science and Applications Program, Lawrence Livermore National Laboratory, Livermore, California 94550, USA Abstract. The high energy pulse can be achieved by exploiting self-similar
prapagation regime. In this regime, mode-lock pulse can be generated without dispersive optics such as gratings or prisms in the cavity.
Introduction Generating a single train of high energy pulses in a fiber-based oscillator is a significant experimental challenge because wave-breaking is more likely to happen than the bulk solid-state oscillator due to the inherent high nonlinearity of a fiber. In the case of using nonlinear polarization evolution as mode-locking mechanism, the angular alignment tolerance of the cavity’s wave plates tends to be inversely proportional to the product of the pulse energy and the fiber length. We estimate that for a 25 nJ, 100 fs pulse propagating through a 10 m section of fiber having a modal effective area of 30 μm2, a change in the polarization orientation of the order of a few hundredths of a degree is sufficient to alter the nonlinear phase shift by 2π, allowing an additional pulse stream to circulate. A simple solution to overcome the severe alignment tolerance, is to minimize the length of fiber in the cavity. Note that by reducing the fiber length from 10 to 1 m we increase the alignment tolerance by an order of magnitude. In addition, the GDD of 1 m of fiber is roughly the amount required to form 10 nJ pulses under self-similar propagation regime [1] and thus the cavity should no longer require a grating pair to trim its dispersion.
Experimental Methods The oscillator depicted in Figure. 1 is a unidirectional ring cavity. The core of the Ybdoped, double-clad gain fiber has a diameter of 7 μm and a numerical aperture (NA) of 0.12 (Nufern SM-YDF-7/210). We varied the fiber length from 1 to 2 m. The gain fiber is pumped by a diode laser array having a center wavelength of 976 nm and a maximum output power of 60W; it is coupled to a fiber bundle having a diameter of 400 _m and an NA of 0.22 (LIMO 60-F400-DL976). A pair of dichroic mirrors, which transmit the pump and reflect the 1053 nm oscillator signal, couples the pump into the cavity. The half-wave plate between the isolator and polarizing beam splitter adjusts the fraction of the power that exits the cavity; it was typically set so that 97% exits.
765
2 m Yb-doped fiber
DP DM PBS DM HW
HW ISO HW QW
Fig. 1. Schematic diagram of a grating-less fiber oscillator: HW, half-wave plate; QW, quarterwave plate; ISO, isolator; PBS, polarizing beam splitter; DP, beam dumper; DM, dichroic mirror which is HT for 978 nm and HR for 1053 nm
Results and Discussion We obtained higher pulse energies with longer lengths of gain fiber: 25 nJ for a 2.0 m length and 20 nJ for a 1.2 m length. In both cases higher pump power was available, but if applied resulted in multi-pulsing. Net dispersions of fiber are 0.004 ps2 for 1.2 m length and 0.006 ps2 for 2 m length. Corresponding maximum pulse energies are about 20 and 25 nJ for each, according to the numerical simulation plotted in Fig. 3(a) of [1]. We verified the single-pulse operation by monitoring the long range autocorrelation (600 ps range) and a fast photodiode signal (0.3 ns resolution). We also monitored the stability of the pulse train with a radio-frequency (RF) spectrum analyzer. The RF analyzer always showed a signal-to-noise ratio better than 80 dB RF (40 dB optical) and a linewidth less than 8 kHz which is shown in Figure. 2(b). -20
(b)
(a)
0.8
Power (dB)
Intensity (a.u.)
1
0.6 0.4
-60
80.25 MHz
-100
0.2 0 1020
-140
1040 1060 Wavelength (nm)
-20 -10 0 10 20 Relative frequency (kHz)
Fig. 2. (a) Spectrum of 25 nJ pulses. (b) RF spectrum of the pulse trains.
766
Conclusions We have demonstrated a novel design for a high energy fiber laser that does not contain a grating pair. This design exploits the energy scaling properties of selfsimilar pulses to eliminate the need for dispersion compensation inside the cavity, resulting in a simpler design whose polarization evolution can be readily controlled and maintained. With the grating-less Yb-doped fiber oscillator, 25 nJ pulse energy has been achieved at a repetition rate of 80 MHz, producing 2 W average output power. The pulses were compressible to 150 fs. Further experiments to improve the pulse energy and analysis on the current propagation regime are in progress. Details of this work were described in Ref .2 Acknowledgements. This work was performed under the auspices of the U.S. Department of Energy by the University of California, Lawrence Livermore National Laboratory under Contract No. W-7405-Eng-48. J. An and D. Kim also acknowledge the support of the National Research Laboratory project (No. M1050000006606J0000-06610) funded by Korean Science and Engineering Foundation (KOSEF), Brain Korea 21 project funded by Korean Research Foundation (KRF), and Core Technology Development Program funded by the Ministry of Commerce, Industry and Energy of Korea. 1 2
F. O. Ilday, J. R. Buckley, W. G. Clark, and F. W. Wise in Physical Review Letters, Vol. 92, 213902, 2004. J. An, D. Kim, J. W. Dawson, M. J. Messerly, and C. P. J. Barty, in Optics Letters, Vol. 7, 2010, 2007.
767
Femtosecond passively mode-locked fiber lasers using saturable Bragg reflectors Hyunil Byun, Jason Sickler, Jonathan Morse, Jeff Chen, Dominik Pudo, Erich P. Ippen, and Franz X. Kärtner Department of Electrical Engineering and Computer Science, and Research Laboratory of Electronics, Massachusetts Institute of Technology, 77 Massachusetts Avenue, Cambridge, Massachusetts 02139, USA Email: [email protected] Abstract. We demonstrate a soliton fiber laser with 280-fs pulses at 408-MHz repetition rate, and a stretched-pulse regime fiber laser with 102-fs pulses at 234-MHz repetition rate. Both use saturable Bragg reflectors for mode-locking and/or self-starting.
Introduction Compact sources of femtosecond laser pulses are an attractive and versatile technology for a variety of applications, such as frequency metrology [1] and ultrafast sampling [2]. Passive mode-locking enables low jitter femtosecond pulses and alleviates the need for an external microwave oscillator. In the past, both polarization additive-pulse mode-locking (P-APM) and/or saturable Bragg reflector (SBR) modelocking [3,4] were used. The former has been successfully used in high repetition rate fiber lasers in the soliton [5] and stretched-pulse [6] regimes. The latter can lead to a more compact cavity with fewer components required. Used in combination with PAPM, SBRs can enable self-starting and increase stability when fibers are too short for P-APM alone to self-start, allowing scaling up the repetition rate, while allowing for ultrashort pulse durations. In this paper, we demonstrate two high repetition-rate, self-starting, passively mode-locked femtosecond erbium-doped fiber (EDF) lasers using commercial SBRs. The first is a simple, soliton-regime linear cavity modulated solely by an SBR, and the second is a compact, stretched-pulse-regime sigma cavity modulated by a combination of P-APM and an SBR. The linear cavity soliton source generates a pulse train at up to 408 MHz, with a corresponding full-width half maximum (FWHM) inferred pulse width of 280 fs, while the stretched-pulse sigma cavity results in a pulse train at 234 MHz repetition rate, consisting of 242 pJ, 102 fs pulses.
Experimental Results Linear soliton laser The experimental setup is depicted in Fig. 1. The laser cavity consists of a 25 cm section of EDF with a group-velocity dispersion (GVD) of -20 fs2/mm. One end of the cavity is butt-coupled to an SBR, and the other to a dielectric mirror, which acts as the output coupler. The SBR is a commercial unit with 14% modulation depth, a 2 ps recovery time, and a saturation fluence of 25 μJ/cm2. Pump is provided by a 980 nm laser diode, free-space coupled through a dichroic beamsplitter, and focused by a collimating lens through the output coupler and into the EDF. The output signal follows the same path in reverse, and is separated from the pump by the dichroic
768
beamsplitter. The output pulses are then amplified using an EDFA (980 nm pump, 200mA), detected using a 10 GHz photodiode, and measured with a 500 MHz sampling scope and a signal source analyzer (Agilent E5052).
Fig. 1. Experimental setups of the linear soliton laser.
The 25 cm section of EDF yields a 408 MHz pulse repetition rate. Fig. 2 depicts a sampling scope trace, an optical spectrum, a 2 GHz bandwidth RF spectrum, and phase noise/timing jitter traces, respectively. Integrating the phase noise from 1 kHz to 10 MHz yields a root-mean-squared timing jitter of 196 fs.
Fig. 2. Measurement traces at 400 MHz: a) sampling scope trace, b) optical spectrum, c) RF spectrum, and d) phase noise.
The 40 dB RF side-mode suppression ratio indicates an energy stability to better than 1%. The 9.1 nm FWHM optical bandwidth implies 280 fs duration transform-limited pulses. An autocorrelation measurement is in progress. All measurements were done with 130 mW of cavity-coupled pump power; the intracavity signal power was measured to be 136 mW, resulting in 330 pJ intracavity pulse energies. The laser was self-starting; as the pump power increased, the laser first operated in an unstable Qswitching state, changing to a continuous-wave soliton mode-locked state at pump powers of 115 mW. For pump powers greater than 160 mW, multiple pulsing occurred. We subsequently increased the EDF length to 50 cm, resulting in a 197.8 MHz repetition rate. With 58 mW of cavity-coupled pump power, the intracavity signal power was measured to be 22.4 mW, yielding 113 pJ intracavity pulse energies. The optical signal exhibited a 6 nm FWHM bandwidth, corresponding to a 420 fs transform-limited pulse duration. Here, the continuous-wave soliton mode-locking, and multiple-pulsing threshold pump powers were 50 mW and 60 mW, respectively. The smaller threshold powers (as compared to the 408 MHz setup) come from the fact that a lower repetition rate at the same intracavity power correspondingly increases the pulse energy, and along with it, the nonlinear phase shift. Phase noise measurements yielded a timing jitter of 502 fs from 1 kHz to 10 MHz. Stretched-pulse laser The stretched-pulse laser is shown in Fig. 3a. The cavity is in a sigma configuration to provide a point of reflection for the SBR, and includes an isolator for unidirectional
769
operation, a polarizing beam splitter as the P-APM analyzer and output coupler, and various waveplates to control the polarization evolution. A silicon slab is included in the cavity to prevent residual pump power from reaching the SBR, and to provide normal GVD that, together with the anomalous GVD of the 60 cm of gain fiber, leads to stretched-pulse operation.
a)
b)
c)
d)
Fig. 3. The stretched pulse schematic is shown in a). Laser performance is demonstrated with b) the autocorrelation trace (gaussian = 102 fs), c) optical spectrum (FWHM=35.8 nm), and d) RF spectrum of the detected pulse train.
Fig. 3b-d shows the laser performance. The laser is free-space pumped with 980 nm light, resulting in an average laser output power of 56.7 mW. At the repetition rate of 234 MHz, this corresponds to 242 pJ pulses. The output pulses are normally chirped, and compress to 102 fs using 71.7 cm of single-mode fiber. The optical spectrum FWHM is 35.8 nm, which corresponds to 73 fs transform-limited pulses, indicating the presence of some residual chirp. The clean RF-spectrum indicates single pulse operation as was the case for the linear cavity laser.
Discussion and conclusion We demonstrated two stable, passively mode-locked lasers using SBRs. The first is a soliton laser generating 280 fs pulses at 408 MHz, using 25 cm of EDF as the cavity. Such a design provides a simple, fiber-compatible low-jitter femtosecond source without the need for driving electronics for active modulation. The laser provides good stability in a simple and potentially scalable design. No polarization control or active stabilization is required, and the laser self-starts. Still higher repetition rates can be achieved by further reducing the cavity length while optimizing the pumping scheme. The second is a stretched-pulse laser that produces 242 pJ pulses compressible to at least 102 fs at 234 MHz repetition rate. The SBR enables self-starting, and P-APM provides a strong saturable absorber mechanism, leading to very stable, and most importantly, shorter pulses. The laser offers a higher pulse energy alternative to the linear cavity, and with improved cavity and pump optimization, such a system should also be scalable to higher repetition rates. 1 2 3 4 5 6
770
S. T. Cundiff, in Nature, Vol. 450, 1175, 2007. A. Bartels, R. Cerna, C. Kistner, A. Thoma, F. Hudert, C. Janke, and T. Dekorsy, in Review of Scientific Instruments, Vol. 78, 035107, 2007. H. A. Haus, in Journal of Applied Physics, Vol. 46, 3049, 1975. R. Paschotta, U. Keller, in Applied Physics B, Vol. 73, 653, 2001. J. Chen, J.W. Sickler, E.P. Ippen, and F.X. Kartner, in Conference on Laser and ElectroOptics. OSA, 2007. T. Wilken, T.W. Hansch, R. Holzwarth, P. Adel, M. Mei, in Conference on Laser and Electro-Optics. OSA, 2007.
Noncollinear optical parametric amplification of cw light, continua and vacuum fluctuations Markus Breuer, Christian Homann, and Eberhard Riedle LS für BioMolekulare Optik, Ludwig-Maximilians-Universität München, Oettingenstraße 67, 80538 München, Germany E-mail: [email protected] Abstract: Seed sources for NOPAs are compared. Single-mode cw light renders Fourierlimited femtosecond and fully tunable picosecond µJ output pulses, OPG leads to random spectral fluctuations and a sapphire continuum delivers identical pulses on every shot.
Influence of the seed light on the output of parametric amplifiers Optical parametric amplifiers (OPAs) seeded by a continuum or by parametric generation (OPG) are the prime source of tunable radiation for ultrafast spectroscopy. In chirped pulse OPAs (OPCPAs) unprecedented levels of peak power and pulse shortness are reached and envisioned [1]. White light seeded OPAs are known to deliver reproducible pulses, however, for ultrabroadband pulses the spectrum starts to deviate from a Gaussian shape and weak satellites are found in the temporal structure. Even more severely, OPGs are known for large intensity and spectral fluctuations. In the nanosecond regime superior pulses are generated by amplification of monomode cw seed light in dye amplifiers and optical parametric oscillators [2,3]. For the ultrafast regime the principle has so far not been exploited, likely due to the small number of seed photons contained within the temporal amplification interval. Yet, the ever increasing availability of low cost diode sources would make it an attractive method. We report both a picosecond and femtosecond noncollinear OPA seeded with cw light that display smooth output spectra with Fourier-limited width. From the observed bandwidth for the femtosecond NOPA we conclude that the spectral modulations and pulse-to-pulse variations in an OPG seeded NOPA are evidence that the OPG starts from vacuum fluctuations. In contrast, the output pulses of a continuum seeded NOPA are identical for each shot.
Amplification of cw light in femtosecond and picosecond pumped NOPAs For the evaluation of the different seeding sources under comparable conditions, we have used a BBO-based two-stage OPA with a noncollinear amplifier geometry [4]. Noncollinear phase matching is used in many blue or green pumped systems for its extremely broad amplification bandwidth. In addition, it provides high small signal gain and efficiency and avoids the need for dichroic optics. The first NOPA was pumped by 130 µJ of the 532 nm frequency doubled output of a 10 ps, 5 kHz Nd:YVO4 system. As seed laser we used a fiber coupled single mode cw diode laser system with automated tunability from 1260 to 1630 nm. Due to the low damage threshold of the BBO amplifier crystals for ps pumping, weak focusing and a moderate intensity around 10 GW/cm2 had to be used. The resulting low gain coefficient was compensated by the use of 8 mm BBO crystals, in accord with previous work on OPG/OPA systems [5]. The resulting ps NOPA was tunable over the full tuning range of the seed laser (Fig. 1a) with an output energy of up to 4 µJ. Compared 771
to the mW seeding power, this is a total amplification of 109. Only a slight readjustment of the phase matching angle of the amplifier crystal was needed. Due to the continuous seed light no adjustment of the seed pump delay is needed. Figure 1b) shows the amplified spectra, which result by electronic tuning of the cw-seed laser only, without any adjustment of the phase matching. The spectral bandwidth was only 0.63 nm (at 1.30 µm) with a clean Gaussian distribution very close to the Fourier limit of the 5.3 ps duration. The background due to spontaneous (non-seeded) emission was on the order of only 1 %.
Fig. 1. Typical output spectra of the ps-pumped cw-seeded noncollinear optical parametric amplifier, showing a) wide range tunability and b) fine tunability by electronic adjustment only
Motivated by the success of the cw seeded ps NOPA we also investigated the possibility of cw seeding a femtosecond unit. The 150 fs duration of frequency-doubled Ti:sapphire pump pulses corresponds to just 500 seed photons of the low power 532 nm laser. Nonetheless we were able to observe a µJ output with 88 fs duration and 6 nm bandwidth, corresponding to a time bandwidth product of 0.55 without external pulse compression. The bandwidth of the output pulses is considerably broader than the bandwidth of the seed-light. The additional spectral components originate from the 100 fs amplification. When the 150 fs pump pulse was stretched to 300 fs by the dispersion of a fused silica slab, the bandwidth decreased further. The present implementation of the cw seeded fs NOPA produces about an equal amount of amplified output and amplified parametric superfluorescence. The ratio depends critically on the alignment, in particular the spatial overlap of the beams. In the ps NOPA we also find that the imaging of the fiber output into the amplifier BBO with a high quality aspherical lens is essential for the optimum performance. As it has been recognized that the spontaneous emission of the first amplifier in an OPCPA is a crucial issue [6] and the proper imaging has been shown to improve the contrast ratio [7], we believe that our investigations of low power cw seeding can contribute important fundamental understanding toward the proper design and alignment of the OPCPAs. Indeed, the peak power level of a highly chirped ultrabroadband Ti:sapphire oscillator is not much higher than the seeding level used by us.
Comparison of cw-, continuum- and OPG-seeded NOPAs in the fs-regime When the single longitudinal mode cw laser is used as seed in the fs NOPA, all individual pulses show the same 6 nm wide spectrum (middle row in Fig. 2b). The output pulse energy does, however, fluctuate due to the low number of seed photons and some mode beating of the cw laser. These fluctuations are not found for the ps system
772
since 100 times more seed photons are available over the longer pump pulse duration. A white light seeded NOPA was investigated for comparison and we find fluctuations of only 1 % and again identical, yet slightly structured and non-Gaussian spectra for each shot (top row of Fig. 2b). The width of these spectra can be chosen by varying the chirp of the continuum generated in a sapphire plate. Typically the pulses can be compressed close to the Fourier limit. When neither the cw seed nor the continuum is used, the second amplifier stage is seeded by the parametric superfluorescence generated in the first stage. This is the typical OPG/OPA configuration. The single shot spectral analysis shows that the spectrum is highly structured and varies from shot to shot dramatically (bottom row in Fig. 2b). About 500 shots averaging is needed to converge to a stable and smooth distribution. The width of each individual spectral structure is identical to the one found in cw seeding but the height varies largely (Fig. 2a). We conclude that we amplify individual photons out of the vacuum fluctuations [8] and each amplification process starts at a different depth in the first amplifier crystal.
Fig. 2. a) dots: single shot spectrum of OPG-seeded NOPA; dashed line: Gaussian fits with nearly same width; solid line: fit to amplified OPG-seeded spectrum. b) shot-to-shot comparison of amplified spectra of a fs-NOPA with different seed sources
The OPG/OPA setup makes it possible to directly visualize the vacuum fluctuations. By suitable imaging of the parametric superfluorescence ring from the first noncollinear amplifier, a spatial and spectral blinking is observed on a white card by the bare eye. It is the simultaneous parametric generation and amplification in the BBO crystal at well chosen pump intensity that allows for this startling effect. A detailed analysis will allow the evaluation of the statistics in the OPG and OPA process. 1 2 3 4 5 6 7 8
R. Butkus, R. Danielius, A. Dubieitis, A. Piskarskas, and A. Stabinis, Appl. Phys. B 79, 693 (2004). M. M. Salour, Opt. Commun. 22, 202 (1977). O. Votava, J. R. Fair, D. F. Plusquellic, E. Riedle, and D. J. Nesbitt, J. Che. Phys. 107, 8854 (1997). E. Riedle, M. Beutter, S. Lochbrunner, J. Piel, S. Schenkl, S. Spörlein, and W. Zinth, Appl. Phys. B 71, 457 (2000). X. D. Zhu and L. Deng, Appl. Phys. Lett. 61, 1490 (1992). F. Tavella, A. Marcinkevicius, and F. Krausz, Opt. Expr. 14, 12822 (2006). S. Witte, R. Th. Zinkstok, A. L. Wolf, W. Hogervorst, W. Ubachs, and K. S. E. Eikema, Opt. Express 14, 8168 (2006). R. Glauber and F. Kaake, Phys. Lett. 68A, 29 (1978).
773
Modeling of Octave-Spanning Sub-Two-Cycle Titanium:Sapphire Lasers: Simulation and Experiment Michelle Y. Sander, Helder M. Crespo, Jonathan R. Birge, and Franz X. Kärtner Department of Electrical Engineering and Computer Science and Research Laboratory of Electronics, Massachusetts Institute of Technology, Cambridge, Massachusetts, 02139, USA E-mail: [email protected] Abstract. It is shown that a one-dimensional temporal laser model under optimized intracavity dispersion settings can quantitatively predict the spectral output and temporal pulse shape of octave-spanning, sub-two-cycle Ti:sapphire lasers.
Introduction Various models have been derived to describe the pulse generation in mode-locked lasers [1]. For femtosecond lasers, the dispersion managed mode-locking (DMM) theory [2] takes into account pulse shaping effects of broadening and recompression during each cavity round-trip, based on the sign of group delay dispersion. Simulations based on DMM have been used to accurately describe the behaviour of standard non-octave-spanning mode-locked lasers [3]. To our knowledge, this approach has not yet been demonstrated with sub-two-cycle octave-spanning lasers. In this work, we present an extensive one-dimensional temporal numerical analysis of an actual sub-two-cycle octave-spanning laser. This model incorporates high-order material dispersion and real mirror data from measurements. Our model can capture the octave-spanning output spectrum and generated subtwo-cycle pulse in great detail, allowing for a direct and precise comparison with experimental measurements. Thus, the laser dynamics, stability and steady-state operation output based on realistic parameters can be predicted and optimized.
Laser Model During one cavity round-trip, the pulse shaping is determined by nonlinear propagation through the Ti:sapphire crystal and linear evolution through the air path, mirrors and wedges. The pulse propagation in the crystal is modelled by applying the Split Step Fourier Method to the Nonlinear Schrödinger Equation [4]. The Kerr-Lens mode-locking mechanism is approximated with a fast saturable absorber and the pulse experiences gain saturation with Lorentzian filtering. Properly balancing selfphase modulation, filtering and dispersion (second and higher order material dispersion from the crystal, double-chirped mirrors (DCMs) and wedges), the system evolves to a steady-state solution. Therefore, this approach allows us to accurately follow the temporal and spectral breathing of the pulse within the laser cavity for any general laser configuration. To accurately reproduce the experimental conditions, the DCM pairs are described by the measured reflectivity and group delay data shown in Figs. 1(a) and (b). The DCMs were designed and optimized to provide exact compensation of the material dispersion in the cavity while maintaining a smooth phase over a broad wavelength
774
range [5]. Important features of the mirrors are the pump window around 532 nm (mirrors M1, M3) and a 50% transmission around 1160 nm and 580 nm, corresponding to the 1f to 2f frequency components that are used for carrier-envelope offset (CEO) phase stabilization of the laser. The output coupler with a 2% reflectivity for the main part of the spectrum supports the generation of a broad intracavity spectrum while a reflectivity higher than 50% below 590 nm and above 1040 nm effectively enhances and couples out the spectral wings. 1.0
Group Delay (fs)
0.8
Reflectivity
300
(a) M1, M3 M2, M4 OC
0.6 0.4 0.2 0.0
(b)
250
Designed Measured
200 150 100 50
600
0
800 1000 1200 Wavelength (nm)
600
800 1000 1200 Wavelength (nm)
Fig. 1. (a) Reflectivity curves for the double-chirped mirrors and output coupler; (b) Group delay of the designed and measured DCM pair; measured data was determined by white-light interferometry for a normal incidence angle
Experimental Setup The simulated laser system is a Kerr lens mode-locked Ti:sapphire ring laser with a 2-mm-long crystal and a repetition rate of 500 MHz. The crystal is placed near the focus between two concave DCMs with a 5 cm radius of curvature and the cavity consists of two other flat DCMs (see Fig. 2). A broadband output coupler is coated on a fused silica wedge, which is used in combination with a BaF2 wedge to fine-tune the intracavity dispersion. This configuration simultaneously produces a main carrierenvelope phase stabilized pulse from the reflective output coupler and a 1f-2f output coupled out through the flat mirror M3, used for detection of the CEO frequency in the 1f-2f interferometer [6]. 1f-2f interferometer AOM
DPSS pump laser L
M3 1f-2f output
M1
Ti:S
Phase-locking electronics M2
OC
PR BBO
APD
LO W
S
λ/2
M4
Main output FSW
Fig. 2. Octave-spanning sub-two-cycle, carrier envelope phase stabilized Ti:Sapphire ring laser system [6]
PBS F
IAC / 2DSI
Pulse Dynamics in the Laser The temporal numerical analysis demonstrates the capability of capturing all of the significant laser characteristics: the output spectrum in Figs. 3(a) and (b) shows excellent agreement with the measured data and accurately reproduces the main features in the spectral wings. In addition, we can directly predict the optimized power levels achievable for the 1f and 2f output (with 1160 nm and 580 nm as 1f and 2f frequencies as seen in Fig. 3(b)). As the pronounced ripples in the measured main 775
(a)
-20 -30 Measurement Simulation
-40 -50
600
800
1000
1200
Wavelength (nm)
0
(b)
-10 -20 -30 -40 -50
600
Measurement Simulation 800 1000
1200
Wavelength (nm)
1.0
Pulse Intensity (a.u.)
0 -10
Power Spectral Density (dB)
Power Spectral Density (dB)
output spectrum are assumed to mostly have been caused by etalon filtering effects in the experimental setup, the numerical analysis reproduces the envelope of the measured output spectrum but does not exhibit the same strong oscillations. The importance of optimized dispersion compensation becomes obvious when adjusting the inserted material dispersion by a few fs2. The wings in the spectrum are notably decreased (around -18 dB at 1160 nm for changes in material insertion corresponding to a group delay dispersion of ~7 fs2 at 800 nm), in excellent agreement with the required experimental fine-tuning of the wedges. With external compression of the main output pulse (using four additional octave-spanning DCMs, a BaF2 plate, and two thin FS wedges for fine dispersion compensation), the numerically achievable minimum pulse duration is 4.9 fs. This result exactly reproduces the retrieved pulse from two-dimensional spectral shearing interferometry (2DSI) measurements [7], as shown in Fig. 3(c). Therefore, the numerical analysis clearly establishes that sub-two-cycle pulses can be generated in Ti:sapphire oscillators. (c)
0.8
Retrieved Simulated
0.6
4.9 fs pulse 0.4 0.2 0.0
-20
0
20
40
Time (fs)
Fig. 3. (a) Simulated and measured main output spectrum; (b) simulated and measured 1f-2f output spectrum; (c) retrieved pulse intensity from 2DSI measurements and simulated output pulse after external compression
Conclusions We successfully demonstrated that under optimized dispersion settings, a realistic one-dimensional numerical analysis can accurately describe the laser dynamics of octave-spanning lasers with sub-two-cycle pulses. By optimizing the dispersion compensation we can predict the best achievable performance in terms of output spectrum, frequency enhancement of the wings, and achievable pulse shape and duration in an actual experimental setup, providing a powerful tool for analysis and laser design. Acknowledgements. This work was supported in part by NSF grant ECS-0501478, DARPA grant HR0011-05-C-0155 and AFOSR grant FA9550-07-1-0014. 1 2 3 4 5 6 7
776
H. A. Haus, IEEE, J. Sel. Top. Quantum Electronics 6, 1173, 2000. Y. Chen, F. X. Kärtner, U. Morgner, S. H. Cho, H. A. Haus, E. P. Ippen, and J. G. Fujimoto, J. Opt. Soc. Am. B. 16, 1999, 1999. M. V. Tognetti, M. N. Miranda, and H. M. Crespo, Phys. Rev. A 74, 033809, 2006. G. P. Agrawal, Nonlinear Fiber Optics, Academic Press, Boston, 2001. J. R. Birge and F. X. Kärtner, Appl. Opt. 46, 2656, 2007. H. M. Crespo, J. R. Birge, E. L. Falcão-Filho, M. Y. Sander, A. Benedick, and F. X. Kärtner, Opt. Lett, 33, 833, 2008. J. R. Birge, R. Ell, and F. X. Kärtner, Opt. Lett. 31, 2063, 2006.
Ultra-Broadband Infrared Pulses from a Potassium-Titanyl Phosphate Optical Parametric Amplifier for VIS-IR-SFG Spectroscopy Oleksandr Isaienko and Eric Borguet Chemistry Department, Temple University, 1901 N. 13th Street, Philadelphia, Pennsylvania, 19122, USA E-mail: [email protected] Abstract. A non-collinear KTP-OPA to provide ultra-broadband mid-infrared pulses was designed and characterized. With proper pulse-front and phase correction, the system has a potential for high-time resolution vibrational VIS-IR-SFG spectroscopy.
Many vibrational spectral features, e.g. OH stretching modes, extend over several hundreds of wavenumbers. However, because of phase matching conditions, the bandwidth of IR pulses from most available optical parametric amplifiers is limited to ~150 cm-1. For this reason, the term “broadband Sum-Frequency Generation (SFG) spectroscopy” is applied to the studies in which the infrared pulses have bandwidth of ~150 cm-1 (see, e.g., [1]), with only few accounts of SFG-acquisition over ~500-600 cm-1 bandwidth [2]. Rather than acquiring the spectrum in a single shot, SFG spectra from such systems as, e.g., the air-water interface, are normally obtained by tuning the output of an IR source over the broad spectral feature. Moreover, with ~150 cm -1 pulses the time-resolution of surface vibrational dynamics measurements is limited to ~100 fs, so information about ultrafast processes (e.g., charge transfer, vibrational dephasing) is inaccessible. Recently, there has been a growing interest in the generation of ultra-broadband ultrashort near- and mid-IR pulses via non-collinear optical parametric amplification (NOPA) in nonlinear optical crystals [3, 4]. We applied the concept of NOPA to bulk potassium-titanyl phosphate (KTP) [5]. Calculations of phase matching curves for type-II OPA (o-pump e-signal + o-idler) in XZ-plane of KTP with 800-nm pump, suggested that the combination of a non-collinear signal-pump geometry and the use of a divergent white-light (WL)-seed (rather than collimated) can phase match the near-IR signal in a broad wavelength range [5]. Generation of broadband signal pulses covering simultaneously ~1.05-1.45 m range was achieved experimentally by stretching the pump pulses to compensate for any possible chirp in the WL-seed [5]. The optical setup for generation of broadband near- and mid-IR pulses is shown in Fig. 1, together with the setup for acquisition of idler+800-nm SFG-spectra. In order to obtain the mid-IR idler pulses with sufficient energies, we added a second KTPNOPA stage to amplify the broadband signal generated in the first NOPA-stage, further called “pre-amp seed”. In order to stretch the pump pulses, two equilateral SF18 prisms with face size 2.5 cm were used at optimal geometry [5]. At KTP-1, the internal phase matching angle for pump was set at ~48-49o, the internal non-collinear signal-pump angle was adjusted to ~3.5-4.0o; the angular divergence of the WL-seed was created by slightly displacing the spherical mirror collecting the WL from the sapphire plate. The geometry of the pre-amp seed and pump at KTP-2 was as close as possible to that of the WL-seed and pump at KTP-1.
777
Fig. 1. Optical setup for generation of broadband infrared pulses and (inside the dashed contour) acquisition of the broadband SFG-spectra from silica-water interface: KTP - 2-mm thick crystals ( =42o, =0o); P-ZnSe – 2-mm thick polycrystalline ZnSe crystal; S – 2-mm thick sapphire plate; PBS – polarizer-beam splitter; HW – half-wave plates; SF18 – equilateral prisms; BS – BK7 beam splitter; DL – delay line; CaF2 – 50-mm lens; all other optics - BK7based lenses, flat and spherical mirrors. Orientation of the z- and x-axes is shown. SFG – idler-800nm sum-frequency generation from reference or samples (see text). SHG – secondharmonic generation from the signal pulses. Inset: side-view of the beam geometry at the sample interface; “FS” – IR-grade fused-silica hemicylinder prism; “aq.” – aqueous solution
The spectra of the pre-amp seed and amplified signal were acquired by measuring their SHG-spectra in reflection off a polycrystalline (P-) ZnSe crystal [6], as shown in Fig. 1. This material ensures high-efficiency conversion and has insignificant phasematching restrictions on the converted bandwidth. The spectra of idler from KTP-2 were measured via SFG with 800-nm pulses (Fig. 1) with P-ZnSe placed instead of the sample.
Fig. 2. (a) SHG-spectra of the pre-amp seed (dashed line) and amplified signal pulses at different pump – pre-amp seed relative delays at KTP-2 (thick lines). (b) Spectra of the idler pulses from KTP-2
A spectrum of the pre-amp seed after KTP-1 is shown as a dashed line in Fig. 2(a), along with spectra of the signal after OPA in KTP-2 at different relative time delays between the pre-amp seed and the pump pulses. Although the pump pulses are stretched (from ~180 fs to >500 fs [5]), apparently, it was not enough to compensate for the chirp of pre-amp pulses. Tuning of the signal SHG-spectra excluded 1- or 2photon fluorescence from ZnSe as a source. Idler spectra were derived from idler+800nm SFG on P-ZnSe (Fig. 2(b)). As the entire bandwidth of the pre-amp
778
seed pulses could not be amplified at once, the idler spectra obtained were narrower than one would expect based on the pre-amp seed spectrum. Additionally, to get idler frequencies in higher wavenumber region (>5000 cm-1), we also needed to slightly adjust the pump phase-matching angle within ~1–2o. The energy of the near-IR preamp seed pulse is ~2-3 J, and that of the signal and idler pulses from KTP-2 is on the order of 11-14 J and 4-5 J, respectively, indicating ~3-4% energy conversion. As the idler pulses are obtained in NOPA, they inevitably have a large angular dispersion [7]. Despite this fact, we decided to perform proof-of-principle experiments to check the feasibility of ultra-broadband VIS-IR-SFG spectroscopy on the fused silica - water interface, a system that has vibrational features spanning over ~1000 cm-1. For this, the idler pulses were nearly collimated with a 50-mm CaF2 lens (Fig. 1), and focused tightly at the sample interface with another (not shown). In conclusion, we have demonstrated a KTP-NOPA that generates ultra-broadband near- and mid-IR pulses in ~1.0-1.4 m and ~1.9 – 3 m regions, respectively. Preliminary results on the vis-IR-SFG spectroscopy from the silica-water interface show that ultra-broadband SFG-vibrational spectroscopy of features extending over >1000 cm-1 is feasible. As suggested from the idler bandwidth, generation of sub-20 fs mid-IR pulses should be possible once the pulse-front of the idler pulses is corrected with subsequent compression. Use of a grating -telescope (or prismtelescope) setup ([7]) will provide for compensation of the signal and idler angular dispersion. Additionally, introduction of a larger chirp into the pump pulses should enable production of ~2000 cm-1 broad idler pulses. The NOPA described here will allow the generation of broadband mid-IR pulses for a wide range of applications. Our initial focus will be studies of vibrational dynamics of processes such as heat/energy transfer between different broad modes and/or chemical species with sub-20 fs time resolution. Another interesting perspective is a realization of time-domain ultra-broadband IR spectroscopy with pulses covering simultaneously vibrational transitions of multiple species of interest, in analogy to pulsed NMR spectroscopy. Acknowledgements. The authors acknowledge the support of the US Department of Energy – Office of Basic Energy Sciences. 1. Bonn, M., Ueba, H., and Wolf, M., Theory of sum-frequency generation spectroscopy of adsorbed molecules using the density matrix method - broadband vibrational sumfrequency generation and applications. J. Phys.: Condens. Matter 17(8), S201-S220 (2005) 2. Hommel, E.L., Ma, G., and Allen, H.C., Broadband vibrational sum frequency generation spectroscopy of a liquid surface. Anal. Sci. 17(11), 1325-1329 (2001) 3. Cirmi, G., Brida, D., Manzoni, C., Marangoni, M., De Silvestri, S., and Cerullo, G., Fewoptical-cycle pulses in the near-infrared from a noncollinear optical parametric amplifier. Opt. Lett. 32(16), 2396-2398 (2007) 4. Brida, D., Manzoni, C., Cirmi, G., Marangoni, M., De Silvestri, S., and Cerullo, G., Generation of broadband mid-infrared pulses from an optical parametric amplifier. Opt. Express 15(23), 15035-15040 (2007) 5. Isaienko, O. and Borguet, E., Generation of ultra-broadband pulses in the near-IR by noncollinear optical parametric amplification in potassium titanyl phosphate. Opt. Express 16(6), 3949-3954 (2008) 6. Chinh, T.D., Seibt, W., and Siegbahn, K., Dot patterns from second-harmonic and sumfrequency generation in polycrystalline ZnSe. J. Appl. Phys. 90(5), 2612-2614 (2001) 7. Shirakawa, A., Sakane, I., and Kobayashi, T., Pulse-front-matched optical parametric amplification for sub-10-fs pulse generation tunable in the visible and near infrared. Opt. Lett. 23(16), 1292-1294 (1998)
779
Chirped-pulse Raman amplification for two-color high-intensity laser experiments Peng Dong1 , Franklin Grigsby1 , and Mike Downer1 1 FOCUS
Center, University of Texas at Austin, Department of Physics, Austin, TX 78712, USA E-mail: [email protected]
Abstract. We report generation and compression of millijoule-level first Stokes sideband (873nm) of 800nm TW pulses by inserting a multi-stage barium nitrate Raman shifter-amplifier into a conventional Ti:sapphire chirped pulse amplification system.
Introduction In many high-field experiments it is desirable to accompany the main terawatt (TW) pulse with a moderately powerful (∼0.1 TW), temporally synchronized ultrashort pulse at a slightly shifted (∼100 nm) center wavelength outside the bandwidth of the main pulse. Zhavoronkov et al. [1] demonstrated the generation of 80 µJ, 190fs, sub-GW, 870 nm pulses by stimulated Raman scattering (SRS) of chirped 1.5 mJ, 800 nm pulses in barium nitrate. Because it uses chirped pulses to avoid damage and self-phase modulation in the Raman-active medium, this technique is well suited to TW-scale laser systems based on chirped pulse amplification (CPA), but so far has been demonstrated only on a much lower power system. In this paper, we report scaling the chirped-pulse Raman amplification (CPRA) technique to a TW CPA laser system, resulting in 873 nm pulses up to 3 mJ, 0.03 TW synchronized with 800 nm, 5 TW pulses. This energy was achieved purely by CPRA, without adding any pump lasers to the CPA system and without compromising the amplified energy of the main 800 nm pulses. Further amplification of the 873 nm output by conventional methods e.g. a multi-pass Ti:sapphire amplifier with an additional pump laser prior to compression, appears straightforward. Fig. 1 shows the setup of Raman shifter and amplifier.
Fig. 1. Details of stimulated Raman shifter and two-pass Raman amplifier.
780
Results and Discussion The two main hurdles encountered in scaling CPRA from microjoule [1] to millijoule output energy are: 1) onset of self-focusing and filamentation; and 2) competition between first Stokes and higher-order Raman amplification. The first issue limits output energy of the first (SRS) stage to 1800 nm, and there is some linear response distorting the 1:8 ratio of the IAC. The trace denotes a ~5-cycle (35fs) compressed pulse, nearly compressed to its transform limit (27 fs). We are currently pursuing alternative diagnostics capable of measuring a 2-cycle pulse [6] and optimizing the OPCPA for a larger signal bandwidth. Previous results have already verified the CEP stability of the OPCPA process with self-CEP-stabilized seed pulses generated by DFG [3]. CEP measurements on the presented OPCPA system are also in progress. 1 2 3 4 5 6
788
A. Gordon and F. X. Kärtner, Opt. Express 13, 2941, 2005. F. X. Kärtner, F. Ö. Ilday and W. Graves, MIT Workshop on Seeded FELs, June 17th, 2004, Cambridge, MA. T. Fuji, N. Ishii, C. Y. Teisset, X. Gu, T. Metzger, A. Baltuška, N. Forget, D. Kaplan, A. Galvanauskas, and F. Krausz, Opt. Lett. 31, 1103, 2006. F. Tavella, A. Marcinkevicius, and F. Krausz, New J. of Phys. 8, 219, 2006. J. Moses, C. Manzoni, S.-W. Huang, G. Cerullo, and F. X. Kärtner, Ultrafast Phenomena, Stresa, Italy, 2008. J. R. Birge, R. Ell, F. X. Kärtner, Opt. Lett. 31, 2063, 2006.
Generation of sub-20-fs, two-color deep-ultraviolet pulses by four-wave mixing through filamentation in gases Takao Fuji, Takuya Horio, and Toshinori Suzuki Chemical Dynamics Laboratory, RIKEN, Hirosawa 2–1, Wako, Saitama, 351–0198, Japan E-mail: [email protected] Abstract. Generation of ultrashort pulses at 260 nm and 200 nm by four-wave mixing through filamentation in neon gas was demonstrated. Fundamental (ω1 ) and second-harmonic (ω2 ) pulses of 25 fs Ti:sapphire laser output were focused into neon gas, and 260 nm (ω3 ) pulses were produced by a four-wave mixing process, ω2 + ω2 − ω1 → ω3 , through an ∼15 cm filament. At the same time, 200 nm (ω4 ) pulses were also genereted by a cascaded process, ω3 + ω2 − ω1 → ω4 , and/or by a sum frequency process, ω2 + ω1 + ω1 → ω4 . The both pulses were simultaneously compressed by a grating-based compressor, and characterized by a trangient grating frequencyresolved optical gating. The estimated pulse widths of the 260 nm and 200 nm pulses were 14 fs and 16.5 fs, respectively.
For many experimental studies in fundamental chemical dynamics, ultrashort pulses (∼10 fs) in deep-ultraviolet (DUV, 200∼300 nm) region are required since a number of small molecules have resonances in the wavelength region and the motion of important atoms (O, C, etc.) during chemical reaction is 10 fs time scale. In particular, a light source which generates synchronized twin femtosecond pulses at different wavelengths in the DUV region is very useful for pump-probe photoelectron spectroscopy [1]. One interesting method to generate ultrashort DUV pulses with different colors at the same time is to use third order cascaded processes in a gas-filled hollow waveguide [2]. By mixing the second harmonic (2ω ) and the fundamental (ω ) of powerful Ti:sapphire laser pulses in noble gas, one can generate 3ω , 4ω , and 5ω pulses all together with the cascaded third order nonlinear processes. However, the compression was successful only for the 3ω (267 nm) component. The compression for shorter wavelength components was too difficult because of the large nonlinear dispersion of air in the wavelength region and the power of the generated pulses ( 1 (n < 0) denotes frequency upconverted (downconverted) pulses. The central frequency of the total spectrum and the separation between sidebands can be adjusted by tuning the pump frequencies. The analytical solution for CFWM assuming perfect phase-matching and constant coupling coefficients [7] is formally equivalent to that obtained for molecular modulation, and it can be shown that the sign of chirp in the total CFWM field depends on the initial phase difference between pumps.
795
Experimental setup and results The experimental setup is shown in Fig. 1(a) and includes generation of the CFWM beams, pulse synthesis by beam recombination, and diagnostics. In this proof-ofprinciple experiment, like in previous work [5], two synchronized visible laser pulses from a dual-wavelength (615 and 563 nm) 10 Hz amplifier were used as pumps. The noncollinear horizontally-polarized orange and green pulses (90 fs, 32 µJ and 45 fs, 38 µJ, respectively) are focused in a 150-µm-thick fused silica slide FS1, producing 2 frequency downconverted pulses and a fan of 21 upconverted pulses (with wavelengths down to 200 nm) as shown in Fig. 1(b). The total energy in the cascaded beams is ∼ 6 µJ (10% efficiency). The corresponding 2-octave spectrum measured with a calibrated fiber-coupled spectrometer (250-1100 nm) is shown in Fig. 1(c). (from laser)
(b)
PG‐XFROG delay
P2
P3
D GTP
BS
FS2
λ/2 D
L
(a)
Screen
Cascaded FWM beams FS1
Pump synch.
Log Normalized Intensity
GTP
GTP (to FCS)
P1
P4
1 0.1
(c)
0.01 1E-3 1E-4 300
400
500
600
700
800
Wavelength (nm)
Fig. 1. (a) Experimental setup: GTP, Glan-Thomson Polarizers; D, diaphragms; BS, ultrafast beamsplitter; P1-P4: Al-coated off-axis parabolic mirrors; FS1-FS2: fused-silica slides; FCS, fiber-coupled spectrometer. (b) Image of the generated cascade projected onto a white screen and (c) corresponding measured spectrum. The white arrows denote the pump beams.
Two Al-coated off-axis parabolic mirrors P1 and P2 collimate and focus the angularly separated beams in a second slide FS2; the residual pumps are attenuated with an apertured screen. Simultaneous phase-matching and ensures that the CFWM pulses are phase-locked; the short air path between the two slides provides a small amount of second-order dispersion while minimizing higher-order dispersion, and the λ/8 mirror quality minimizes phase distortions. Hence a train of white-light ultrashort pulses is expected to be synthesized at the focal plane of P2. To characterize the pulses, we used polarization gating induced in slide FS2 by a portion of the orange pump. A second set of parabolic mirrors P3 and P4 sends the gated pulses through a GlanThompson analyzer and into the fiber-coupled spectrometer. Polarization gating cross-correlation frequency resolved optical gating (PG XFROG) traces were obtained by averaging 10 gated spectra for each time delay [Fig. 2(a)].
700 -100
0 100 Delay (fs)
0.3
0.6
0.2
0.4
0.1
0.2 0.0
1.0
0.0 -50
0
Time (fs)
50
-0.1 100
10
(c) 5
0.8 0.6
0
0.4 -5 0.2 0.0 -60
-40
-20
0
20
40
60
Temporal phase (rad)
600
1.2
0.4
Intensity (a.u.)
500
0.8
0.5
Temporal phase (rad)
400
(b)
1.0
(a) Intensity (a.u.)
Wavelength (nm)
300
-10
Time (fs)
Fig. 2. (a) Measured PG XFROG trace. (b) Retrieved first frequency upshifted pulse (30.6 fs FWHM duration; 531 nm center wavelength). (c) Synthesized pulse in the time domain.
796
The measured traces reveal multiply synchronized CFWM pulses. Due to the practically instantaneous nature of the nonlinear response and the low dispersion of the experimental setup, each CFWM order is also near-transform-limited. Since the relative phase between the pumps is not stabilized, the synthesized pulses will still be randomly time-shifted and their duration will change from shot to shot. This, together with the use of a long gate pulse, results in smearing of fine temporal structure in the measured traces. The complete electric field of each of the 15 gated cascaded orders can nevertheless be retrieved unambiguously, as illustrated in Fig. 2(b) for the first frequency upconverted order. By assuming zero relative phase between each of the generated CFWM pulses (equivalent to introducing small, fixed time delays between pulses) we obtain a synthesized field composed of a main 2.2 fs pulse - less than 1.3 optical cycles at the center wavelength of 513 nm - and 2 smaller side pulses at a distance of 23 fs (the pump beat period), as shown in Fig. 2(c). The energy in the main pulse is ∼ 10 µJ, corresponding to an instantaneous power of ∼ 4 GW.
Conclusions A technique to Fourier-synthesize high-power near-single-cycle pulses from multiband coherent spectra in the visible-UV generated through cascaded four-wave mixing was presented. Broadband PG XFROG measurements show the possibility of synthesizing 2.2 fs 1.3-cycle pulses with multi-gigawatt powers without significant manipulation of the intermediate cascaded beams. The use of more stable pump pulses and shorter (∼ 25 fs) gate pulses provided by a Ti:sapphire laser amplifier coupled to an optical parametric amplifier or hollow-fiber compressor is expected to significantly improve the generation and measurement of near-single-cycle pulse trains and isolated single-cycle pulses using this technique. Acknowledgements. This work was partly supported by the Access to Research Infrastructures activity in the Sixth Framework Programme of the EU (contract RII3CT-2003-506350, Laserlab Europe). The authors gratefully acknowledge J. Etchepare and G. Mourou for fruitful discussions, and A. dos Santos for technical support. 1 2 3 4 5 6 7
G. Sansone, E. Benedetti, F. Calegari, C. Vozzi, L. Avaldi, R. Flammini, L. Poletto, P. Villoresi, C. Altucci, R. Velotta, S. Stagira, S. De Silvestri, M. Nisoli, “Isolated singlecycle attosecond pulses,” Science 314, 443 (2006). E. Matsubara, K. Yamane, T. Sekikawa, and M Yamashita, “Generation of 2.6 fs optical pulses using induced-phase modulation in a gas-filled hollow fiber,” J. Opt. Soc. Am. B 24, 985 (2007). T. Binhammer, E. Rittweger, U. Morgner, R. Ell, and F. X. Kärtner, “Spectral phase control and temporal superresolution toward the single-cycle pulse,” Opt. Lett. 31, 1552 (2006). M. Y. Schverdin, D. R. Walker, D. D. Yavuz, G. Y. Yin, and S. E. Harris, “Generation of a single-cycle optical pulse,” Phys. Rev. Lett. 94, 033904 (2005). H. Crespo, J. T. Mendonça, and A. Dos Santos, “Cascaded highly nondegenerate fourwave mixing phenomenon in transparent isotropic condensed media,” Opt. Lett. 25, 829 (2000). L. Misoguti, S. Backus, C. G. Durfee, R. Bartels, M. M. Murnane, and H. C. Kapteyn, “Generation of broadband VUV light using third-order cascaded processes,” Phys. Rev. Lett. 87, 013601 (2001). J. T. Mendonça, H. Crespo, and A. Guerreiro, “A new method for high harmonic generation by cascaded four-wave mixing,” Opt. Comm. 188, 383 (2001).
797
2 MHz repetition rate - 15 fs fiber amplifier pumped optical parametric amplifier Steffen H¨adrich1 , Jan Rothhardt1 , Fabian R¨oser1 , Damian N. Schimpf1 , Jens Limpert1 , and Andreas T¨unnermann1,2 1 2
Friedrich Schiller Universitt Jena, Institute of Applied Physics, Albert-Einstein-Str. 15, 07745 Jena, Germany Fraunhofer Institute for Applied Optics and Precision Engineering, Albert-Einstein-Str. 7, 07745 Jena, Germany E-mail: [email protected]
Abstract. An optical parametric amplifier pumped by a fiber amplifier producing ultrashort pulses with durations of 15.6 fs at 2 MHz repetition rate is presented together with scaling considerations to tens of µJ pulse energy.
Introduction The availability of ultrashort and high peak power optical pulses has driven breathtaking advances in applications ranging from industrial to fundamental science. Unfortunately, a number of ultrafast processes initiated by such pulses are characterized by a low probability or a low conversion efficiency. Hence, detection systems comprise very sophisticated and sensitive apparatus precluding these from real world applications. An increase of few orders of magnitude in the repetition rate could dramatically decrease the necessary measurement times. It is well known that fiber laser systems are average power scalable due to the fiber design itself, and femtosecond fiber amplification systems have been demonstrated with average powers above 100 W and pulse energies up to 1 mJ [1]. Though, their gain bandwidth does not support pulse durations of few 10 fs [2]. On the other hand optical parametric amplifiers (OPA) offer an enormous amplification bandwidth [3], up to 200 THz in a non-collinear configuration (NOPA), and are inherently immune against thermo-optical problems due to the fulfilled energy conservation during the nonlinear amplification. Furthermore, a high gain can be achieved in just few millimeter long crystals, therefore, the B-integral (accumulated nonlinear phase) is negligible. Based on these facts, OPA is a promising way in order to generate high peak-power pulses at high repetition rates. Our approach involves the transfer of the high pulse energy, at high average power, of a 1 µm laser source to pulses of significantly shorter pulse durations via parametric amplification.
Experiment and Results The experimental setup is shown schematically in fig. 1. A cavity dumped Ti:Sapphire oscillator providing 25 fs , 15 nJ pulses at 810 nm wavelength is used to seed the NOPA and the fiber based amplification stage as well. To generate a signal for the fiber amplification stage pulses with an energy of 1 nJ are coupled into a photonic crystal fiber with a zero dispersion wavelength of 975 nm providing soliton generation. The coupled pulse energy is carefully aligned to generate a single soliton at 1030 nm center wavelength and a few pJ of pulse energy and a spectral bandwidth of more than 30 nm. These pulses are pre-amplified in a double pass 6 µm core and a single pass 10 µm core Yb
798
doped fiber amplifier to 360 mW average power. Approximately 40 m of delay fiber are used within the first double-pass amplifier to ensure temporal overlap with the next signal pulse in the NOPA (at 2 MHz repetition rate).
. Fig. 1. Experimental setup of the fiber amplifier pumped NOPA.
Before further amplification, the pulses are stretched to about 320 ps and spectrally cut to 6 nm bandwidth using a grating-stretcher (1740 lines/mm). The power amplifier is operated in a single-pass configuration using a polarizing-large-mode-area doubleclad photonic crystal fiber. After amplification the pulses are compressed by a grating pair (1740 lines/mm) to a pulse duration of 640 fs with a pulse energy of 10 µJ. The infrared pulses are frequency doubled in a 2 mm critical phase matched LBO crystal with 41 % conversion efficiency, resulting in a 4.1 µJ, 515 nm pump source for the NOPA. With a pump-signal angle of 2.6◦ broadband phase matching and therefore broadband amplification can be obtained. The Ti:Sapphire laser pulses are parametrically amplified either directly or after additional spectral broadening. Direct amplification leads to 500 nJ pulses which are compressed in a simple fused silica prism compressor (efficiency 90 %) to high quality pulse with 20.1 fs width and a spectral bandwidth of 59 nm resulting in a peak power of 20 MW. Additional spectral broadening in a 1.5 cm photonic crystal fiber results in an increased bandwidth of 124 nm (fig. 2). Parametric amplification of these pulses results in 300 nJ pulse energy and a pulses duration of 15.6 fs (fig. 2), which is only slightly above the transform limit (13.2 fs).
. Fig. 2. (a) Normalized spectrum of the Ti:Sapphire oscillator (gray) , the spectrally broadened pulse (dotted) and amplified spectrum (black). (b) Autocorrelation trace of the compressed pulses (black) and fourier transformation of the measured spectrum (dotted).
799
Scaling Considerations Recent demonstration of a mJ fiber-chirped-pulse amplification system [1] has proven the capability of scaling the output of this NOPA experiment to much higher pulse energies at moderate repetition rates. First proof of principle experiments revealed that the necessary pump pulse generation for the parametric amplification can be achieved by frequency doubling the output of the fiber CPA system at repetition rates as high as 200 kHz with a pulse energy of 205 µJ (41 W) and a pulse duration of 720 fs (fig. 3) .
. Fig. 3. Autocorrelation trace of the SHG signal at 205 µJ pulse energy and 200 kHz repetition rate.
Using the same experimental configuration as shown in fig.1 combined with the high energy fiber CPA system offers the possibility to generate ultrashort pulses (20 fs) and pulse energies of tens of µJ at a repetition rate of 200 kHz.
Conclusions A new approach for parametric amplification that combines the advantages of a broadband Ti:Sapphire oscillator and a high average power Yb-doped-fiber-amplifier system is presented. Using the frequency doubled output pulses of a fiber amplifier with up to 4.1 µJ pulse energy, efficient parametric amplification of the Ti:Sapphire oscillator pulses is possible. Output energies up to 500 nJ were achieved with a 5 mm long BBO amplifier crystal. The shortest pulse duration of 15.6 fs was obtained with additional spectral broadening in a photonic crystal fiber. The scalability of this approach to pump pulse energies of 205 µJ at 200 kHz repetition rate has been proven with the help of a state-of-the-art fiber-chirped-pulse amplification system and is currently under investigation experimentally. Acknowledgements. This work has been partly funded by the German Federal Ministry of Education and Research (BMBF) with project 03ZIK455 ’onCOOPtics’. 1
F. R¨oser, T. Eidam, J. Rothhardt, O. Schmidt, D.N. Schimpf, J. Limpert, and A. T¨unnermann, Opt. Lett. 32, 3495-3497, 2007. 2 R. Paschotta, J. Nilsson, A. Tropper, and D. Hanna, IEEE J. Quantum Electron . 33, 1049-1056, 1997. 3 G. Cerullo and S. Silvestri, Review of Scientific Instruments 74, 1, 2003.
800
Octave-wide tunable NOPA pulses at up to 2 MHz repetition rate Christian Homann, Christian Schriever, Peter Baum, and Eberhard Riedle LS für BioMolekulare Optik, Ludwig-Maximilians-Universität München, Oettingenstraße 67, 80538 München, Germany E-mail: [email protected] Abstract: Based on noncollinear parametric amplification, we demonstrate frequency conversion of the 230 fs pulses of a high repetition rate ytterbium-doped fiber amplifier system to octave wide tunable femtosecond pulses with down to 20 fs duration.
Complete spectral coverage for ultrafast spectroscopy For numerous applications in ultrafast spectroscopy and nonlinear microscopy, spectrally tunable sources of intense femtosecond laser pulses are needed. Fiber amplifiers offer high peak power together with a high repetition rate, high energy stability and excellent beam quality. However, they so far only provide pulse durations above 100 fs and lack flexibility in output wavelength. Therefore a device is desirable that can convert the fiber amplifier pulses to an as wide as possible wavelength range, ideally in combination with temporal shortening to the 20 fs regime. Here we present a noncollinear optical parametric amplifier (NOPA [1]) that converts the 1035 nm, 230 fs output pulses of an ytterbium-doped fiber amplifier laser system to pulses with continuous tunability from 440 to 990 nm [2]. Pulse durations of down to 19.8 fs and pulse-to-pulse energy stability of 1.3% (rms) at up to 2 MHz repetition rate are demonstrated. The NOPA is pumped with UV pulses from the third harmonic of the fiber amplifier at 345 nm and is seeded by a spectrally smooth supercontinuum generated in sapphire. The leftover second harmonic light from the frequency tripling process is used to pump an additional independently tunable NOPA with a tuning range of 600 to 970 nm. Together the two NOPAs provide powerful sources for two-color pump-probe spectroscopy at MHz repetition rates.
Fig. 1. Typical output spectra of the 345 nm pumped noncollinear optical parametric amplifier, showing the more than octave wide continuous tuning range
801
Interference experiments show that the two NOPAs have a precisely locked relative phase, despite of being pumped by different harmonics with random phase jitter. This directly proves that parametric amplification preserves the phase of the seed light.
Octave-wide tunability with femtosecond UV pumping As primary pump source we use a commercially available ytterbium-doped fiber oscillator/amplifier system (IMPULSE; Clark-MXR, Inc.) that delivers 10 µJ output pulses with 230 fs duration at 1035 nm and at a variable repetition rate between 200 kHz and 2 MHz. Pulses with an energy of 1.5 µJ are split off and focused into a 4 mm thick rotating sapphire disc for supercontinuum generation. The rotation is needed to prevent accumulative damage at high repetition rates. The remaining part of the output pulses is used to generate the NOPA pump pulses. We use type I frequency doubling and subsequent type II sum-frequency mixing of the fundamental and second harmonic pulses in specially selected BBO crystals to generate the third harmonic at 345 nm by collinear propagation only. Due to the group velocity mismatch, the second harmonic pulses leave the first type-I BBO crystal later than the transmitted fundamental pulses. With the particular choice of type-II sum frequency mixing in the second BBO crystal, the sign of the group velocity mismatch is reversed and the two pulses restore perfect temporal overlap inside the crystal for efficient conversion (total 15%). The third harmonic and the remaining second harmonic are each separated from the left over fundamental with a dichroic mirror and used to pump two independently tunable NOPA units. Fig. 1 shows output spectra from the 345 nm pumped NOPA unit. The output pulses are continuously tunable over more than one optical octave (440 to 990 nm). Autocorrelation measurements yield pulse durations of about 2030 fs between approximately 500 and 700 nm. The smooth Gaussian shaped spectra allow for pulse shapes without temporal satellites, as observed in the autocorrelation traces. Towards the short and long wavelength side of the tuning range, the Fourier-Limits of the measured spectra increase. For short wavelengths, this is the result of the strong dispersion of the seed light and can be overcome in future experiments by precompression of the seed light. A calculation of the phase matching bandwidth lets us expect 13 fs blue pulses. On the long wavelength side, above the NOPA’s degeneracy point at 690 nm, the group velocity mismatch between signal and idler cannot be compensated for by noncollinearity. However, the secondary NOPA unit operates with visible pump pulses and renders pulse durations of ~10-30 fs all the way up to 970 nm [3]. The output pulse energies are in the range of 150 nJ for the 345 nm pumped NOPA, best at the center of its tuning range. Simultaneously, the second harmonic pumped NOPA delivers pulses with up to 250 nJ energy. These pulse energies are well sufficient for frequency doubling and thereby extending the available spectral range to the ultraviolet region. The demonstrated octave-wide tuning range together with a single additional nonlinear conversion will lead to a gapless coverage from below 250 nm to nearly 1 µm. The high output stability of our setup is evident in pulse-to-pulse energy fluctuations of less than 1.3% rms, measured by recording sets of 100.000 subsequent single pulse energies.
Investigation of phase dependencies in optical parametric amplification The overlapping amplification regions of the two differently pumped NOPA units allow for an instructive interference experiment that renders direct information about 802
the phase dependencies in the parametric amplification process. The two NOPA units, seeded by the same supercontinuum, are tuned to the same center wavelength of 720 nm and brought to interference with a small angle on a distant screen (see Fig. 2a). In earlier experiments it was shown that such an arrangement leads to stable interference fringes, when the NOPAs were seeded with the same supercontinuum and when pump pulses with similar phase fluctuations were applied [4]. In contrast, the presented experiment involves pump pulses that are derived from different harmonics of the primary fiber laser system, which is not phase stabilized. The second and third harmonic pulses therefore have carrier-envelope phases with twice or threefold the original phase fluctuations, which makes their relative phase jitter at random.
Fig. 2. Interference experiment to investigate phase dependencies in optical parametric amplification (OPA). a) Experimental setup. SCG: supercontinuum generation, BS: beam splitter. b) Measured spatial interference pattern. c) and d) Phase of the interference pattern over time
Nevertheless we observe a very stable interference pattern of the NOPA outputs (Fig. 2b), which, as can be seen in Fig. 2c) and d), shows very low residual phase fluctuations of less than 20 mrad rms (0.1-1000 Hz). This demonstrates that optical parametric amplification, despite possible saturation effects, preserves the phase of the seed light independently of phase fluctuations of the pump pulses to an extreme precision. In combination with the ultrashort pulse durations, the octave-wide tunability, and the high repetition rate, the presented system has potential to significantly advance nonlinear microscopy and ultrafast spectroscopic applications. 1 2 3 4
T. Wilhelm, J. Piel, and E. Riedle, “Sub-20-fs pulses tunable across the visible from a blue pumped single pass noncollinear parametric converter”, Opt. Lett. 22, 1494, 1997. C. Homann, C. Schriever, P. Baum, and E. Riedle, “Octave wide tunable UV-pumped NOPA: pulses down to 20 fs at 0.5 MHz repetition rate”, Opt. Express 16, 5746, 2008. C. Schriever, S. Lochbrunner, P. Krok, and E. Riedle, “Tunable pulses from below 300 to 970 nm with durations down to 14 fs based on a 2 MHz ytterbium-doped fiber system”, Opt. Lett. 33, 1, 2008. P. Baum, S. Lochbrunner, J. Piel, and E. Riedle, “Phase-coherent generation of tunable visible femtosecond pulses”, Opt. Lett. 28, 185, 2003.
803
Asymptotic pulse shapes and pulse self-compression in femtosecond filaments Carsten Kr¨uger1,2 , Ayhan Demircan1 , Stefan Skupin3 , Gero Stibenz2 , Nickolai Zhavoronkov2 , and G¨unter Steinmeyer2 1 2 3
Weierstraß-Institut f¨ur Angewandte Analysis und Stochastik, Mohrenstr. 39, 10117 Berlin,Germany, e-mail: [email protected] Max-Born-Institut, Max-Born-Straße 2a, 12489 Berlin, Germany, e-mail: [email protected] Max-Planck-Institut f¨ur Physik komplexer Systeme, N¨othnitzer Str. 38, 01187 Dresden, Germany, e-mail: [email protected]
Abstract. The balance of Kerr-type and plasma-mediated self-amplitude modulations can give rise to self-stabilizing asymptotic pulse shapes in filament propagation. These soliton-like solutions resemble experimental data and constitute the major mechanism for self-compression in femtosecond filaments.
Self-guided propagation of femtosecond pulses in filaments has found novel applications, including detection of atmospheric pollution, terahertz generation [1], and fewfemtosecond pulse generation in the VUV [2]. One of the most surprising effects, however, is the self-compression of millijoule pulses in the filamentary channel [3]. Using input pulses of 2.6 mJ input energy and 45 fs duration, a sixfold compression down to 7.3 fs duration has been achieved in argon, without any need for external dispersion compensation schemes, see Fig. 1. Such a source is highly interesting for applications in high-field physics, because it also removes energy constraints imposed by guiding fibers as they are used in hollow fiber compressors. Consequently, filament self-compression allows for the tightest concentration of optical energy at kHz repetition rates. Elaborate numerical models have been developed for an analysis of the self-compression scenario [4], and these models reproduce most experimental findings very well. However, the exact origin of stable self-compressing light bullets is still not completely understood.
Fig. 1. (a) Experimentally measured pulse shapes generated by self-compression in an argon filament. (b) Numerically simulated sequence of pulse shapes illustrating the pulse shaping mechanism during propagation in the filament.
It is rather undisputed that a balance between Kerr-type self-focusing and plasmainduced self-defocusing causes self-stabilization of the beam profile, giving rise to a stable beam profile with a typical diameter of 200 microns, the longitudinal extension of which can easily exceed the confocal parameter by a factor 5 to 10. Here we in804
troduce an unconditionally stable pulse shape arising from a similar balance along the direction of propagation. The resulting pulses display the same characteristic asymmetry as observed in the experiments, with a slowly rising edge and a steep trailing edge. These temporal pulse profiles are also observed in numerical simulations, see Fig. 1(b). The balance of nonlinear effects can compensate for small perturbations, making these pulse shapes an asymptotic solution in filamentary propagation. Similar to solitons in nonlinear fiber optics, such pulse shapes automatically appear once a certain threshold is reached. Excess energy is stripped off into the spatial reservoir surrounding the filament, similar to the transfer into the temporal continuum in fiber soliton propagation. It is important to understand that the formation of a stable filament is a condition that requires a balance of focusing and defocusing effects in every temporal point of a propagating pulse. Kerr-type self-focusing is an instantaneous effect in noble gases. Plasma-induced defocusing, however, will monotonically increase over the pulse as recombination of electrons occurs on much longer time scales than the pulse duration. The balancing condition q2e n2 I(t) = η0 2me ε0 ω02
Zt
w(I(t 0 ))dt 0
(1)
−∞
therefore gives rise to characteristically asymmetric pulse shapes. Here n2 is the nonlinear refractive index, η0 the number density of atoms, ε0 the dielectric constant, ω0 the angular frequency of the light, and w the ionization rate, e.g., as calculated by the Ammosov-Delone-Krainov (ADK) theory. qe and me denote electron charge and mass, respectively. Typical solutions of Eq. (1) are shown in Fig. 2. Solution of Eq. (1) can be ∗ simplified by replacing the ADK ionization rate by a simple power law w = σN ∗ I N with an effective nonlinearity N ∗ ranging from 8 to 9 and a Keldysh parameter σN ∗ . With these simplifications, Eq. (1) can be solved analytically, yielding solutions of the type ∗ I(t) ∝ (−t)1/(1−N ) . Quite clearly, these root-like pulse shapes reflect the balance between an instantaneous and a non-instantaneous nonlinear optical effect, with a slowly varying pedestal-like rising edge and an indefinitely steep falling edge, see Fig. 2. Other than Schr¨odinger solitons in nonlinear fiber optics, stable pulse shapes in filaments are not localized, may exhibit a pole, and show diverging energy. Fig. 2. Characteristically asymmetric asymptotic pulse shapes calculated from Eq. (1) using ADK-theory. Shown are the cases of helium, argon, and xenon. The gray shaded area illustrates the definitions of τ and ε, describing the perturbation used for the stability analysis. The balance of nonlinear optical effects has to transfer the energy in the gray shaded area into the reservoir for any duration lasting longer than the optical cycle.
Despite their awkward mathematical properties, however, the root-like solutions of Eq. (1) are locally stable. Phenomenologically, this is quite easy to understand, as any positive deviation from the balancing solution I(t) will increase the number of 805
electrons generated, cf. Fig. 2. As the ionization rate w is highly nonlinear with I, the resulting additional defocusing will easily outweigh the increased self-focusing effect, which causes an effective transfer of radiation into the reservoir and restores the balancing condition. The same argument holds for a negative deviation as the thresholdlike intensity dependence of plasma generation causes stalling of defocusing effects. With strongly reduced plasma defocusing, energy in the filament will reconcentrate, i.e., compensate the original perturbation. Mathematically, this self-restoration requires 2n2 me ε0 ω02 ∗ ∂w 200 nJ and a pulse duration 150 nJ. With this scheme, a maximum energy of the white-light pulses of 22 nJ could be achieved.
Fig. 3. Generated supercontinuum spectra for output energies ranging from 5 nJ to 20 nJ
Acknowledgements. This work was funded by the Australian Research Council under the Centres of Excellence scheme 1 2 3 4 5 6 7 8
812
A. Fernandez, T. Fuji, A. Poppe, A. Fuerbach, F. Krausz, and A. Apolonski in Optics Letters 29, 1366, 2004. D.Herriott, H. Kogelnik, R. Kompfner in Applied Optics 3, 523, 1964. R.R. Alfano, The Supercontinuum Laser Source, 2nd ed., Springer, 2006. P.S.J. Russell in Science 299, 358, 2003. P. Dombi, P. Antal, J. Fekete, R. Szipoecs, Z. Varallyay in Appl. Phys. B 88, 379, 2007. K.M. Hilligsøe, T.V. Andersen, H.N. Paulsen, C.K. Nielsen, K. Mølmer, S. Keiding, R. Kristiansen, K.P. Hansen, J.L. Larsen in Opt. Express 12, 1045, 2004. T.P. White, B.T. Kuhlmey, R.C. McPhedran, D. Maystre, G. Renversez, C.M. de Sterke, L.C.Botten in J. Opt. Soc. Amer. B 19, 2322, 2002. G. P. Agrawal, Nonlinear Fiber Optics, 2nd ed., Academic Press, San Diego, 1995.
An All-Optical Synchrotron Light Source H. Schwoerer1,2, H.P. Schlenvoigt2, K. Haupt1, A. Debus2, E. Rohwer1, J. Gallacher3, R. Shanks3, D. Jaroszynski3 1
Laser Research Institute, Physics Department, Stellenbosch University, Private Bag X1, Matieland 7602, South Africa E-mail: [email protected] 2 Institut für Optik und Quantenelektronik, Universität Jena, Max-Wien-Platz, 07743 Jena, Germany 3 Department of Physics, Scottish Universities Physics Alliance, University of Strathclyde, Glasgow G4 0NG, United Kingdom Abstract. We report on the generation of synchrotron radiation from laser accelerated relativistic electrons propagating through an undulator. We indicate that this provides exciting novel opportunities in ultrafast spectroscopy.
Introduction and Motivation Ultrashort coherent light pulses are an invaluable tool to study the microscopic dynamics of matter. Femtosecond lasers in combination with optical nonlinear devices deliver pulses with wavelengths between the UV and the NIR spectral region and are therefore in the range of relevant electronic excitations of molecules and solids. Shorter wavelengths down to a few nm allow a direct view onto the molecular structure through diffraction. They can be generated by synchrotron radiation using electron storage rings or linear accelerators equipped with undulators. In particular if the undulator, operated in the free electron laser mode (FEL), extremely brilliant, ultrashort, polarized and coherent light pulses are produced. Tunability is achieved by varying the electron energy, coherence is obtained by an intrinsically generated modulation of the electron pulses during the interaction with the self-generated light field. This process is called self-amplified spontaneous emission (SASE), which is the basis of all present-day FEL at short wavelengths. The short duration of the emission arises from the temporal structure of the SASE electron pulse itself. Currently, constructed free electron lasers promise a wide applicability spanning from atomic and cluster physics, through temporally resolved structural analysis of complex molecules to plasma physics and even quantum electrodynamics in high external fields [1]. However, nowadays FEL require kilometer long electron accelerators due to the maximum energy gain per length in the order of tens of MeV/m which is set by the material breakdown of the radio frequency cavity. An alternative electron acceleration mechanism is accomplished by the fs lasers themselves: if the output of conventional fs lasers is extended by powerful short pulse laser amplifiers, laser pulses with peak powers of tens of terawatts and peak intensities in the laser beam focus of more than 1020 W/cm2 can be achieved [2]. If these laser pulses correctly interact with a self- or externally generated plasma, electrons can be accelerated to energies up to a GeV with a few percent bandwidth and within a well collimated beam [3]. The underlying acceleration mechanism is called forced wake field or bubble acceleration and relies on the generation of a spatially confined plasma wave, which captures electrons and accelerates them within a few millimeters to several hundreds of MeV. The energy gain per length is significantly larger than in radio frequency accelerators because the acceleration is 813
based on acceleration in a plasma, which has to be avoided in the conventional approach. The electron pulse duration has been measured to be not longer than the laser pulse duration (< 50 fs), simulations suggest that it might be even shorter. We report here on a first successful generation of synchrotron radiation from laser accelerated relativistic electrons [4]
Fig. 1. The experimental setup of the all optical synchrotron light source consists of a laser wakefield accelerator as source of a relativistic electron beam, electron beam diagnostics, an undulator, an electron spectrometer and an optical spectrometer. All parts are aligned along a line and are located inside vacuum chambers, except the optical spectrometer. An essential feature of the experiment is that the acceleration region, the electron spectrum and the undulator spectrum were simultaneously recorded for each individual shot
Results and Prospects Ultrashort monochromatic electron pulses with energies around 60 MeV are produced by the interaction of a high intensity laser (Jena high-intensity Titanium:Sapphire laser JETI, 80 fs, 430 mJ on target) with a He gas jet. These electrons propagate through a static undulator of 2 cm period length, where they undergo oscillations perpendicular to the magnetic field and the propagation direction and therefore emit polarized radiation. The wavelength λ of the emitted light is mainly determined by the undulator period λu and the electron energy Ee = γ m0 c 2 (γ being the Lorentz factor, m0 the electron’s rest mass) and to second order by the undulator parameter K, determining the amplitude and therefore the anharmonicity of the oscillatory motion. Furthermore the energy of the emitted photons is peaked in the forward direction. For K < 1 the emitted wavelength in forward direction is approximately λ ≅ λu / 2γ 2 . For electron energies around 60 to 70 MeV the wavelength emitted by our undulator ( λu = 2 cm ) is in the visible spectral range, see Fig. 2. In the presented simple approach of laser electron acceleration, the generation of the plasma with its required complex density profile, and the acceleration of the electrons therein, have to be accomplished by the same intense laser pulse. Since several nonlinear processes are involved, this scenario is not perfectly predictable and repeatable, resulting in shot to shot variations of energy, spectral width, yield and direction. However, different plasma target designs as colliding pulses or capillary discharges have proven that more stable conditions can be achieved [5,3]. Due to the favourable 1/ γ 2 scaling of the undulator radiation, ultrashort, incoherent light pulses
814
in the UV and soft x-ray spectral range can be produced with today’s table-top high intensity lasers by the method described here. The generation of coherent undulator radiation is more challenging since it requires electron pulses shorter than the emitted wavelength. With today's laser accelerated electrons this might be possible in the infrared spectral range, using undulators with a long period. In the UV or even x-ray region one has to rely on the SASE process. An interesting aspect of a purely laser driven synchrotron source is the temporal coupling of the laser light and the undulator radiation: for each time resolved experiment at least two pulses are required, one to start the fast process to be observed and one to probe it after a well defined temporal delay. One of the two pulses typically is a femtosecond laser pulse, which can be generated by the same laser source which drives the electron acceleration and is by that perfectly synchronized to the electron pulse and the undulator radiation respectively. In conclusion, the combination of laser electron acceleration and undulator technology opens exciting novel opportunities for ultrafast spectroscopy with short wavelength brilliant light pulses.
Fig. 2. Electron spectrum (inset) and undulator radiation spectrum (dark grey line). Both spectra were recorded for the same laser shot. The electron spectrum is peaked at 64 MeV, has a width of 3.4 MeV (FWHM) and contains a charge of 10 pC. The undulator radiation is peaked at 740 nm with a bandwidth of 55 nm, containing 250,000 photons. Expected undulator radiation was simulated (light grey line) from the measured electron spectrum (inset) taking into account undulator parameters and solid angle of radiation detection. There is an excellent agreement of the spectral position and width between measured and expected undulator radiation and a reasonable agreement in yield [4] 1 2 3 4 5
H. Chapman et al., in Nature Physics, Vol. 2, 839, 2006. M. Pittman, S. Ferré, J. Rousseau, L. Notebaert, J. Chambaret, G. Chériaux, in Applied Physics B, Vol. 74, 529, 2002. W. Leemans, B. Nagler, A. Gonsalves, Cs. Toth, K. Nakamura, C. Geddes, E. Esarey, C. Schroeder, S. Hooker, in Nature Physics, Vol. 2, 696, 2006. H.-P. Schlenvoigt, K. Haupt, A. Debus, F. Budde, O. Jäckel, S. Pfotenhauer, H. Schwoerer, E. Rohwer, J. Gallacher, E. Brunetti, R. Shanks, S. Wiggins, D. Jaroszynski, in Nature Physics, Vol. 4, 130, 2008. J. Faure, C. Rechatin, A. Norlin, A. Lifschitz, Y Glinec, V. Malka, in Nature, Vol. 444, 737, 2006.
815
Compression of an Ultraviolet Pulse by Molecular Phase Modulation and Self-Phase Modulation Yuichiro Kida1, Shin-ichi Zaitsu1, and Totaro Imasaka1,2 1
Department of Applied Chemistry, Graduate School of Engineering, Kyushu University, 744, Motooka, Nishi-ku, Fukuoka 819-0395, Japan E-mail: [email protected] 2 Division of Translational Research, Center for Future Chemistry, Kyushu University, 744, Motooka, Nishi-ku, Fukuoka 819-0395, Japan E-mail: [email protected] Abstract. A compression scheme for an ultraviolet pulse to sub-15 fs is reported. Frequency modulation of an ultraviolet pulse by molecular rotations and by self-phase modulation results in a compressed pulse with small intensities of sub-pulses.
Introduction An ultrashort ultraviolet (UV) pulse is a useful mean for the investigation of the ultrafast phenomena of an organic compound. For generation of the pulse, techniques based on self-phase modulation (SPM), difference-frequency mixing (DFM), noncollinear parametric amplification (NOPA), and Molecular Phase Modulation (MPM) have been reported. In the technique based on MPM, the frequency of a UV pulse (probe pulse) is modulated by a coherent molecular motion induced by an intense pulse (pump pulse). In this case, the energy of the modulated pulse is determined by the energy of the original probe pulse [1]. A high-energy ultrashort UV pulse is, therefore, generated by use of a high energy probe pulse whose width is long enough to prevent the ionization of the molecules which leads to undesired phenomena such as deterioration of beam profile. The use of the long probe pulse, however, leads to generation of sub-pulses in the temporal profile [2]. To reduce the intensities of the sub-pulses, SPM induced by the probe pulse is considered here. When the energy of the probe pulse is high enough for inducing SPM, the frequency of the probe pulse is simultaneously modulated by a coherent molecular motion and by SPM. The spectrum of the modulated probe pulse has a broad width arising from the molecular phase modulation, and also has a dense structure arising from SPM.
Experimental Methods A near-infrared pulse provided from a Ti:sapphire chirped pulse amplifier (784 nm, 1.2 mJ, 110 fs) was passed through a LiB3O5 crystal to generate a UV pulse (392 nm) that was used as a probe pulse. The near-infrared pulse (pump pulse) and the probe pulse emitted from the crystal were separated from each other with a dichloic mirror for controlling the time delay of the probe pulse with respect to the pump pulse. The two pulses were then recombined and were focused into a Raman cell filled with pressurized hydrogen gas (10 atm). The probe beam transmitted from the Raman cell was collimated with a concave mirror before compensation for the GDD and TOD of the probe pulse by use of a prism compressor and a grating compressor. The compressed probe pulse in the compressors was, then, propagated to a self-diffraction
816
(SD) autocorrelator. The time delay of the probe pulse was adjusted to 0.9 ps, and the energies of the input pump and probe pulses were 260 μJ and 80 μJ, respectively.
1.0 (a) 0.8 0.6 0.4 0.2 0.0 360 380
Intensity (a. u.)
Intensity (a. u.)
Results and Discussion
400
420
Wavelength (nm)
440
1.0 (b) 0.8 0.6 0.4 0.2 0.0 360 380
400
420
440
Wavelength (nm)
Fig. 1. (a) The spectra of the probe pulse emitted from the Raman cell (solid line). The spectrum of the probe pulse measured in the case that only the probe pulse was focused into the cell (broken line). (b) The spectrum of the probe pulse measured in front of the autocorrelator (solid line). A typical spectrum measured under the situation that SPM is not induced (Ref. 2) is shown in the figure with broken line.
The spectrum of the probe pulse emitted from the Raman cell is shown in Fig. 1 (a). The frequency of the probe pulse was modulated by the coherent rotation of orthohydrogen and by SPM. The spectral width (FWHM) was broader than that expanded only by SPM [broken line in Fig. 1 (a)], and the spectral structure was denser than that modulated without SPM [broken line in Fig. 1 (b)] [2]. After the dispersion compensation for the probe pulse, the relative intensity of each spectral component was changed to that shown with solid line in Fig. 1 (b). This was due to the fact that the output energy from the compressors was different in each spectral component, and was also due to the fact that the frequency modulation was induced efficiently in the center part of the beam. The center part was extracted using an aperture before the measurements of the spectrum [Fig. 1 (b)] and the autocorrelation trace described below. The spectrum shown in Fig. 1 (b) is broad enough for generation of a 10-fs UV pulse as indicated by the inverse Fourier-transform [IFT, Fig. 2 (a)] of the spectrum [solid line in Fig. 1 (b)]. The waveform of the IFT contains no sub-pulses, while the measured SD-autocorrelation trace shown in Fig. 2 (b) consists of a main peak structure and small sub-pulses in the vicinity of the structure. In principle, the spectral phase of a probe pulse modulated by the phase modulations shows a complicated structure and it can not be compensated only by GDD and TOD compensations. The residual phase distortion leads to the generation of the sub-pulses even when perfect compensation for GDD and TOD are demonstrated. Though complete removal of the sub-pulses from the temporal profile was not demonstrated, the intensities of the sub-pulses were lower than those in the case where SPM was not induced [2]. Hence the generation of SPM is useful for suppressing the generation of the sub-pulses in the temporal profile. The FWHM of the main peak in the autocorrelation trace was 17 fs, corresponding to the FWHM of the SDautocorrelation trace of a 14-fs Gaussian pulse [solid line in Fig. 2 (b)]. The
817
10 fs
0 Time (fs)
50
100
SD Intensity (a. u.)
Intensity (a. u.)
1.0 (a) 0.8 0.6 0.4 0.2 0.0 -100 -50
1.0 (b 0.8 0.6 0.4 0.2 0.0 -100 -50
14 fs
0
50
100
Delay (fs)
Fig. 2. (a) Inverse Fourier transform of the spectrum (Fig. 1 (b)). (b) The measured (solid circles) and calculated (solid line) SD-autocorrelation traces.
good correspondence between the main structure of the measured trace and the trace of the Gaussian pulse suggests that the FWHM of the main structure of the compressed probe pulse was ca. 14 fs. The width of 14 fs is slightly longer than that of the width of the IFT (10 fs), which would be due to imperfect compensation for the TOD. Though a frequency-resolved SD-autocorrelation trace (spectrogram) was measured to estimate the amount of the dispersion, it could not be easily estimated since the spectrogram is not appreciably distorted by a small amount of TOD [3]. The width of 14 fs is, however, much shorter than the width of the original probe pulse (ca. 100 fs). Furthermore, the energy of the compressed pulse was not so small (4.5 μJ) despite the much energy loss in the two compressors.
Conclusions As reported here, the spectrum of a 100-fs UV pulse is modulated by the coherent rotation of hydrogen molecules and by SPM for generation of a 10-fs UV pulse. The resultant spectral width is wide enough for generation of a 10-fs pulse, and the pulse width of the compressed pulse is sub-15 fs which would be determined mainly by the precision in the dispersion compensation. The temporal profile of the compressed pulse has the sub-pulses whose intensities are small compared to those in the case where SPM is not induced. Since the GDD and TOD are compensated with a prism and grating compressors, this approach can be applied to compression of a deep-UV (DUV) pulse by only replacing the probe pulse with a DUV pulse. Acknowledgements. This research was supported by Research Fellowships of the Japan Society for the Promotion of Science (JSPS) for Young Scientists, Grants-inAid for Scientific Research, and a Grant-in-Aid for the Global COE program, “Science for Future Molecular Systems”, from the Ministry of Education, Culture, Science, Sports and Technology of Japan. 1 2 3
818
N. Zhavoronkov and G. Korn, Phys. Rev. Lett. 88, 203901, 2002. Y. Kida, T. Nagahara, S. Zaitsu, M. Matsuse, and T. Imasaka, Opt. Exp. 14, 3038, 2006. K. W. DeLong, R. Trebino, and D. J. Kane, J. Opt. Soc. Am. B 11, 1595, 1994.
Temporal Optimization of Ultrabroadband Optical Parametric Chirped Pulse Amplification Jeffrey Moses1, Cristian Manzoni2, Shu-Wei Huang1, Giulio Cerullo2, and Franz X. Kärtner1 1
Department of Electrical Engineering and Computer Science and Research Laboratory of Electronics, Massachusetts Institute of Technology, Cambridge, MA 02139, USA E-mail: [email protected] 2 ULTRAS-INFM-CNR Dipartimento di Fisica, Politecnico, Piazza L. da Vinci 32, 20133 Milano, Italy E-mail: [email protected] Abstract. Critical optimization considerations are presented for ultrabroadband, high-power optical parametric chirped-pulse amplifiers, where simultaneous suppression of superfluorescence and maximization of both conversion efficiency and bandwidth is required. Numerical simulations verify theory.
Today’s demands on light sources for high-intensity ultrafast optics research are stringent: peak power must be maximized by scaling both to high energy and nearsingle-cycle duration, signal to noise contrast must be high, and often pulses at nontraditional wavelengths must be generated. These requirements have led to the rapid development of ultrabroadband optical parametric chirped pulse amplification (OPCPA) pumped by powerful picosecond pulses, in which gain bandwidth is stretched to near-octave breadths by group-velocity matching between signal and idler. In recent years, several problems in the construction of these amplifiers have become relevant. The coupling of temporal gain narrowing and spectral narrowing results in a trade-off between conversion efficiency and bandwidth. Additionally, the amplifier seed energy is often low while total gain is high, resulting in high levels of parametric superfluorescence and poor signal to noise ratio [1, 2]. While the effect of temporal gain narrowing on ultrabroadband OPCPA has been investigated [3], a study of simultaneous optimization of conversion efficiency, signal bandwidth and signal-to-superfluorescence ratio has not yet been presented, and several details of the temporal optimization problem have been neglected. In this paper we study the simultaneous optimization of conversion efficiency, signal bandwidth and noise suppression in ultrabroadband OPCPA as a function of the ratio of pump and seed pulse durations, their relative energy and the total parametric gain set by the pump intensity. The optimal pump-seed pulse duration ratios for maximization of the efficiency-bandwidth product and signal-to-noise ratio are found to depend on the total gain. The basic compromise between conversion efficiency and amplified signal bandwidth in OPCPA is conceptually well understood: for a chirped seed pulse, frequency is mapped to time, and thus when conversion efficiency is maximized by stretching the seed pulse to cover the full temporal gain profile of the amplifier, the short- and long-wavelength wings of the seed pulse experience significantly lower gain than the peak. Hence, temporal gain narrowing results in spectral narrowing. The difference between chirped pulse and non-chirped pulse parametric amplification is depicted in Fig. 1 for a Gaussian pump pulse and a peak gain of 100. If pump beam depletion can be neglected, parametric gain is G = 1 + (Γ2/g2)sinh2(gL), where g
819
= [Γ 2 − (∆k/2)2]1/2, ∆k is the wavevector mismatch, and Γ2 is the nonlinear drive, proportional to the pump intensity Ip. Given a pump profile Ip(t), with perfect phase matching and at high gain we may find the temporal region of overlap (|t| < tg) of the seed with the pump pulse where gain is at least e-a times the gain at t = 0 by setting
(1) where a = 1. Assuming the pump has a Gaussian temporal profile, we can find tg by setting Ip(tg)/Ip(0) = exp(-tg2/τ02). We find (2)
Fig. 1. Schematics of high-conversion-efficiency (a) optical parametric amplification, (b) OPCPA in which amplified signal bandwidth is maximized.
The case for an unchirped seed is depicted in Fig. 1(a). In order to maximize conversion efficiency and suppress superfluorescence noise, the temporal extent of the seed pulse should exactly fill the region |t| < tg [shaded region in Fig. 1(a)]. If the pulse is too short, a portion of the available pump energy remains unconverted to signal after amplification and regions where there is significant gain remain unseeded, resulting in noise amplification without quenching of the gain by the signal. If the pulse is too long, additional gain is needed to saturate the amplifier. An important feature of Eq. 2 is the dependence of the gain temporal width on the peak gain. As a consequence, the optimal seed duration varies from stage to stage in a multiple-stage amplifier. Note that in the pump-depletion regime the region of significant gain becomes slightly wider, since the gain at the temporal and spectral wings increases with respect to the peak gain. In the case of a chirped seed pulse [Fig. 1(b)], the spectral and temporal gain profiles simultaneously affect the width of the significant gain region. In order to employ the full gain bandwidth 2δωb of the amplifier [defined as G(ω0 ± δωb) = e-1G(ω0)] the gain temporal width is significantly narrower than in the unchirped seed case: at a certain temporal coordinate, both the drop in pump intensity and the instantaneous wavevector mismatch combine to reduce the gain by e-1. An appropriate definition of the new gain temporal width, |t| < t’g, and the amplification bandwidth, 2δωb’, is where each effect reduces the gain by e-1/2, or when a = 1/2 in Eq. 2. In this case the gain bandwidth is reduced by 30% in a regular OPCPA and by only 16% in the case of broadband group-velocity-matched amplification.
820
When suppression of superfluorescence is necessary, however, the optimization problem is more complex: the seed is temporally chirped but the incoherent noise is not. In Fig. 1, the signal and noise fields have different gain widths, with ratio ~0.7:1. As a result, the average noise gain across the pump pulse is higher than the average signal gain. Moreover, if the seed is stretched such that it fills the region |t| < t′g [to maximize amplifier bandwidth, as in Fig. 1(b)], the low seed energy in the region t′g < |t| < tg, will allow significant conversion of pump to noise. Thus, we conclude that the seed should be chirped to at least cover the region |t| < tg (Eq. 2, a = 1), resulting in some reduction in bandwidth but maximizing conversion efficiency and preventing strong degradation of signal to noise ratio. 3.0
4
Ep/Es = 10
seed duration
Total Efficiency
0.5 0.4 0.3 0.2 0.1 0.0
(a)
1.045 ps 2.0
2.5
3.0
3.5
4.0
4.5
5.0
5.5 2
Pump Peak Intensity (GW/cm )
6.0
6
10 4 10 2 10
25 20
2.0
15
7 ps 8.5 ps
12 ps 1.5
10
10 5
6
1.0 0
(b)
2.5
2
4
6
8
10
12
Noise-to-Signal (%)
Gain Bandwidth product (THz)
12.588 ps
0.6
14
Seed duration (ps)
Fig. 2. (a) Simulation results for conversion efficiency (signal + idler) for several values of seed pulse duration and G = 104. The squares denote where maximum pump depletion occurs. (b) Efficiency-bandwidth product (at maximum possible conversion efficiency) versus seed pulse duration for G = 102, 104, and 106. Triangles denote noise-to- signal ratio for G = 106.
To verify predictions of the analytical model, we numerically simulated an OPCPA by solving the nonlinear coupled equations for the case of collinear interaction and plane-waves. We considered a 3-mm thick PPLN crystal, pumped by 9-ps Gaussian pulses at 1.047 µm for broadband degenerate amplification at 2 µm. The superGaussian seed had a bandwidth corresponding to 12-fs transform limit, larger than the acceptance of the crystal. By varying both the pump intensity and the seed second order chirp, we explored the parameter space as shown in Fig. 2(a) for G = 104. Maximum conversion efficiency (squares) increases strongly with increasing seed duration until the significant gain region of the pump is filled. At this point the growth in conversion efficiency relaxes, while at the same time bandwidth begins to strongly decrease. Using these datapoints from Fig. 2(a), the efficiency-bandwidth product [panel (b), filled squares] is maximized when the seed duration is close to the gain region width predicted by Eq. 2. The same numerical analysis was conducted for the cases of a pump to signal energy ratio of 102 (circles) and 106 (open squares): a comparison between the three amplification regimes shows that the optimal seed to pump duration ratio decreases with increasing pump-to-signal energy ratio, as predicted by Eq. 2. In other words, the optimal seed duration decreases for increasing gains. Finally, a minimum in noise to signal ratio is found close to the maximum gain-bandwidth product (Fig. 2(a), triangles [G = 106]). 1 2 3
F. Tavella, A. Marcinkevicius, and F. Krausz, New J. of Phys. 8, 219, 2006. T. Fuji, N. Ishii, C. Y. Teisset, X. Gu, T. Metzger, A. Baltuška, N. Forget, D. Kaplan, A. Galvanauskas, and F. Krausz, Opt. Lett. 31, 1103, 2006. S. Witte, R. T. Zinkstok, W. Hogervorst, K. S. E. Eikema, Appl. Phys. B 87, 677, 2007.
821
Third Harmonic X-waves Generation by Filamentation of Infrared Femtosecond Laser Pulses in Air Han Xu1, Hui Xiong1, See Leang Chin2, Ya Cheng1 and Zhizhan Xu1 1
State Key Laboratory of High Field Laser Physics, Shanghai Institute of Optics and Fine Mechanics, Chinese Academy of Sciences P.O. Box 800-211, Shanghai 201800, China 2 Centre d’Optique, Photonique et Laser (COPL) and Département de physique, de génie physique et d’optique, Université Laval, Québec, Québec G1K 7P4, Canada E-mail: [email protected], [email protected] Abstract. We report the first measurement of the hyperbolic featured angularly resolved spectra of the X-shaped third harmonic generated in infrared femtosecond pulses pumped filament in air. We show that at low pump intensity, phase matching between the fundamental and third harmonic waves dominates the nonlinear optical effect and induces a ring structure of the third harmonic beam; whereas at high pump intensity, the dispersion properties of air begins to affect the angular spectrum, leading to the formation of nonlinear X-wave at third harmonic.
Introduction In propagation of intense laser pulse in Kerr nonlinear medium, stable optical filament could be generated, and efficient third harmonic (TH) generation [1-5] could arise due to the very high intensity achieved in filamentation. Due to the potential application in laser frequency conversion to shorter wavelength, TH generation through optical filaments in air has been intensively studied. Spatiotemporal coupling during nonlinear propagation would naturally result in complex spatiotemporal structures in both the fundamental and TH pulses. The traditional spectrum detection technique, however, is not sufficient for the diagnosis of the complex wavepacket, while a simple but powerful diagnostic method, namely, the measurement of angularly resolved spectra, has been developed [6] for this purpose. Angularly resolved spectra could provide surprisingly detailed portrait of complex wavepacket. In this Letter, we report on the experimental measurement of angularly resolved spectrum of third harmonic generated by intense infrared pulse after its filamentation in air. The X featured angularly resolved spectrum of TH indicates that the TH wavepacket is transformed into nonlinear X wave after filamentation [6], and the intensity-dependent spectrum gives a clear evidence of the strong nonlinear phase locking between the FW and the TH during their co-propagation [1-3].
Experimental Methods The carrier wavelength from our OPA is tuned to 1270nm with single pulse energy of up to 487nm and pulse duration of about 20fs-25fs. After being separated from the idler using a broadband reflection mirror (M1, high reflection at 1100nm-1300nm), the infrared pump pulse is then tightly focused by a gold coated concave mirror with
822
a focal length of 250mm to generate a single light filament in air near the geometrical focal point. After filamentation, an angularly resolved spectrometer was employed to portrait the angular resolved spectrum of the output TH pulse; the far-field spatial pattern as well as the filament is thus captured on a digital camera (Coolpix995, Nikon, Japan). The angular-resolved (ky-ω) spectrum is recorded by angularly resolved spectrometer consists of a positive lens (L, focal length: 300mm) and a grating spectrometer (SpectraPro 300i, Acton)
Fig.1. (a) The experimental setup, and the digial camera captured images of (b) the side view of the filament in air, and (c) the far-field spatial pattern of the third harmonic.
Results and Discussion Figure 2 shows the evolution of the angular-resolved spectrum (θ-λ) of TH with the increase of the pumping pulse power from 30 μJ to 487 μJ. We note that there are mainly two mechanisms contributing to the process of third-order harmonic generation. When the IR pulse energy is between ~200 μ J and ~400 μ J, the mechanism of longitude phase matching between the generated TH and the fundamental wave dominates the TH generation process, which requires: kz(3ω)=3k(ω).
(1)
The conversion efficiencies of both the ring and axial components increase with the increasing pump energy, and the ring TH is much stronger due to the longitude phase matching condition could be satisfied on the ring. When the IR pulse energy is further increased to > 400 μJ, two hyperbolic structured tails (tail2, solid curve and tail3, dash-dot curve) emerge and grow rapidly. TH frequency components growing along these tails obey the mechanism of group velocity matching between the generated TH and the pump wave, which requires [7]: kz(ω)= k(ωs)+(ω-ωs)/Vg
(2)
,where ωs is the frequency of the scattered TH wave, and Vg is group velocity of the pump wave. By fitting the measured angularly resolved spectra with equation 2, the group velocity of tail 2 and tail 3 could be obtained, which reveals that both the tails
823
propagating at the same group velocity and locked to the same peak of the pump pulse.
Fig. 2. Normalized angular resolved spectrum of TH in air with FW pulse energy at (a) 135μJ , (b) 240μJ , (c) 410μJ , and (d) 481μJ (in logarithmic scale and normalized by maximum intensity of TH).
Conclusions We systematically investigate the evolution of angular spectrum of TH wave generated by IR ultrashort pulse filamentation in air. We speculate that there exist two physical mechanisms governing the evolution of angular spectrum of TH wave, implying different origins of conical ring and X-wave. Acknowledgements. This work was supported by the National Basic Research Program of China (Grant No. 2006CB806000), Shanghai Commission of Science and Technology (Grant No. 07JC14055), and National Natural Science Foundation of China (Grant No. 10523003). S. L. C acknowledges the support of the Canada Research Chairs program.. 1 2 3 4 5 6 7
824
N. Aköbek, A. Iwasaki, A. Becker, M. Scalora, S.L. Chin, and C.M. Bowden, Phys. Rev. Lett. Vol. 89, 143901, 2002 F. Théberge, N. Aközbek, W. Liu, J.–F. Gravel, and S. L. Chin, Optics Communications. Vol. 245, 399, 2005 F. Théberge, N. Aközbek, W. Liu, J. Filion, and S. L. Chin, Optics Communications. Vol. 276, 298, 2007. H. Xu, H. Xiong, R. Li, Y. Cheng, Z. Z. Xu, and S. L. Chin, Appl. Phys. Lett., Vol. 92, 011111, 2008 H. Xiong, H. Xu, Y. Fu, Y. Cheng, Z. Z. Xu, and S. L. Chin, Phys. Rev. A, Vol. 77, 043802, 2008 D. Faccio, P. Di Trapani, S. Minardi, A. Bramati, F. Bragheri, C. Liberale, V. Degiorgio, A. Dubietis, A. Matijosius, J. Opt. Soc. Am. B, Vol. 22, 862, 2005 D. Faccio, et al., Opt. Express, Vol. 15, 13077, 2007
Generation and control of coherent conical pulses in seeded optical parametric amplification Ottavia Jedrkiewicz1, Matteo Clerici1, Daniele Faccio1 and Paolo Di Trapani1,2 1
Cnism and Dipartimento di Fisica e Matematica, Università dell’Insubria, Via Valleggio 11, 22100 Como (Italy) E-mail: [email protected] 2 Department of Quantum Electronics, Vilnius University, Sauletekio 9, LT-10222 Vilnius (Lithuania)
Abstract. We propose a new technique for high-energy conical pulse generation based on continuum seeded parametric amplification process in quadratic nonlinear media. We show that by using an appropriate broadband week seed we are able to lock the many spatiotemporal modes of the parametric radiation to obtain a single, coherent conical pulse.
Introduction Conical waves are peculiar wave-packets, in which the energy flow is not directed along the beam axis, as in conventional waves. In contrast here, the energy arrives laterally, i.e. from a cone-shaped surface, leading to the appearance of a very intense and localized interference peak at the cone vertex, as in Bessel (or Durnin) beams, in the continuous wave limit, and in the so-called X-waves, in the ultra-short pulse regime. The conical nature of these waves allow them to carry angular dispersion, i.e. a controlled dependence of temporal frequencies on angles. As a consequence, they can propagate stationary in linear as well as in nonlinear media, in spite of diffraction and material dispersion, and exhibit tunable “effective” phase and group velocities, which result in unique dispersion-management features. Recent experiments in phasemismatched second harmonic generation [1] have highlighted the existence of interlinked spatial and temporal processes, which cannot be considered separately within the nonlinear dynamics, as consequence of the presence of angular dispersion. In quadratic processes, this space and time coupling has also been observed in parametric down-conversion where the generated superfluorescence could be interpreted as a stochastic “gas” of quasi-stationary X-type modes characterized in the spatiotemporal domain by a skewed correlation surrounding a very sharp peak [2,3]. The work presented here extends that previous study with the goal of generating and controlling single coherent conical pulses in seeded Optical Parametric Amplification (OPA). Here we show that by using an appropriate broadband week seed in OPA we are able to lock the many incoherent spatiotemporal modes of the parametric downconverted radiation to obtain a single, coherent, conical pulse.
T = 105° 825
a
T = 105°
c
T = 130°
b
d
T = 130°
Fig. 1. Spatiotemporal spectrum of the parametric radiation generated in OPA recorded for two different crystal temperatures (two different phase-matching configurations) in non seeded (a and c) and continuum seeded (b and d) configurations respectively.
Experimental Methods The OPA was seeded by a broadband coherent week radiation, which was obtained by spectral windowing a filament induced by focusing in a 5cm BK7 bulk sample a 1ps laser pulse at 1055nm. The continuum seed generated over a smooth bandwidth in the range of 600-800nm (with energy in the µJ range) was suitably focused and injected collinearly with the pump inside the LBO crystal. The temporal overlap of the pump pulse and seed pulse inside the crystal was controlled by means of a delay line. The statistical and coherence properties of the radiation generated in the OPA process can be studied by recording the far-field spatiotemporal (θ,λ) spectrum, in analogy to [2,3]. To this end the radiation at the output of the LBO crystal was collected by means of a telescope and the far-field diagnostic was based on an imaging spectrometer coupled with a 16 bit high efficiency CCD camera (Andor) for the visible, and a InGAs CCD (Xenics) for the infrared region.
Results and Discussion The single shot spatiotemporal spectra of the signal recorded in non seeded and seeded configuration respectively are presented in Fig.1 for two different crystal temperatures, both corresponding to emission out of degeneracy (for best superposition range with spectral characteristics of continuum).
826
a
b
Fig. 2. Portion of single shot spatiotemporal spectrum of the generated idler field recorded by an InGAs CCD.
The parametric down converted radiation is generated along the so-called phasematching curve of the quadratic nonlinear process, satifying the energy and momentum conservation laws. Thus when the broadband coherent continuum seed is injected inside the crystal and the temporal superposition is guaranteed, the amplification process occurs over the same broadband region dictated by the phasematching curves. In addition we can clearly observe the spatiotemporal mode locking process, which leads to smoothed spatiotemporal spectral curves (in single shot) of the amplified signal. Note that because of the phase-conjugation property of amplified signal/idler radiation in OPA, the single shot idler spectrum in infrared region turns out to be similarly smoothed. Experimentally, because of the huge spatial and spectral bandwidth of the idler radiation, only a portion of the expected ring type (similarly to the signal) spectrum of the idler could be recorded, as shown in Fig.2. The spectral distribution of the obtained OPA radiation is characteristics of the socalled O-waves, which are stationary modes in materials with anomalous dispersion (λ>1µm) [4]. Moreover in analogy to what done in [2,3], the evaluation of the spacetime coherence function of the generated radiation leads to an “onion” type structure in the (r,t) space, characterized by a central core of about 10µm and 10fs respectively.
Conclusions The results presented here permit to evidence the possibility of generating in seeded OPA single ultrashort localized conical wave modes in space and time, with peak dimensions of the order of few µm and fs respectively, in analogy with previous work [2,3]. Further work on pulse reshaping and control in OPA, in particular around degeneracy, with the aim of generating X-type or Bessel like pulses is in progress, together with the study of stationarity of the single wave modes. 1 2 3 4
P. Di Trapani, G. Valiulis, A. Piskarskas, O. Jedrkiewicz, J. Trull, C. Conti, S. Trillo, Phys. Rev. Lett. 91, 093904 (2003). O. Jedrkiewicz, A. Picozzi, M. Clerici, D. Faccio, P. Di Trapani, Phys. Rev. Lett. 97, 243903 (2006). O. Jedrkiewicz, M. Clerici, A. Picozzi, D. Faccio, and P. Di Trapani, Phys. Rev. A 76, 033823 (2007). M. A. Porras and P. Di Trapani, Phys. Rev. E 69, 066606 (2004).
827
Generation of Ultrashort Optical Pulses Using Multiple Coherent Anti-Stokes Raman Scattering Signals in a Crystal and Observation of the Raman Phase Eiichi Matsubara, Taro Sekikawa, and Mikio Yamashita Department of Applied Physics, Hokkaido University, and Core Research for Evolutional Science and Technology (CREST), Japan Science and Technology Agency, Kita-13, Nishi-8, Kita-ku, Sapporo, 060-8628, Japan E-mail: [email protected] Abstract. We demonstrate Fourier synthesis of the multiple coherent anti-Stokes Ramanscattering signals in LiNbO3. Both angle and temporal dispersions of the signals are compensated for by a modified 4f configuration. As a result, isolated pulses with 25-fs duration (640−780 nm) are generated and discrete phase shifts due to Raman coherence are observed.
Introduction Generation of one-optical-cycle pulses is attractive for many applications. One of the promising methods to achieve it is the Fourier synthesis of multiple Raman sidebands. So far, by using a cooled D2 gas as a Raman medium, a train of 1.6-fs pulses have already been generated [1]. However, its repetition period is too short (11 fs) because spectra of the adjacent Raman sidebands do not overlap at all. Recently, it has been found that multiple coherent anti-Stokes Raman scattering (CARS) signals with broad bandwidths can be generated even in crystals [2] by the excitation with two-color femtosecond laser pulses at room temperature, in which all the spectra of adjacent CARS peaks overlap so that the entire spectrum is continuous. This property gives two advantages over the cases of a gas. One is the potential of the generation of “isolated” pulses. The other is that the spectral phase can be directly measured by spectral phase interferometry for direct electric field reconstruction (SPIDER). This is important not only from the viewpoint of the ultrafast optical pulse technology but also from the viewpoint of physics in quantum optics because whether or not discrete phase shifts of multiple CARS signals due to Raman coherence formation is an interesting issue.
Experimental Setup and Results The experimental set up is shown in Fig. 1. The YZ-surface of a LiNbO3 crystal with a 0.5-mm thickness (LN) was simultaneously irradiated by two slightly chirped (+100 to 200 fs2) fundamental laser pulses from a multi-pass Ti:sapphire laser amplification system (center wavelength: 810 nm, duration: 30 fs, repetition rate: 1 kHz). The input pulse energies were 6 and 11 µJ. The relative angle between the two beams was 1.9° in the air, because in the case of crystals refractive index dispersions are large so that two collinear input beams cannot generate CARS signals efficiently. The angle dispersion of the signals was compensated into one white-continuum beam by a modified 4f configuration which consists of a spherical mirror with a 100-mm focal length (SM), a rectangular mirror (RM), a grating with a groove density of 1200
828
lines/mm and a blaze wavelength at 500 nm (G). To tune the magnification of f1/f2 arbitrarily, only one spherical mirror was used. The collimated beam was picked up by a small square mirror (SQM), and collimated both vertically and horizontally by two cylindrical mirrors (CM1 and CM2: focal lengths were 100 and 250 mm, respectively). Figure 2 shows the intensity spectrum and the spectral phase profiles of the collimated CARS beam measured by the modified SPIDER. It was found that the sign of the dispersion could be both positive (upper dotted curve) and negative (lower dotted curve). By appropriately adjusting the rotation angle and the position of the grating, the spectral phase became almost flat, within a 10-rad variation in the frequency range from 13000 to 15600 cm−1 (middle dotted curve). Here, we can see many step-like changes in the spectral phase. They originate from the Raman phase which is expected to slip discretely with increasing order of CARS signals. The temporal intensity profile of the generated pulse (black solid curve) and that of the Fourier-transform-limited (TL) one (black dotted curve) are shown in Fig. 3. The gray solid curve shows the temporal phase profile of the generated pulse, which is almost flat during the pulse duration. An isolated pulse with a duration of 25 fs is observed, while that of the TL one is 11 fs. This difference seems to come from the incompletely flattened spectral phase.
Discussion Now we discuss the present results. First, we explain how the compensations of both the angle and the temporal dispersion of the multiple CARS signals are achieved in our experimental setup. The position and the angle of the grating (G) play the most important role. It is known that angle dispersive elements, such as prisms and gratings in a 4f configuration, introduce group delay dispersion (GDD: d2Ψ/dω2) in such a manner as [3], 2 dθ 1 d 2Ψ d 2θ dθ sin θ + ω cosθ = − ( z ′M 2 + z) 2 +ω 2 2 (1) dω c dω dω dω
Here, θ is the diffracted angle of a ray with an angular frequency ω, and is measured from the direction of the center-frequency component. z is the distance between the crystal and the first focal point, and z’ is the distance between the second focal point and the grating. Because dθ/dω is a function of the angle of the grating, the GDD depends on both the angle and the position of the grating. Using parameters of z=0, z’= ± 300 µm, and M=0.74, the applied GDD is m 200 fs2 at 700 nm. If we rotate the grating by 1°, the GDD changes by 5 fs2. Thus we can understand that the temporal dispersion of the collimated beam is adjustable by slightly moving the grating. Next we discuss why the spectral phase was not flattened perfectly. Although it is almost flat within 1-rad fluctuation in the frequency range from 13200 to 14500 cm−1, it increases a little steeply to 11 rad in the range from 14600 to 15600 cm−1. We think this is due to the nonnegligible higher-order dispersion applied by the modified 4f configuration. It will be certain that, by using a programmable multi-channel spatial light modulator [4] or well-designed chirped mirrors, we can generate TL pulses with this scheme. Finally we discuss the discrete phase shift due to Raman phase. Figure 4 shows the expanded spectral phase profile in Fig. 2 (the middle one). As eye-guided by dotted and dashed vertical lines, ripples with spacings of 155 and 369 cm−1 are seen, which correspond to the frequencies of TO (E) phonons. On the other hand, the
829
spectral resolution of SPIDER determined by the spectral shear (170 cm−1) is 340 cm−1 according to the Nyquist limit, so that one might think that such phonon frequencies are not observable. However, a simple simulation shows that some information on periodicity can still be retrieved even if the sampling step is nearly equal to the frequency of interest.
Fig. 1. Experimental setup for the generation of ultrashort optical pulses using multiple CARS signals in LiNbO3.
Fig. 2. Intensity spectrum (black solid curve) and spectral phase profiles (black dotted curves) of the generated pulse.
Fig. 3. Temporal-intensity profiles of the generated pulse (solid curve) and Fouriertransform limited one (dotted curve). Gray curve shows the temporal phase profile of the generated pulse.
Fig. 4. Spectral-phase profile shown in Fig. 2. As eye-guided by dotted and dashed lines, ripples with spacings of 155 and 369 cm−1 are seen.
Conclusions We demonstrated Fourier synthesis of the multiple CARS signals from a LiNbO3 single crystal. Isolated pulses with 25-fs duration were generated. In addition, discrete phase shifts of multiple CARS signals due to Raman coherence were observed.
References 1 2 3 4
M. Y. Shverdin, D. R. Walker, D. D. Yavuz, G. Y. Yin, and S. E. Harris: Phys. Rev. Lett. 94, 033904 (2005). M. Zhi and A. V. Sokolov, New J. Phys. 10, 025032 (2008), and references therein. J-C. Diels and W. Rudolph: “Ultrashort Laser Pulse Phenomena", pp75-78, Academic press, San Diego (1996). M. Yamashita, K. Yamane, and R. Morita: IEEE J. Sel. Top. Quantum Electron. 12, 213 (2006).
830
Generation of High-power Visible and UV/VUV Supercontinua and Self-compressed Single-cycle Pulses in Metal-dielectric Hollow Waveguides J. Herrmann, A. Husakou Max Born Institute for Nonlinear Optics and Short Pulse Spectroscopy, Max-Born-Str. 2a, D-12489 Berlin, Germany email: [email protected] Abstract. We investigate high-power soliton-induced supercontinuum generation in visible and UV/VUV based on argon-filled metal-dielectric hollow waveguides. We predict the generation of MW/nm spectral power densities with ∼0.1 mJ energy and self-compressed isolated 1.7-fs pulses.
The discovery of supercontinua (SC) in microstructure fibers [1] using fs pulses with nJ energy has attracted widespread interest, it has been studied extensively in recent years and found several fascinating applications. The dramatic spectral broadening for low pulse energies is related to the crucial role of soliton dynamics and soliton emission [2], requiring anomalous dispersion at the input wavelength and thus small fiber radii. However, simultaneously small radii severely restrict the possible total pulse energies to few nJ and spectral power densities to tens of W/nm. In this contribution, we predict that using specifically designed metal-dielectric hollow waveguides one can generate two-octave-broad SC with five order of magnitude higher power densities in the range of MW/nm and 0.1 mJ total energy. Besides, for optimized waveguide parameters self-compressed isolated sub-cycle pulses with a duration of 1.7 fs are predicted. In addition, we show that by pumping with the third harmonic of Ti:sapphire laser VUV supercontinua in the wavelength range of 150-500 nm can be generated. Hollow dielectric waveguides are a key element in modern ultrafast nonlinear optics, e. g. for pulse compression and attosecond pulse generation. Unfortunately, these waveguides provide tolerable loss only for relatively large radii in the range of 50 µm, thus inhibiting the control over group-velocity dispersion (GVD). In this contribution we predict that metallic hollow waveguides coated with a nm-scale dielectric layer have moderate loss even for small radii in the range of 10 to 25 µm, with a broad range of anomalous dispersion even for high gas pressures. We consider cylindrical straight hollow waveguides with metallic walls coated on the inner surface by a layer of dielectric to improve the guiding properties. The calculation of the waveguide loss and group velocity dispersion is made in the formalism of transfer matrices[3], including roughness loss modelled by pointlike scatterers. The simulation of the nonlinear pulse propagation is performed using the Forward Maxwell equations[2] including dispersion to all orders, Kerr nonlinearity and higher-order nonlinear effects, energy transfer to higher-order modes, as well as photoionization and plasma effects. In Fig. 1(a) the loss (red curve) and the group velocity dispersion (green curve) are presented for a 40-µm-radius silver waveguide coated with 45 nm of SiO2 . One can see that the waveguide loss remains in the range of 0.01-0.1 dB/m over the broad spectral range including the whole visible range, in contrast to dielectric hollow
831
Fig. 1. Characteristics of the waveguide (a) and peak spectral power density (b) for a silver-wall waveguide with a radius of 40 µm coated by a 45-nm layer of SiO2 with a average scatterer diameter of 100 nm and argon gas filling at 1 atm. In (b), we assume a 100-fs, 100 TW/cm2 input pulse at 800 nm; the propagation length is 50 cm. The output pulse duration was used for the calculation of the peak spectral power density.
waveguides which have much higher loss for this radius. The group velocity dispersion is anomalous for all wavelengths above 600 nm for argon filling at 1 atm. The output peak spectral power density for a 100-fs, 100-TW/cm2 input pulse is presented in Fig. 1(b). One can see that a spectrum reaching from 300 nm to 1500 nm is achieved, with peak power spectral density in the range of MW/nm and energy of 0.11 mJ, which are
Fig. 2. Waveguide characteristics (a), spectrum (b), temporal shape (c) and the phase of the compressed pulse (d) for the a silver-wall waveguide with radius of 15µm, argon-filled at 13 atm, coating layer of 40 nm, an average scatterer diameter of 100 nm, and length of 71.8 cm. The input pulse has a duration of 50 fs, peak intensity of 20 TW/cm2 , and central wavelength of 800 nm.
832
4 to 5 orders of magnitude higher than in microstructure fiber. Such dramatic increase of the spectral power density is brought about both by the larger radius of the waveguide and by the increased allowable input intensity related to the high ionization threshold of argon. The mechanism of supercontinuum generation is due to soliton fission and soliton emission, similar to microstructured fibers. To study SC generation for a higher nonlinerity, we consider a higher gas pressure which requires a lower waveguide radius. The loss [red curve in Fig. 2(a)] in the case of a 15-µm waveguide remains in the range of 1-10 dB/m. The GVD [green curve in Fig. 2(a)] is anomalous in the broad spectral range, despite the high argon pressure of 13 atm. In Fig. 2(b) the output spectrum is presented, exhibiting a spectral width of roughly two octaves. For propagation lengths in the range from 71 to 73 cm an ultrashort pulse is formed, as illustrated in Fig. 2(c) with smooth phase [Fig. 2(d)]. Its FWHM duration is ∼1.7 fs, and it has only a weak pedestal. Thus, we predict the generation of a coherent isolated sub-cycle pulse in a hollow waveguide without any external chirp compensation. Further, we have modeled the influence of quantum noise and calculated the first-order coherence [or visibility V(λ )] of the generated spectra. As illustrated by the green curve in Fig. 2(b) the whole spectrum is highly coherent with an average coherence of 0.98. Finally we studied the possibility to shift the SC to the VUV range using aluminium walls coated with SiO2 and pumping with the third harmonic of Ti:sapphire lasers at 266 nm. We have predicted that VUV supercontinua in the wavelength range of 150-500 nm can be generated in this way (not shown). 1
J. K. Ranka, R. S. Windeler, and A. J. Stentz, “Visible continuum generation in air-silica microstructure optical fibers with anomalous dispersion at 800 nm”, Opt. Lett. 25, 2527 (2000). 2 J. Herrmann et al., “Experimental evidence for supercontinuum generation by fission of higher-order solitons in photonic fibers”, Phys. Rev. Lett. 88, 173901 (2002). 3 P. Yeh, A. Yariv, and E. Marom, “Theory of Bragg fiber”, J. Opt. Soc. Am. 68, 11961201 (1978).
833
Part X
Frequency Combs and Waveform Synthesis
CEO-Phase Stabilized Few-Cycle Field Synthesizer Stefan Rausch1, Thomas Binhammer2, Anne Harth1, Franz X. Kärtner3 and Uwe Morgner 1,4 1
Institute of Quantum Optics, Leibniz University of Hannover, Welfengarten 1, 30167 Hannover, Germany E-mail: [email protected] 2 VENTEON Femtosecond Laser Technologies by Nanolayers, Maarweg 30, 53619 Rheinbreitbach, Germany 3 Research Laboratory of Electronics, Massachusetts Institute of Technology (MIT), Cambridge, MA, USA 4 Laser Zentrum Hannover (LZH), Hollerithallee 8, 30419 Hannover, Germany Abstract. We present an optical field synthesizer consisting of a CEO-phase stabilized octavespanning Ti:sapphire laser oscillator and prism-based pulse shaper allowing for full control of the electric field on a sub-femtosecond time-scale.
Introduction The electric field of few-cycle femtosecond laser pulses can be controlled on a subfemtosecond time-scale by manipulating the spectral phase and amplitude together with the carrier envelope offset (CEO) phase of the pulse train. An octave-spanning Ti:sapphire oscillator provides the spectral width required for direct CEO-phase stabilization and few-cycle shaping experiments. This field synthesizer is a versatile tool for coherent control and CEO-phase sensitive experiments.
Field Synthesizer The light source is a prism-less Ti:sapphire oscillator with specially designed output coupling mirror and double chirped mirror pairs [1], delivering an average output power of about 100 mW at a pulse repetition rate of 80 MHz. The octave-spanning output spectrum, shown in Fig. 1, supports a Fourier-limited pulse duration as short as 3.7 fs.
Fig. 1. Octave-spanning output spectrum supporting Fourier-limited pulses as short as 3.7 fs, shown on a linear (left axis) and logarithmic scale (right axis).
837
To obtain full control over the electric field of the few-cycle laser pulses the oscillator has to be CEO-phase stabilized. Due to the octave-spanning output spectrum no additional spectral broadening is necessary for the f-2f measurement approach [2]. The beat signal reveals a signal-to-noise ratio greater than 30 dB @ 100 kHz resolution bandwidth and is sufficient to stabilize the laser using a PLL locking electronic to control the pump power. This is schematically illustrated in Fig. 2 (upper right corner).
Fig. 2. Field synthesizer - overall working-scheme; The system consists of a CEO-phase stabilized Ti:sapphire oscillator, dispersion compensation and beam adaptation, prism-based double-LCD pulse shaper, SPIDER measurement system and personal computer.
The phase-stabilized pulses propagate through a prism-sequence to compensate for the highly dispersive prisms used in the 4-f geometry in which the double-LCD spatial light modulator is positioned in the Fourier-plane to independently manipulate the spectral phase and amplitude of the input femtosecond pulses. This shaper allows for highly efficient pulse manipulation covering the whole spectral octave. In the next step the pulses are analyzed with SPIDER. The computer-based analysis allows for a direct control of the shaper. If the required shape is generated and verified, the pulses can be guided towards the desired experiment. This system allows for flexible manipulation of the CEO stabilized few-cycle input pulses, and full control over the electric field of these pulses on the sub-cycle scale is achieved.
Results and Discussion The systems capability with respect to spectral phase and amplitude shaping is demonstrated in Fig. 3. The upper row shows a simple example: A shaped pulse with rectangular spectral amplitude and flat spectral phase. The calculated and measured time-domain sinc-shaped pulses match nearly perfectly. Various spectral shapes can be realized, such as Gaussian or triangular. Also sharp and flexible wavelength filtering, e.g. edge or band pass, can be achieved using this technique. The bottom part for Fig. 3 illustrates simulation results showing a double pulse sequence generated by simultaneously modulating the spectral phase and amplitude resulting in two clean sub-two-cycle double pulses in the time domain. The duration and shape of the pulses contained within the sequence is thereby similar to the input pulse, whereas the time delay is variable depending on the modulation frequency.
838
Fig. 3. Pulse shaping results; A) Sinc-shaped pulse formed by phase and amplitude shaping. B) Pulse sequence generated by simultaneous modulation of the spectral phase and amplitude (simulation results).
By using super-resolution techniques as presented in [3] pulse durations even below 3.7 fs are possible by accepting stronger wings in the pulse intensity profile. Such extreme pulse durations with controlled CEO-phase allow for novel experiments with sub-cycle scale resolution where the position and control of the electric field becomes crucial.
Conclusions In conclusion, we demonstrated a phase-stabilized few-cycle field synthesizer build from an octave-spanning laser oscillator, a prism-based double-LCD pulse shaper for independent phase and amplitude shaping, and SPIDER measurement system. This unique combination is capable of forming versatile pulse shapes and sequences in time and frequency domain down to the single cycle. Acknowledgements. The author thanks VENTEON Technologies by Nanolayers for the close cooperation. 1 2 3
Femtosecond
Laser
T.R. Schibli, O. Kuzucu, J.W. Kim, E.P. Ippen, J.G. Fujimoto, F.X. Kärtner, V. Scheuer, G. Angelow, in IEEE J. of Selected Topics in Quantum El., Vol. 9 (4), 990-1001, 2003. U. Morgner, R. Ell, G. Metzler, T. R. Schibli, F. X. Kärtner, J. G. Fujimoto, H. A. Haus, and E. P. Ippen, in Phys. Rev. Lett. 86, No. 24, 5462-5465, 2001. T. Binhammer, E. Rittweger, R. Ell, F.X. Kärtner and U. Morgner, in Opt. Lett. 31, 15521554, 2006.
839
High-power, mHz linewidth Yb:fiber optical frequency comb for high harmonic generation T. R. Schibli1, I. Hartl2, D. C. Yost1, M. J. Martin1, A. Marcinkevičius2, M. E. Fermann2, and J. Ye1 1
JILA, National Institute of Standards and Technology and University of Colorado, Boulder, CO 80309, USA E-mail: [email protected] 2 IMRA America, Inc., 1044 Woodridge Ave., Ann Arbor, MI 48105, USA E-mail: [email protected] Abstract. We present a fully phase-stabilized, high-power Yb:fiber frequency comb with record-low sub-mHz relative linewidths. Utilizing coherent pulse-addition inside a passive optical cavity, we achieve >3 kW average power and 100 fs pulse duration for high harmonic generation at >100 MHz pulse repetition rate. On a phosphor screen we visually observe up to the 21st harmonic generated in Xe at 136 MHz pulse repetition rate.
Introduction Growing demands for high average and peak powers in extreme nonlinear optics, attosecond pulse, and XUV-comb generation experiments can find a powerful solution in mode-locked fiber lasers. Fiber lasers have demonstrated the capability to reliably produce sub-100 fs pulse trains with unprecedented average powers. In this paper we report on a fully phase-stabilized Yb:fiber laser system capable of producing an ultracoherent fs-comb with sub-mHz relative linewidth at >10 W average powers and more than 3 kW with cavity-enhancement, empowering HHG at 136 MHz repetition rate.
Fig. 1. a) Setup of the fiber comb and its precise characterizations: SA: Saturable absorber; PZT: Piezo actuator; FBG: Fiber Bragg grating; PM LMA: Polarization-maintaining largemode-area fiber; PCF: Photonic crystal fiber; BS: Beam splitter; 698 nm: sub-Hz linewidth cwlaser for repetition rate locking; 1064 nm: cw-Nd:YAG laser locked to a Ti:sapphire frequency comb for the out-of-loop comparison. b) Two independent records of the out-of-loop heterodyne beat note between the stabilized 1064 nm laser and the Yb:fiber comb, showing an measurement limited linewidth of 950 µHz. Top inset: 15 s record of the out-of-loop beat note (dots) with a sinusoidal fit (line). Lower inset: Relative stability between the two combs.
840
Experimental Methods and Results Figure 1a) shows the experimental setup for the high-power fiber comb. A FabryPerot type Yb-similariton oscillator [1,2] modelocked with a sub-picosecond lifetime saturable absorber is used as a seed-source. The dispersion of the optical fibers inside the oscillator is compensated by a chirped fiber Bragg grating to a net-normal dispersion value. The oscillator generates ~40 nm of bandwidth centered at 1065 nm with >100 mW average power. This 136 MHz pulse train is amplified in a chirpedpulse amplifier (CPA) in which the pulses are first stretched in an anomalous thirdorder dispersion single-mode fiber to ~70 ps to avoid nonlinear phase shifts in the subsequent power amplifier. After amplification, the pulses are recompressed using two fused-silica transmission gratings, yielding 75 fs pulses with >10 W of average power. The spectrum is then broadened to more than one frequency octave (675 nm 1450 nm) in a 15 cm long photonic crystal fiber (PCF). After the PCF, a fraction of the continuum is sent into a nonlinear f-2f interferometer, which yields the carrierenvelope offset frequency f0. A (for fiber lasers) record low linewidth for the freerunning f0 signal of less than 10 kHz (inset of Fig. 2a) is routinely observed. Since the CPA is operating in a linear regime, the signal-to-noise ratio and the linewidth of f0 do not degrade at higher amplifier output powers (Fig. 2a). To phase-stabilize the frequency comb, we lock f0 to a Cs-clock using two loop filters, one of which controls the pump power of the fiber oscillator and the other the temperature of the fiber Bragg grating in the oscillator. When locked, the f0 beat note contains 90% of the RF power within a coherent, mHz line-width carrier. The repetition rate of the oscillator is stabilized by phase-locking the heterodyne beat note between one of the comb teeth around 698 nm and a sub-Hz linewidth external cavity laser diode that is stabilized to a high finesse optical cavity (F ~ 250,000) [3]. This sub-Hz cw-laser provides better short-term stability than any commercial microwave source available to date. The heterodyne beat note was locked to the same Cs-reference using two loop filters, one controlling a fast and the other a slow PZT actuator inside the fiber oscillator. The locked heterodyne beat note again contains ~90% of the RF power within the coherent carrier. To evaluate the comb’s performance we conduct a thorough comparison between this novel, high-power fiber comb and an octave-spanning Ti:sapphire laser-based frequency comb. Fig. 1a) shows the experimental setup for this comparison: The Ti:sapphire laser and the Yb:fiber laser are both locked to the same sub-Hz linewidth reference laser at 698 nm. f0 of each of the combs is independently stabilized using f2f setups. Utilizing the 698 nm cw-laser as a common optical reference allows us to compare the two combs without being limited by the stability of the reference laser. However, since the 698 nm reference and the two combs are each separated by tens of meters it is crucial to actively stabilize the fiber links between the three setups to a small fraction of an optical wavelength [4]. Finally, we use a Nd:YAG non-planar ring oscillator (NPRO) as a transfer laser by locking it to one of the comb teeth of the stabilized Ti:sapphire comb at 1064 nm. The light of this transfer laser is then delivered through a third noise-cancelled fiber link to the fiber-comb setup. The heterodyne beat note between the stabilized Nd:YAG laser and the fiber comb is recorded with an FFT analyzer and a frequency counter. From the FFT analyzer we find measurement-limited linewidths of less than 1 mHz (Fig. 1b). The top inset of Fig. 1b) shows a 15 s record of the out-of-loop beat note (dots) with a sinusoidal fit (line). The lower inset shows the Allan deviation of the out-of-loop beat. By coherent addition of the fs-pulses produced by this laser system, we have obtained more than 3 kW average power and 100 fs pulse duration inside a passive
841
optical cavity [5]. Inside the passive enhancement cavity we reach intracavity intensities as high as 3·1014 W/cm2. To confirm these record-high levels of intensity at >100 MHz repetition rates, we measure electric currents through the laser induced plasma for several noble gases (see Fig. 2b). A theoretical calculation for Kr (green, dotted line; calculation by Robin Santra, Argonne National Laboratory, IL) shows an excellent agreement with the measured data. In addition we observed strong highharmonic radiation produced in Xe at 136 MHz repetition rate (inset Fig. 2b).
Fig. 2a) Free running f0 beat notes obtained from a nonlinear f-2f interferometer at four different levels of average output powers (4.7W, 6.9W 9.1W and 10W). Each trace shows an average of 25 single scans at 100 kHz RBW (20 s accumulation per trace.) The four traces are offset in x-direction for clarity. Inset: Free running beat note on a linear scale with 10 kHz RBW (sweep time: 65 ms.) b) Electrical currents through the laser-induced plasma for various noble gases at 750 mTorr under 10 V/mm bias as a function of laser power and peak intensity. The perfect agreement with a theoretical calculation (dotted line) and the clear onset of saturation confirm the 3·1014 W/cm2 peak intensity levels. The inset shows the fluorescence of high-harmonic radiation produced in Xe at 136 MHz repetition rate on a sodium salicylate screen after separation on a VUV grating. The power per harmonic across the VUV range (5th – 11th harmonic) exceeds 1 µW while the average power for the 13th – 19th harmonic is ~100 nW/harmonic. The 21st harmonic marks the beginning of the cut-off region.
Conclusions We have demonstrated that Yb:fiber lasers in conjunction with chirped pulse Yb:fiber amplifiers are not only perfectly suited to produce very high average and peak powers, but are also capable of producing ultra-precise optical frequency combs. In conjunction with coherent pulse addition inside a passive optical cavity we observed strong HHG at 136 MHz, yielding all odd-harmonics up to the 21st. This work certainly marks an important milestone towards coherent as-comb generation in the XUV domain as well as highly-nonlinear light-matter interaction experiments that require precise control over the optical phase. 1
2 3 4 5
842
M. E. Fermann, "Ultrafast fiber oscillators', in Ultrafast lasers:technology and applications,” (eds. M. E. Fermann, A. Galvanauskas, G. Suchaeds) Marcel Dekker, New York, 2003. F. İ. Ilday, et al., Phys. Rev. Lett., 92, 213902, 2004. A. D. Ludlow, et al., Opt. Lett., 32, 641-643, 2007. S. M. Foreman, et al., Rev. Sci. Instrum. 78, 021101/1-25, 2007. I. Hartl, et al., Opt. Lett, 32, 2870-2872, 2007.
High Harmonic Frequency Combs for High Resolution Spectroscopy A. Ozawa1, J. Rauschenberger1, 2, Ch. Gohle1, M. Herrmann1, D. R. Walker1 V. Pervak2, A. Fernandez1, A. Apolonski2, R. Holzwarth1, F. Krausz1, 2 T. W. Hänsch1, 2 and Th. Udem1 1
Max-Planck-Institut für Quantenoptik, Hans-Kopfermann-Strasse 1, 85748 Garching, Germany E-mail: [email protected] 2 Department für Physik der Ludwig-Maximilians-Universität München, Am Coulombwall 1, 85748 Garching, Germany Abstract. Intracavity high harmonic generation is demonstrated in an external cavity, seeded by a Ti:sapphire mode-locked laser at a repetition rate of 10.8MHz. Harmonics up to 19th order at 43 nm were observed with plateau harmonics at the µW power level.
Introduction The extreme ultraviolet (XUV) frequency comb technique is expected to play an important role in extending high-resolution laser spectroscopy to the XUV region where many interesting transitions are located. Conventional frequency combs generated at infrared wavelengths can be frequency converted to XUV wavelengths with high harmonic generation (HHG). However, until recently, a low repetition rate in the kHz range was required to generate sufficient pulse energy to produce high harmonics. The resulting frequency comb is far too dense to be useful for highresolution spectroscopy. Direct frequency comb spectroscopy requires the separation of the modes, i.e., the repetition rate, to be much larger than the measured linewidth. This problem has been solved by employing intracavity HHG, achieving the required intensity by resonantly enhancing the pulse train from a mode-locked laser without compromising on the repetition rate [1-4]. Unfortunately, the powers generated in the XUV with this technique have been far too low for a reasonable excitation rate of a narrowband XUV transition. We report a dramatic enhancement of the XUV output power by almost 5 orders of magnitude. This is achieved by an elaborate dispersion compensating scheme and by reducing the repetition rate to 10.8MHz, which is still sufficient to resolve narrowband atomic transitions.
Experimental Methods Our mode-locked laser has a long cavity and runs in the positive GDD regime. Its linear cavity has a length of 13.9m which is achieved by using a Herriott-type multipath delay line. The output of about 1.5W of average power is compressed by a pair of LaK16 (SCHOTT) prisms (Fig. 1). The enhancement cavity is located in a vacuum setup and consists of 10 mirrors including 8 quarter-wave-stack mirrors, 4 of which are curved, 2 chirped mirrors, and one input coupler. Two homemade chirped mirrors are used to compensate the dispersion of the Brewster XUV output coupler made of sapphire and the contribution from the remaining cavity mirrors. Firstly, we design one chirped mirror by using the estimated total dispersion of the cavity. Then the
843
residual dispersion is measured by analyzing the spectral cavity enhancement as described in [5]. Using this information, we generate several different coating designs with the correct dispersion properties and choose the design which is the least sensitive to the coating errors. With this dispersion compensating scheme, we obtain broad enhancement over 40nm. The circulating power is determined to be 100W by measuring the residual transmission through one of the highly reflecting cavity mirrors. This corresponds to an average power enhancement of 100. Intracavity pulsewidth is measured to be 57fs. Asymmetric focusing with two differently curved mirrors (radii of curvature, 0.1m and 0.24m) is used to focus to a waist size of w0= 13 µm into the gas target emerging from a nozzle. At the focus, a peak intensity of > 5 1013 W/cm2 is obtained. When injecting Xe, Ar, or air through the gas nozzle, a fluorescing plasma can be observed. The gas flow is estimated to be 1 10-2 mbar l/s. The XUV output is used to illuminate the entrance slit of a scanning grating spectrometer (Jobin-Yvon, LHT30) that has an estimated resolution of 1.4 nm and is equipped with a channeltron detector (Burle, CEM4839). The pulse compressor, gas flow rate, nozzle position, and carrier-envelope offset frequency are optimized to maximize the XUV signal. The measured XUV spectrum is divided by the specified wavelength dependent grating diffraction efficiency. In order to independently calibrate the XUV power, the (spectrally unresolved) total power is measured with a calibrated Si photodiode (IRD, AXUV20HS1) placed directly after the Brewster XUV output coupler. A 150µm thickness Al filter (Lebow) is used to remove the residual reflection of the fundamental laser beam. The transmission of the Al filter is then measured by comparing the XUV spectra with and without it. Knowing the undistorted XUV spectrum, the total power and the spectrally resolved filter transmission, we determine the absolute spectral power density of outcoupled XUV beam.
Results and Discussion The obtained spectral power density of XUV beam is shown in Fig. 2. A µW power level is obtained at plateau harmonics. Compared to previously reported powers [1], this represents an improvement of 104-105. High harmonics up to the 19th order (41 nm) are clearly observed, which agrees with the calculated cutoff located between the 13th and 15th harmonics. In addition to the odd harmonics, a broad peak at 104 nm is observed. Its origin is not yet understood but appears to be related to the occupancy of excited bound states of the Xe atom [1]. With the power level obtained here, precision spectroscopy in the XUV comes into reach for the first time. An example could provide the 1S-2S transition in He+. In this system, the ground-state ions may be excited to the 2S state by two-photon absorption by using the 13th harmonic around 60.8 nm. A third photon can further ionize to He2+ which can be accumulated in the ion trap and detected with unity efficiency. If the full 0.84µW generated so far could be focused on the He ion with a waist size of 0.5 µm, an ionization rate close to 1Hz (0.88Hz) could be obtained. This is a typical rate for a high precision spectroscopy experiment on trapped ions. Since there is virtually no background for this type of detection, we believe precision spectroscopy is realistic even with count rates below 1Hz.
844
Fig. 1. Experimental Setup
Fig. 2. Spectrum power density of outcoupled high harmonics
Conclusions Generation of high power XUV frequency combs with µW power level is demonstrated. An elaborate dispersion compensation scheme and the use of a moderate repetition rate allow for this significant improvement in output power. With this power level and repetition rate, high-resolution spectroscopy in the extreme ultraviolet (XUV) region becomes conceivable. Our improved XUV source demonstrates that generating high harmonics in external enhancement cavities has now moved beyond proof-of-principle experiments to become a tool capable of performing spectroscopy at previously inaccessible wavelengths. 1 2 3 4 5
Ch. Gohle, Th. Udem, M. Herrmann, J. Rauschenberger, R. Holzwarth, H.A. Schuessler, F. Krausz, and T. W. Hänsch, in Nature, 436, 234, 2005. R. J. Jones, K. D. Moll, M. J. Thorpe, and Jun Ye, in Phys. Rev. Lett., 94, 193201, 2005. I. Hartl, T. R. Schibli, A. Marcinkevicius, D. C. Yost, D. D. Hudson, M. E. Fermann, and Jun Ye, in Opt. Lett., 32, 2870, 2007. D. C. Yost, T. R. Schibli, and Jun Ye, in Opt. Lett. 33, 1099, 2008. A. Schliesser, Ch. Gohle, Th. Udem and T. W. Hänsch, in Opt. Express, 14, 5975, 2006.
845
Ultrafast double pulse parametric amplification for precision Ramsey metrology D.Z. Kandula, A. Renault, C. Gohle, A.L. Wolf, S. Witte, W. Hogervorst, W. Ubachs, and K.S.E. Eikema Laser Centre, Vrije Universiteit Amsterdam, De Boelelaan 1081, 1081HV Amsterdam, Netherlands E-mail: [email protected] Abstract. An optical parametric chirped pulse amplifier system for pulse pairs is presented. The differential phase stability of the pulse pairs is 20 mrad, giving good prospects for high resolution Ramsey spectroscopy in the extreme ultraviolet.
Introduction The convergence of ultrafast optics and high precision optical frequency metrology has led to fascinating new possibilities. The advent of femtosecond laser optical frequency combs (OFC) [1] turned the determination of arbitrary optical frequencies into a routine task. This enabled new high accuracy test of quantum electrodynamics (QED) [2] and the determination of tight lower bounds on the current time dependence of fundamental physical constants [3] as well as it opened the possibility of creating optical frequeny standards which have the potential to give 18 digits of accuracy. Simultaneously OFC opened the door to attosecond physics [4]. In recent years a new branch of high-precision physics is emerging from the crossfertilization between laser frequency metrology and ultrafast technology. Ultrashort pulses provide an extremely high peak power suitable for relatively efficient nonlinear frequency conversion of the available power into frequency ranges where no laser sources exist. If these pulses originate from a phase stable OFC, they interfere to create a frequency comb inside the THz bandwidth of a single femtosecond pulse. Each of the modes of this frequency comb can be extremely narrow and are suitable as a probe for high resolution spectroscopy experiments. As the nonlinear conversion can preserve this comb structure, high-precision experiments in the extreme ultraviolet spectral range come into reach, and first proof of principle experiments have been perfomed to demonstrate this idea [5,6]. Quite a few interesting atomic transitions exist in the extreme ultraviolet (XUV) wavelength range below 100 nm. As an example, the 1s2 1S0 – 1s4p 1P1 transition in atomic helium at 52 nm wavelength could be used to improve the value of the ground state Lamb shift in this atom by at least an order of magnitude. One specifically intriguing possibility is to determine the frequency of the 1s-2s transition in He II, a 2-photon transition at 60 nm. He II is a Hydrogen like system which is specifically simple to model in QED, and comparisons between theory and experiment could in principle be carried out at an extreme level of accuracy. Compared to Hydrogen, the ground state energy in He II is four-fold lower which leads to stronger relativistic effects. Specifically, the ground state Lamb shift (i.e. the QED correction to the Dirac energy level structure) is 16 fold stronger than in Hydrogen. Currently, there exist two approaches to coherent XUV generation for high resolution spectroscopy. The first possibility is to enhance the entire bandwidth of a MHz repetition rate femtosecond frequency comb laser in a passive optical resonator
846
in order to achieve the intensities required for high order harmonic generation (HHG) [7]. On the other hand it can be shown that optical Ramsey type spectroscopy, where only two identical pulses with a variable but precisely known delay and carrier envelope phase are used to excite the atom, is essentially equivalent. Such pulse pairs can be obtained by pulsed amplification of an OFC seed source. While the former is a continuous wave technique which avoids many systematic effects due to transients in the optically active materials in the setup, we prefer the latter as it can rely on cutting edge non collinear optical parametric chirped pulse amplifier (NOPCPA) technology [8]. This offers more than three orders of magnitude larger peak power, so that the nonlinear conversion becomes a lot more efficient, facilitating the actual spectroscopy experiment.
Phase-stable double pulse NOPCPA The double-pulse NOPCPA is based on the single-pulse system presented previously [9]. To achieve double-pulse amplification, the pump pulses are split, delayed with respect to each other and superimposed in a symmetrized relay imaged MachZehnder interferometer. In this way two identical (in spatial profile) pump pulses with a delay of 6.6 ns are generated. This delay matches the time between two pulses from our 151 MHz OFC seed oscillator. In this way we can amplify a pair of subsequent pulses from our OFC to the millijoule level with a pulse duration down to 10 fs. This pair can be upconverted into the XUV spectral and used for a Ramsey experiment with a spacing of the Ramsey fringes of 151 MHz.
Fig. 1. NOPCPA and Mach-Zehnder interferometer for measurement of the differential phase shift accumulated during the amplification in a NOPCPA, and the setup used to test the reliability of the measurement (dotted beam line). BS: Beam splitter, PC: Pockels cell, sPC – slow Pockels cell, NG: neutral grey filter, Comp.: compressor, G: 1200 l/mm grating
847
If the wavefront and the intensity of the two pulses is not identical, the two amplified pulses can acquire different phase shifts during the amplification process. This would lead to a shift of the Ramsey fringes with respect to the original comb mode positions. Such a shift is multiplied up by the upconversion process, which means that it needs to be accurately measured and monitored. Figure 1 shows a schematic of the setup used to monitor differential phase shift induced by the NOPCPA system on the amplified pulse pair. It consists of a MachZehnder interferometer including the NOPCPA, the stretcher and compressor in one arm. The output of this interferometer is analysed using spectral interferometry. In order to maintain a good contrast in the interference signal, the output of the interferometer is spatially filtered using a single mode optical fiber and temporally gated to remove the pulses from the seed oscillator which have no amplified counterpart using a double passed Pockels cell (PC1). Like this we obtain a fringe contrast of almost unity with a background from leaking oscillator pulses of less than 10-4. The two remaining pulses are spatially split using a second Pockels cell so that two interferograms can be recorded, one for each pulse. The information on the differential phase shift is now contained in the relative position of the two spectral interferograms. As only differential phase shifts are of importance for the spectroscopy the absolute position of the two interferograms is not important and the interferometer needs only to be stable on a 10 ns time scale (the separation between the pulses). This is passively guaranteed even for the almost 10 meter arm length of the interferometer.
Results The accuracy of the phase measurement scheme was found to be better than 10 mrad. The single shot stability of the interferometer output is on the same order of magnitude, so that shot to shot fluctuations can be accurately measured and taken into account. The rms phase stability of the NOPCPA system was found to be 20 mrad. This is already sufficient for the observation of clear Ramsey fringes even at the 15th harmonic of the fundamental signal frequency at 800 nm, that is required for the 1s2 1 S0 – 1s4p 1P1 transition in He I. The achievable accuracy for the lamb shift due to the amplifier phase shift is therefore at least 15 MHz which would be a threefold improvement over previous results. 1 2 3 4 5 6 7 8 9
848
T. Udem, R. Holzwarth and T.W. Hänsch, Nature. 416, 233, 2002. M. Niering, R. Holzwarth et al., Phys. Rev. Lett., 84, 5496, 2000. M. Fischer, N. Kolachevsky et al., Phys. Rev. Lett., 92, 230802, 2004. A. Baltuska, Th Udem et al., Nature, 421, 611, 2003. S. Witte, R.T. Zinkstok, W. Ubachs, W. Hogervorst and K.S.E. Eikema, Science, 307, 400, 2005. R.T. Zinkstock, S. Witte, W. Ubachs, W. Hogervorst and K.S.E. Eikema, Phys. Rev. A, 73, 061801, 2006. C. Gohle, Th. Udem, M. Herrmann, J Rauschenberger, R. Holzwarth, H. A. Schuessler, F. Krausz and T.W. Hänsch, Nature, 436, 234, 2005. S. Witte, R. T. Zinkstok, A.L. Wolf, W. Hogervorst, W. Ubachs and K.S.E Eikema, Opt. Express, 14, 8168, 2006 A. Renault, D.Z. Kandula, S. Witte, A.L. Wolf, R.T. Zinkstok, W. Hogervorst and K.S.E Eikema, Opt. Lett., 32, 2363, 2007
Towards Versatile Coherent Pulse Synthesis using a Femtosecond Laser and Synchronously Pumped Optical Parametric Oscillator Barry J. S. Gale, Jinghua Sun, and Derryck T. Reid Ultrafast Optics Group, School of Engineering and Physical Sciences, Heriot-Watt University, Edinburgh EH14 4AS E-mail: [email protected] Abstract. Pulses from a femtosecond optical parametric oscillator and its Ti:sapphire pump laser were phase-locked as a prerequisite to coherent synthesis from different wavelengths. Mutual coherence was demonstrated using spectral interferometry and cross-correlation.
Introduction Pulse compression and shaping are limited by the available spectral components of a single laser. Micro-structured fibres are capable of broadening a laser spectrum to a supercontinuum sufficient to support a pulse of 2.6 femtosecond duration [1], but they require high pulse power and have little flexibility to obtain variable spectral shapes. Coherent synthesis of the outputs from multiple laser sources offers a direct and intuitive way to achieve the desired spectral components [2-4], but it requires not only carrier-envelope phase slip (CEPS) locking, but also precise synchronization of pulse repetition frequencies (Frep) to ensure the coherence of the combined pulses over significant time periods. An alternative route uses multi-colour pulses from an optical parametric oscillator (OPO). This requires active control of only the carrierenvelope phase slip (CEPS) frequencies, because of the inherent passive synchronization between the pump and all OPO outputs, which makes coherent synthesis easier and more robust. We previously generated coherent outputs from a femtosecond OPO centered at 1240 nm and 1330 nm [5]. Here we report the achievement of mutual coherence between the pump and frequency-doubled signal pulses, both at 780 nm, as a prerequisite for coherent synthesis from different wavelengths.
Experiment The 1.25 W output from a self-mode-locked Ti:sapphire laser was split by mirror M1 (see Fig. 1), with 1 W and 0.25 W used to pump a quasi-phase-matched MgO:PPLN OPO and directed to an f-to-2f non-linear interferometer respectively. The latter employed photonic crystal fiber to generate a super-continuum reference spectrum against which the pump and signal CEPS frequencies were measured. The OPO was operated close to degeneracy with a signal (idler) wavelength of 1560 (1640) nm. The non-phase-matched sum-frequency mixing (SFM) between pump and signal (529 nm), which was reflected from the PPLN crystal rear surface, was collected through mirror M4 and heterodyned with the pump super-continuum in a second interferometer to obtain the signal CEPS frequency. Signal secondharmonic (SH) light with an average power of 15 mW at 780 nm was generated by an intra-cavity frequency-doubling BBO crystal, and collected through mirror M7. Pump
849
pulses at 780 nm with average power 10 mW were selected with an interference filter (IF) from the residual pump leaking through OPO mirror M5, and after suitable delay combined with signal SH pulses in polarizing beam splitter PBS3. The pump and signal CEPS frequencies were locked to 50 MHz (Frep/4) and 25 MHz (Frep/8) respectively, so that the signal SH CEPS frequency was also 50 MHz. The locking bandwidths were 1.2 kHz and 2.7 kHz respectively. The CEPS frequencies of the pump and signal were controlled using a travelling-wave acoustooptic modulator (not shown) to fine-tune the power from the Verdi laser, and a piezoelectric translator (PZT) mirror M6 to tune the OPO cavity length.
Fig. 1 Optical layout. BBO, β barium borate crystal. FM, flipper mirror. IF, interference filter. P, polariser. PBS1-3, polarising beam splitters. PZT1,2, piezoelectric transducers.
Results When the CEPS frequencies of the pump and signal SH were locked to the same value (50 MHz), spectral interference between the two outputs at 780 nm showed deep fringes (Figure 2a) on the screen of the optical spectrum analyzer (see Figure 1) during a scan time of 4 ms, indicating strong coherence between the sources. Fringe visibility was limited by imperfect spectral overlap, spectrometer resolution (0.3 nm), and small differences in the beam size and divergence. Scanning PZT2 (see Fig. 1) gave an interferometric second-order cross-correlation trace with high-contrast fringes (Figure 2c) in a sweep time of 20 ms, which was stable for several seconds. The 6:1 contrast ratio (as compared with 8:1 for autocorrelation) does not indicate lack of coherence, only that the input powers were not exactly balanced. Fourier transformation of the in-loop phase-locking errors showed that the pulses from both the Ti:sapphire laser and the OPO had sub-ms coherence times individually, but because the OPO inherited the pump laser noise, the mutual coherence time could be much longer. A phase-noise power spectral density analysis revealed that both oscillators possessed significant acoustic noise at 220 Hz.
850
.6 .4 .2
8
Intensity (a. u.)
.8
Locked Unlocked
(a)
6 4 2
0 -400
1.0
8
.8
(b)
2s p
.6 .4 .2 0.0 765
Locked Unlocked
(c)
0.0 770 772 774 776 778 780 782 784
Intensity (a. u.)
Intensity (a. u.)
Intensity (a. u.)
1.0
-200
0
200
400
0
5
10
(d) 6 4 2 0
770
775
780
Wavelength (nm)
785
790
-10
-5
Delay (fs)
Fig. 2. (a) Spectral interference observed in a 4 ms scan between SH signal and pump pulses with CEPS frequencies locked to 50 MHz (solid line) and unlocked (dashed line). (b) Typical constituent spectra, recorded separately from measurement (a). 2s, SH of the signal. p, the pump. (c) Second-order cross-correlation between SH signal and pump pulses when their CEPS frequencies were locked to 50 MHz (lighter lines) or unlocked (darker lines). (d) Expanded detail of fringes around zero delay.
Conclusions We have generated two coherent pulse sequences by locking the CEPS frequencies of pulses from a femtosecond OPO and its pump laser to sub-harmonics of their common repetition rate. Spectral interferometry and cross-correlation indicated that coherence between the pulses was maintained for at least 20 ms. Such a level of coherence has been shown sufficient for synthesizing waveforms [3]. Acknowledgements. We gratefully acknowledge support for this research from Coherent, Inc, and the Engineering and Physical Sciences Research Council, UK. 1 2 3 4 5
E. Matsubara, K. Yamane, T. Sekikawa, and M. Yamashita, J. Opt. Soc. Am. B 24, 985 (2007). T. W. Hänsch, Opt. Commun. 80, 71 (1990) R. K. Shelton, L. S. Ma, H. C. Kapteyn, M. M. Murnane, J. L. Hall and J. Ye, Science 293, 1286 (2001) Y. Kobayashi, H. Takada, M. Kakehata and K. Torizuka, Appl. Phys. Lett. 83, 839 (2003) J. H. Sun, B. J. S. Gale and D. T. Reid, Opt. Lett. 32, 1396 (2007)
851
Frequency comb spectroscopy on calcium ions in a linear Paul trap A.L. Wolf1 , S.A. v.d. Berg2 , C. Gohle1 , E.J. Salumbides1 , W. Ubachs1 , K.S.E. Eikema1 1 2
Laser Centre Vrije Universiteit, De Boelelaan 1081, 1081 HV Amsterdam, The Netherlands NMi van Swinden Laboratorium BV, Thijsseweg 11, 2629 JA Delft, The Netherlands
Abstract. To add to the debate on a possible variation of the fine structure constant, we perform frequency comb spectroscopy on laser cooled (calcium) ions in a linear Paul trap.
Introduction In 1999, Webb et al. claimed a possible variation of the fine structure constant α of ∆α/α = 0.57(10) × 10−5 over cosmological timescales [1]. These results are based on a comparison of the wavelengths of atomic resonances between absorption lines in quasar spectra observed at high redshift, and the current laboratory values. Since spectral lines have a different dependence on a change in α, such an analysis can be used to find a non-zero value for ∆α/α over time spans of many billion years. Currently, many ionic lines that are interesting for this analysis are only known to a precision of a few tens of MHz. By using cooled and trapped ions, the lines can be measured to much higher precision. An interesting ion for comparison to quasar spectra is Ca+ . We have developed a setup for trapping and laser cooling calcium ions to Coulomb crystallization, to obtain a much improved transition frequency for the 4s 2 S1/2 − 4p 2 P1/2 transition.
Experimental methods Trapping and laser cooling of calcium ions. Calcium ions are trapped and cooled in a linear Paul trap. Atomic calcium is first ionized in the trapping region using a frequency doubled Ti:Sapphire laser (422 nm) and a frequency tripled Nd:YAG laser (355 nm). A 3.3 MHz radio-frequency potential for trapping the ions is supplied by a waveform generator, resonantly enhanced by a helical resonator. A grating stabilized diode laser at 397 nm is used for laser cooling on the 4s 2 S1/2 −4p 2 P1/2 transition, while an additional diode laser at 866 nm is used for repumping of the ions that leak into the 3d 2 D3/2 state. Fluorescence from the trapping region is imaged onto a photomultiplier for signal detection. On scanning the cooling laser over the transition, an asymmetric fluorescence signal is detected (see Figure 1): at energies below the transition the ions are cooled, above the transition ions are heated, and hence blown out of the laser focus, leading to a loss of fluorescence. The phase transition to a Coulomb crystal is clearly visible as a sudden decrease of the fluorescence on the low frequency side of the transition [2]. To make the spectroscopy independent of the coolig dynamics, the cooling laser frequency is fixed, while an additional weak probe laser is used for the spectroscopy.
852
Fig. 1. Measured fluorescence by scanning the 397 nm laser cooling diode laser over the Ca+ 4s 2 S1/2 − 4p 2 P1/2 transition
Spectroscopy on calcium ions The 4s 2 S1/2 − 4p 2 P1/2 transition is measured on the laser cooled calcium ions using a frequency-doubled diode laser. This beam necessarily has a low intensity (∼ 1µW) to prevent heating of the ion cloud by the spectroscopy laser. To keep the ions cold during the measurement, the ions are alternately cooled and probed. The probe laser is calibrated by referencing it to a frequency comb laser. An interference beat note is generated between the spectroscopy laser and the comb laser modes, by overlapping the near-infrared fundamental output of the spectroscopy laser with the output of the frequency comb laser.
Fig. 2. An example of the measured fluorescence spectrum for the 4s 2 S1/2 − 4p 2 P1/2 transition in Ca+ and the corresponding Voigt fit
853
Results and Discussion An example of a calibrated scan over the 4s 2 S1/2 − 4p 2 P1/2 transition in Ca+ , after subtraction of the background and correction for ion loss during the scan, is shown in figure 2. The cooling is insufficient to reach the natural linewidth of the transition, hence a Voigt profile is fitted. The width of the Lorentzian part is fixed to the natural linewidth of the transition (22.4 MHz), while the width of the Gaussian part is fitted. The Gaussian FWHM of the line varies from scan to scan (depending on the cooling conditions), typically between 28 and 43 MHz, corresponding to an average temperature of T ≈ 0.2K. Effects that can introduce systematic and statistical errors have been investigated (see table 1). The 40 Ca+ 4s 2 S1/2 − 4p 2 P1/2 transition follows from the statistical average of the measurements, corrected for the measured shifts, which in total adds up to 755 222 766.2(1.7) MHz. The present result is consistent with the most accurate previously reported value of f = 755 222 740(60) MHz [3]. Table 1. Measured systematic shifts and uncertainty budget (1σ ). All values in MHz.
Effect Zeeman AC Stark repumper AC Stark spectroscopy laser RF Stark effect Comb calibration Statistics Total
Shift(MHz) 0.0 -0.4 -0.4 0.0 0.0 -0.8
1σ Uncertainty (MHz) 0.0 0.6 0.8 1.2 0.2 0.6 1.7
Conclusions We have measured the 4s 2 S1/2 −4p 2 P1/2 transition in 40 Ca+ to be at 755 222 766.2(1.7) MHz, in a laser cooled ion crystal. The level of accuracy, at ∆λ /λ = 2 × 10−9 , is such that for comparison with state-of-the-art astrophysical data, the laboratory value can be considered exact. The technique employed in this work will be used for spectroscopy on other transitions and ions in the near future. We intend to measure the 4s 2 S1/2 − 4p 2 P3/2 and the 4s 2 S1/2 − 5p 2 S1/2 transitions in Ca+ , the latter using direct two-photon frequency comb spectroscopy [4,5] Acknowledgements. This project is financially supported by the Foundation for Fundamental Research on Matter (FOM), and the Nederlands Meetinstituut (NMi). These contributions are gratefully acknowledged. 1 2 3 4 5
J.K. Webb, V.V. Flambaum, C.W. Churchill, M.J. Drinkwater, and J.D. Barrow, Phys. Rev. Lett. 82, 884, 1999. F. Diedrich, E. Peik, J.M. Chen, W. Quint, and H. Walther, Phys. Rev. Lett. 415, L7, 2004 U. Litz´en, Private communication, 2008 A. Marian, M.C. Stowe, D. Felinto, and J. Ye, Phys. Rev. Lett. 95, 023001, 2005 S. Witte, R.Th. Zinkstok, W. Ubachs, W. Hogervorst, K.S.E. Eikema, Science 307, 400, 2005
854
Generation of octave-spanning Raman comb stabilized to an optical frequency standard M. Katsuragawa1, 2, F. L. Hong3, 4, M. Arakawa1, and T. Suzuki1, 2 1
Department of Applied Physics and Chemistry, University of Electro-Communications, 1-5-1 Chofugaoka, Chofu, Tokyo 182-8585, Japan 2 PRESTO, JST, 4-1-8 Honcho Kawaguchi, Saitama, Japan 3 National Institute of Advanced Industrial Science and Technology, 1-1-1, Umezono, Tsukuba 305-8563, Ibaraki, Japan 4 CREST, JST, 4-1-8 Honcho Kawaguchi, Saitama, Japan E-mail: [email protected] Abstract. We show a novel octave-spanning comb generation having precise frequencyspacing of a Raman transition. It is shown that the carrier-envelope-offset of the Raman comb is precisely controlled by stabilizing the driving lasers to an optical-frequency-standard.
Introduction A novel approach using an adiabatic Raman technique to generate ultrashort pulses, has been extensively discussed. The technique relies on the production of maximal coherence through the adiabatic Raman process and the resultant collinear broad Raman generation [1-3]. The generated sidebands are mutually phase coherent and have a wide, equidistant frequency-spacing. Recently, the generation of an ultrahighrepetition-rate train of monocycle pulses has been demonstrated by Fouriersynthesizing such Raman sidebands [4]. It has also been shown based on wellestablished techniques of evaluating a temporal waveform, an intensity autocorrelation [5] and a frequency-resolved optical-gating [6], that high quality ultrashort pulses with potential as an actual light source can be constituted from such Raman sidebands. Here, we demonstrate generation of octave-spanning Raman sidebands with accurate control of the carrier envelope offset (ceo) by stabilizing the two-wavelength driving-laser radiations to an optical frequency standard. The realized broad Raman sidebands have potential to produce monocycle ultrashort pulses with an absolutephase control.
Experimental Methods The conceptual scheme and the experimental setup are illustrated in Fig. 1. The central part of the driving laser system is a dual-wavelength, injection-locked, pulsed Ti:sapphire laser [7]. The key performance of this laser is to emit two-wavelength, transform-limited, nanosecond-pulses from a single laser resonator, enabling the perfect overlap of the two nanosecond-pulsed-outputs (typically 6 ns full-width, halfmaximum) in both time and space. This performance is realized by simultaneously injecting the two-wavelength continuous wave laser radiations as seeds, which are generated from the two independent external-cavity controlled diode-lasers,
855
Fig. 1. Conceptual scheme for absolute-phase control and experimental setup.
respectively. In the present driving laser system, these two diode lasers were further phase-locked with each other through the femtosecond-laser optical-frequency-comb. This femtosecond-laser optical-frequency-comb had an absolute frequency stability equivalent to an optical frequency standard by employing both of the ceo control with an well-known f-2f self-referencing technique and the phase-locking of the comb to the optical frequency standard. As for the optical frequency standard, we employed the iodine stabilized Nd:YAG laser, having a stability better than 2 x 10-14 for 60-s average (absolute frequency uncertainty: 8 x 10-13) [8]. These absolute and relative frequency stabilities were transferred to the two independent diode lasers via the phase-locking procedures. The Raman sidebands were generated by adiabatically driving the pure rotational transition (v = 0, J = 2 ← v = 0, J = 0) in para-hydrogen. The wavelengths of the two driving-lasers, Ω0, Ω-1, were set to 783.9331 and 806.3312 nm, respectively, slightly detuned on the positive side (as shown in Fig. 1) by 700 MHz from the Raman-resonance (10.6235 THz) to satisfy the adiabatic condition. In order to realize an octave-spanning Raman sideband generation, in addition to these two drivinglasers, we further introduced the second harmonic, 2Ω-1 of the driving laser Ω-1. The Raman sidebands are generated from the driving lasers, Ω-1, Ω0 and simultaneously from the second harmonic, 2Ω-1. It should be noted that this sideband generation scheme also provides us the ceo information through an overlap of the both sidebands, similarly to the f-2f self-referencing technique in a femtosecond laser comb.
Results and Discussion Figure 2 shows stability of the phase locking for the diode laser, Ω0, measured with a frequency counter. The stability was better than 4 mHz in 500 s. It was also confirmed that the other diode laser, Ω-1 had nearly same frequency stability. Keeping this stability of the two diode lasers as seeds, we carried out the Raman sideband generation. The broad sidebands over an octave-spanning, 371 ~ 941 nm, were clearly seen with high beam quality. We picked up the sidebands at 487 nm and introduced them into a high-speed detection system having a frequency response broader than 10 GHz. The beat signal corresponding to the ceo of the Raman
856
sidebands was observed directly in time domain as shown in the inset of Fig. 3. The observed clear beat signal reveals that the sidebands generated from the fundamental driving lasers and the second harmonic, were phase-coherent with each other, across the whole beam cross section. Figure 3 plots the ceo estimated from the observed beats as a function of the relative frequency of the seed diode laser, Ω0, which was controlled with a synthesizer set in the phase-locking loop. The red line represents the Fig. 2. Frequency stability of diode laser, Ω0. theoretically predicted line in this ceo control, proportional to 36 times of the tuning frequency of Ω0. It is clearly shown that the ceo is accurately controlled along the theoretical red line. The error bars indicate the standard deviation for 100 measurements. When we further tuned the absolute frequency of Ω0, close to the zero ceo condition, we observed the temporal waveform with a smooth envelope as expected. This implies that the sidebands from fundamental and that from SHG were overlapped in phase at least in this time Fig. 3. Beat frequency corresponding to ceo of scale. Raman sidebands vs the relative frequency tuning of seed diode laser, Ω0.
1 2
S. E. Harris and A. V. Sokolov, Phys. Rev. A 55, R4019-4022 (1997). A. V. Sokolov, D. R. Walker, D. D. Yavuz, G. Y. Yin, and S. E. Harris, Phys. Rev. Lett. 85, 562-565 (2000). 3 J. Q. Liang, M. Katsuragawa, F. Le Kien, and K. Hakuta, Phys. Rev. Lett. 85, 2474-2477 (2000).; M. Katsuragawa, J. Q. Liang, J.Z. Li, M. Suzuki and K. Hakuta, CLEO/QELS '99, Technical Digest, QthE2, pp.195-196, Baltimore, USA, May 23-28 (1999). 4 M. Y. Shverdin, D. R. Walker, D. D. Yavuz, G. Y. Yin, and S. E. Harris, Phys. Rev. Lett. 94, 033904-033907 (2005). 5 M. Katsuragawa, K. Yokoyama, T. Onose, and K. Misawa, Optics Express. Vol. 13, No.15, 5628-5634 (2005); M. Katsuragawa, T. Onose, K. Yokoyama, and K. Misawa, CLEO/QELS 2006, QELS Technical Digest, QFE1, California, USA, May. 21-26 (2006). 6 M. Katsuragawa, T. Onose, T. Suzuki, and K. Misawa, Nonlinear Optics 2007, MB2, Kona, Hawaii, USA, 30 July – 3 August (2007). 7 M. Katsuragawa and T. Onose, Opt. Lett. 30, 2421‐2423 (2005). 8 F. L. Hong, J. Ishikawa, Y. Zhang, R. Guo, A. Onae, and H. Matsumoto, Opt. Comm. 235, 377‐385 (2004).
857
Tunable, octave-spannning supercontinuum driven by X-Waves formation in condensed Kerr media. Alessandro Averchi1,5 , Daniele Faccio1,5 , Antonio Lotti1,5 , Miroslav Kolesik2 , Jerome V. Moloney2 , Arnaud Couairon3 , and Paolo Di Trapani1,4,5 1
CNISM and Department of Physics and Mathematics, University of Insubria, Via Valleggio 11, 22100 Como, Italy E-mail: [email protected] 2 ACMS and Optical Sciences Center, University of Arizona, Tucson, AZ 85721, USA 3 Centre de Physique Th´ ´ eorique, CNRS, Ecole Polytechnique, F-91128, Palaiseau, France 4 Department of Quantum Electronics, Vilnius University, Sauletekio Ave. 9, bldg. 3, LT-10222, Vilnius, Lithuania 5 Virtual Institute for Nonlinear Optics, Centro di Cultura Scientifica Alessandro Volta, Villa Olmo, Via Simone Cantioni 1, 22100 Como, Italy Abstract. We generate an enhanced blue-shifted continuum in bulk Kerr media in ultrashort laser pulse filamentaiton at 1055 nm. At threshold, a spectrally isolated blue peak appears, while at higher energies the continuum expands from the blue peak and spans more than an octave in the spectrum. The central wavelength of the peak can be tuned in a 150 nm range. The effect is explained in terms of X-waves generation.
Supercontinuum (SC) generation, i.e. the generation of an ultra-broadband spectrum starting from a laser pulse is attracting much interest due to the potential applications in a wide range of areas, such as LIDAR, few-cycle pulses generation and ultrafast spectroscopy [1]. In particular, filamentation in Kerr media has been proven to be an efficient way to generate SC: the key mechanism for spectral broadening is self phase modulation (SPM) along with shock fronts formations. Even if the fundamental physical processes behind the SC generation are well known, a number of features have been observed in filamentation whose real nature remains unclear. In recent years it has been demonstrated that the filamentation process is inextricably linked with the reshaping of the input pulse into conical waves (X-waves in the case of normal dispersion) [2]. In this work we study the formation of strongly blue-shifted, spectrally isolated radiation associated to filamentation in bulk fused silica in the normal dispersion regime. Similar observations have been previously reported in anomalous dispersion [3] and with chirped input pulses [4]. With our experiment and simulation we show that this blue spectral peak is the blue-shifted tail of the same X-wave in which the pulse is reshaping into during the filamentation process. Increasing the input pulse energy we obtain a SC whose banwidth spans more than an octave. Notably, unlike usual SC generation, the spectral broadening starts from the blue and not from the pump. We also demonstrate that the spectral position of the blue tail depends on the X-Wave group velocity which can be tuned continuously by controlling the input pulse parameters. Experiments were performed with a 1 ps duration (FWHM) 1055 nm laser pulse delivered by a 10 Hz amplified Nd:glass laser (Twinkle, Light Conversion Ltd., Vilnius, Lithuania). The pulse had a diameter of 5 mm (FWHM) and was focused into a 1 cm long sample of fused silica using a 51 cm focal length lens. The input energy was
858
adjusted using a first order half-wave plate and a polarizer.
Fig. 1. (Left) Experimental spectra with increasing input energy: (a) 20 µJ, (b) 30 µJ, (c) 50 µJ. (Right) Spectra measured for different positions of the sample with respect to the focusing lens. (a) 49 cm, (b) 50 cm, (c) 51 cm, (d) 52.5 cm, (e) 53 cm and (f) 54 cm. We performed a first series of measurements fixing the position of the glass sample at z = 52.5 cm and gradually increasing the energy of the input pulse. Figure 1. (Left) shows the spectra recorded using a 14-bit fiber-coupled spectrometer (Ocean Optics) with increasing input energy. At threshold energy we observe a peak in the spectrum at 450 nm. The peak is clearly spectrally isolated, with more than an octave shift from the pump wavelength, and with an energy roughly around 100 nJ, corresponding to a conversion efficency of 0.5%. When increasing the input energy a SC starts to develop; the most notable feature of the bandwidth increase is that, rather than extending from the pump spectrum, as is commonly observed and as would be expected in a SPMrelated process, the SC grows starting from the blue peak. In a second series of measurements we placed the sample at different positions around the focus of the input lens. Here we kept the energy around 20 µJ, slightly adjusting it in each case so as to be just above the blue peak generation threshold. Figure 1. (Right) from (a) to (f) shows the spectra taken respectively at different positions cm from the input lens. By putting the output facet of the sample close to the focus we observed an isolated peak at 405 nm which may be continuously tuned up to 550 nm when shifting the sample further away. To investigate more in detail the process we measured the angularly resolved spectra at the output of the sample using a commercial imaging spectrometer (lot-Oriel MS260i) and recorded it with a modified digital Nikon D70 camera (Figure 2.). In the figure we note the onset of conical emisison around the pump and, most importantly, the angular dispersion of the blue peak. To proof that X-wave reshaping of the pulse is occurring, p we fit the conical emission at the pump wavlength using the X-wave relation k⊥ = k2 − kz2 with k = (ω/c) and kz (ω) = k(ωlaser ) +
ω − ωlaser vx
(1)
where vx is the group velocity of the X-wave, determined directly from the experimental angular spectrum with a method described in [5] as vx = 2.035 × 108 m/s. As may be seen the curve reproduces the conical emission at the pump wavelength and in the blue part of the spectrum it fits very closely the position and the angular dispersion of the blue peak. It is important to note that the position of the blue peak in the spectrum is determined by the precise value of vx : larger group velocities lead to larger wavelength gaps between the pump and the blue-shifted X tails. 859
Fig. 2. Experimental angularly resolved spectra for input energies of (a) 20 µJ (b) 40 µJ. The solid black line shows a plot of Equation 1. with vx = 2.035 × 108 m/s. This explains the tunability of the blue peak we observed: indeed, it has been shown that the group velocity of the X-waves forming at the beginning of filamentation depends on the peak intensity reached by the pulse during the initial collapse stage [2]. By changing the position of the sample respect to the focus of the lens, we are modifiying the input pulse condition and in turn the collapse dynamic, so that the group velocity of the resulting X-wave is varied. From Equation 1. this determines also the central wavelength of the blue shifted peak. In the experiment we noticed that also keeping the position of the sample fixed but changing the aperture of the beam with an iris allows also the tunability of the spectrum. To further proof the validity of our interpretation we performed a series of numerical simulations using the Unidirectional Pulse Propagation Equation solver [6]. To investigate the influence of the input pulse condition on the group velocity of the X-wave and the position of the blue peak we simply changed the beam diameter before the focusing lens from 5 mm to 3 mm. This produced a corresponding variation in the X-wave forming during the filamentation, as expected, and in turn a shift of 50 nm in the blue peak central wavelength, confirming our understanding (data not shown). In conclusion our measurements show the formation of an isolated blue peak, tunable in wavelength by changing the input pulse conditions and highlight a new mechanism for SC generation. The process can be described as due to the X-waves formation which is occurring during filamentation. 1 R. R. Alfano, The Supercontinuum Light Source, Springer-Verlag, New York, 1989. 2 D. Faccio, M. A. Porras, A. Dubietis, F. Bragheri, A. Couairon, and P. Di Trapani in Physics Review Letters, Vol. 96, 193901, 2006. 3 J. Liu, R. Li, and X. Xu in Physics Review A, Vol. 74, 0143801, 2006. 4 V. Kartazaev, and R. R. Alfano in Optics Letters, Vol. 32, 3293, 2007. 5 D. Faccio, A. Averchi, A. Couairon, M. Kolesik, J. V. Moloney, A. Dubietis, G. Tamosauskas, P. Polesana, A. Piskarskas, and P. Di Trapani in Optics Express, Vol. 15, 13077, 2007. 6 M. Kolesik, J. V. Moloney, and M. Mlejnek in Physics Revew Letters, Vol. 89, 283902, 2002.
860
Toward Ultrafast Optical Waveform Synthesis with a Stabilized Ti:Sapphire Frequency Comb Matthew S. Kirchner1, Tara M. Fortier1, Danielle Braje1, Andy M. Weiner2, Leo Hollberg1, Scott A. Diddams1 1
National Institute of Standards and Technology, Boulder, Colorado 80305, USA E-mail: [email protected] 2 Electrical and Computer Engineering, Purdue University, West Lafayette, Indiana 47907, USA Abstract. We have developed a system for line-by-line control of a stabilized Ti:Sapphire optical frequency comb. We show individually-addressed 20 GHz comb modes around 960 nm and apply simple masks to demonstrate individual mode control.
Introduction Pulse-shaping of ultrafast laser pulses is a valuable tool in the areas of spectroscopy, X-ray generation, quantum control, and communications [1-4]. Recent experiments have shown the ability to perform this pulse shaping in a line-by-line manner by addressing single frequency components of fiber frequency combs or modulated cw lasers[1,5,6]. Such line-by-line shaping enables the generation of waveforms with time durations greater than the repetition period of the source and is a route to arbitrary optical waveform generation [5,6]. We build on these experiments by demonstrating the ability to individually address and manipulate single frequency components of a Ti:Sapphire frequency comb that can be stabilized in both repetition rate and carrier envelope offset frequency. The absolute stabilization of the frequency modes can provide femtosecond timing jitter in the generated waveforms as well as precise control of the carrier phase within the pulse envelope. These features will expand the capabilities of traditional pulse shaping and should enable new applications in secure communication and data transfer. A unique advantage of using this octave-spanning Ti:Sapphire comb is the opportunity to perform this line-by-line manipulation over a broad range of wavelengths from 650 nm to 1050 nm while retaining the low timing jitter provided by locking the comb to a stable optical frequency reference.
Setup We employ a 1 GHz repetition rate octave-spanning Ti:Sapphire frequency comb with spectral coverage from 550-1100 nm, as shown in Fig. 1a [7]. Using dichroic mirrors, we pick off light at 550 and 1100 nm to detect the carrier envelope offset frequency with a SNR of 35 dB in 300 kHz resolution bandwidth. The optical setup allows light at 657 nm to be spectrally separated and compressed to 25 fs for use as a reference pulse in cross-correlations as well as for locking the laser to a calcium optical clock. The rest of the light is double passed through a 20 GHz Fabry-Perot filter cavity (F ~ 300) with mirrors centered at 910 nm [8]. The cavity is designed for low dispersion and has a single-pass suppression of off-resonant modes of 27 dB with an acceptance bandwidth of around 100 nm when centered at the peak of the input laser spectrum (970 nm). The double pass configuration allows for an off-resonant
861
mode suppression of 50 dB while maintaining the high acceptance bandwidth of our mirrors. The output spectrum from the filter cavity is shown in Fig. 1b.
26
Fig. 1. a) Octave-spanning laser spectrum shown on a log scale. Light at 550 and 1100 nm is used to detect and stabilize the carrier envelope offset frequency. Any one of the cw optical frequency references listed can be used to stabilize the absolute frequency of all the comb modes, leaving a broad swath of spectrum from 650-1050 nm available for pulse shaping. In this case we send light from 850-1050 nm to a 20 GHz filter cavity. b) 20 GHz output of the filter cavity showing good coupling over 100 nm. c) Output of pulse shaper showing a hard edge at each end of the aperture of the SLM. Over 600 modes are captured in the aperture of the SLM.
After the comb is filtered to 20 GHz, it is spatially expanded and sent to a pulse shaper, consisting of a 1200 grooves/mm gold grating, a 1 m focal length lens and a 640 pixel liquid crystal spatial light modulator (SLM) in reflection mode that is capable of both amplitude and phase control. The optics are arranged so that the 20 GHz modes nominally match the 100 µm/pixel pitch of the SLM. The modulated light is retroreflected and is picked off by an optical isolator. Its spectrum is shown in Fig. 1c. The reflected light shows a bandwidth of 40 nm centered around 965 nm which should allow for pulses shorter than 50 fs. The total shaper output power is 200 µW. We have amplified the output to 5-10 mW using a semiconductor optical amplifier. In principle, we can manipulate this entire bandwidth; however, in the present configuration the physical separation between modes varies across the aperture of the SLM (due to the non-constant angular dispersion of the grating). We achieve good overlap between optical modes and SLM pixels over a subset of the full aperture before the modes walk off of a SLM pixel. Currently we achieve a bandwidth of about 7.5 nm (120 modes) around 959 nm. Approaches to reduce the pixel walk-off limitations include using a spacing of 2 SLM pixels per comb mode or using a grating plus a prism (grism) to minimize dispersion in comb mode separation and provide 1 pixel per comb mode across the entire aperture of the SLM. Eventually, we hope to have full control over 320 comb modes (in the two pixel per comb mode case) or 640 comb modes (in the grism case).
862
Results To demonstrate the individual addressing of many comb modes, we performed basic amplitude masking on the 120 comb modes from 958 to 965 nm. In this preliminary demonstration, we turned off every other mode to double the repetition rate as shown in Fig. 2a. The extinction was greater than 10 dB, but should be improved by adjusting the mode size at the SLM. We amplified the light from the shaper with a semiconductor optical amplifier and examined the time domain signal with a fast photodiode and oscilloscope as shown in Fig. 2b.
Fig. 2. a) Zoomed in view of all the pixels on (red trace) and every other pixel turned off (black dashed trace). b) Fast oscilloscope trace of the 120 amplified comb teeth with all teeth on showing a 20 GHz repetition rate (red trace) and every other tooth turned off showing a 40 GHz repetition rate (black trace). The photodetector bandwidth is 45 GHz.
Conclusion We have shown individual control over 120 comb modes around 960 nm from a filtered Ti:Sapphire frequency comb. The optical frequency of all comb teeth can be stabilized to better than one part in 1015 by locking the comb to one of the optical references, ultimately enabling novel waveform generation with femtosecond timing jitter.
References 1 2 3 4 5 6 7 8
A.M.Weiner, Rev. Sci. Instr. 71,, 1929–1960 (2000). R. Bartels, S. Backus, E. Zeek, L. Misoguti, G. Vdovin, I.P. Christov, M.M. Murnane, H.C. Kapteyn, Nature 406, ( 164–166 (2000). A.M. Weiner, D.E. Leaird, G.P. Wiederrecht, K.A. Nelson, Science 247, 1317–1319 (1990). N. Dudovich, D Oron, Y. Silberberg, Nature 418, 512–514 (2002). Z. Jiang, D. S. Seo, D. E. Leaird, and A. M. Weiner, Opt. Lett. 30, 1557-1559 (2005). Z Jiang, C. Huang, D.E. Leaird, A.M. Weiner, Nat. Photon. 1, 463-467 (2007). T. M. Fortier, A. Bartels, and S.A. Diddams, Opt. Lett. 31, 1011 (2006). K. Yiannopoulos, K. Vyrsokinos, E. Kehayas, N. Pleros, K. Vlachos, H. Avramopoulos, G. Guekos, IEEE Photon. Tech. Lett. 15,, no. 9, 1294-1296 (2003).
863
Multimillijoule Optically Synchronized and Carrier-Envelope-Phase-Stable Chirped Parametric Amplification at 1.5 µm O.D. Mücke1, D. Sidorov1, P. Dombi1, A. Pugžlys1, S. Ališauskas2, N. Forget3, J. Pocius4, L. Giniūnas4, R. Danielius4, and A. Baltuška1 1
Photonics Institute, Vienna University of Technology, Gusshausstrasse 27-387, A-1040, Vienna, Austria E-mail: [email protected] 2 Laser Research Center, Vilnius University, Saulėtekio av. 10, LT-10223 Vilnius, Lithuania 3 Fastlite, Bâtiment 403, Ecole Polytechnique, 91128 Palaiseau, France 4 Light Conversion Ltd., P/O Box 1485, Saulėtekio av. 10, LT-10223 Vilnius, Lithuania Abstract. Efficient infrared 35-THz-wide parametric amplification at 1.5 µm with the energy of ~10 mJ is obtained in a 4-stage OPCPA using a combination of a 1030-nm 200-fs Yb- and a 1064-nm 60-ps Nd amplifier seeded with a common Yb oscillator.
Optical Parametric Chirped Pulse Amplification (OPCPA) [1] has attracted a lot of attention as a promising route toward intensity scaling of few-cycle laser pulses. Intense phase-stable few-cycle laser pulses have numerous intriguing applications in attosecond science and high-field science including attosecond XUV/soft-X-ray pulse generation by high-harmonic generation (HHG), tomographic imaging of molecular orbitals, and laser-induced electron diffraction. A major challenge for using HHG in studies of time-resolved tomography of molecular dissociative states is the low ionization potential Ip of excited molecular states. The resulting competition between state depletion and HHG prevents generation of broad HHG spectra necessary for tomographic reconstruction. One solution are laser sources with high ponderomotive energy Up∝λ2I at moderate intensity level, i.e., infrared phase-stable few-cycle highpower laser systems. High-Up-sources [2,3] also open the door to experimental investigations of the λ-scaling laws of strong-field physics (Keldysh parameter ∝λ-1, electron energies ∝λ2, HHG cutoff ∝λ2, HHG efficiency ∝λ-5.5, minimum attosecond pulse duration ∝λ-1/2 [4]), and they would benefit laser-induced electron diffraction because of the shorter de Broglie electron wavelength and consequently higher spatial resolution [5]. The main objective of our work is to generate IR pulses with ~40-fs duration that fully satisfy the requirements for external spectral broadening in gas [6]. In addition, with an IR pulse we expect to surpass the energy limitation (4-5 mJ at 0.8 μm) for gas broadening schemes because the critical power of self-focusing also scales as λ2. Using mJ pulses from Ti:sapphire amplifiers at 0.8 μm, coherent X-rays in the keV photon energy range were generated by HHG in helium [7]. A technological problem hindering further scaling of the pulse energy beyond several mJ is gas ionization in the gas-filled hollow-fiber compressors required to achieve few-cycle pulse duration at mJ pulse energies. More fundamentally, ionization in helium saturates when the intensity of a few-cycle pulse at 0.8 μm exceeds ~1 PW/cm2, thus the HHG cutoff and photon flux is limited by ground-state depletion in helium in these experiments. Here, we report on the development of a multi-mJ all-optically synchronized and phase-stable OPCPA at 1.5 μm (see Fig. 1). As opposed to our OPCPA systems developed previously, in this work we modify our approach: (1) with the advent of a 864
mature 200-fs Yb MOPA system it became possible to abandon the Ti:sapphire frontend; (2) we avoid working close to the signal-idler wavelength degeneracy and reduce the quantum defect for the signal wave; (3) we employ (nearly) collinear Type II phase matching that, as opposed to Type I, supports a much narrower bandwidth but is free of parasitic self-diffraction. Following the pioneering work of Miller and coworkers [8], we employ Type II KTP/KTA (1030/1064 nm pump, ~1500 nm signal, ~3500 nm idler) because these crystals are transparent for the mid-IR idler wavelength and exhibit a relatively broad bandwidth around 1500 nm. The repetition rate of the Yb:KGW DPSS MOPA (Pharos, Light Conversion, Ltd.), tunable in the range of 1–100 kHz, was set at 10 kHz as a 500-th harmonic of the flash-lamp pumped Nd:YAG amplifier (Ekspla Ltd.) operating at 20 Hz. In our scheme (Fig. 1), both Yb and Nd RA are simultaneously seeded from a single master oscillator that has a modest FWHM bandwidth of 30 nm. To seed the Nd RA, we pick up the 0thorder diffraction beam behind a transmission grating in the pulse stretcher.
Fig. 1. (a) Scheme of the IR OPCPA setup. MO, master oscillator; RA, regenerative amplifier; PA, double-pass post amplifier; S/C grating-based stretcher/compressor; A, acousto-optic programmable dispersive filter (DAZZLER); WLG, white-light generator in a 4-mm-thick sapphire plate; the CEP-stable idler wave from stage 1 becomes the signal wave in stage 2. Stage 1 (BBO, Type I) is pumped at 515 nm, stage 2 (KTP, Type II) at 1030 nm, stages 3 and 4 (KTP, Type II) at 1064 nm. (b) Spectra of Kerr-lens mode-locked Yb:KGW oscillator (dotted), Yb:KGW RA (solid), SHG of Yb:KGW (grey). The Nd:YAG RA (dashed) contains an intracavity 2-mm-thick etalon to narrow the ps amplifier bandwidth.
The output of the Yb:KGW CPA system is used to pump the first two OPA stages. The frequency-doubled output at 515 nm is used to generate white light continuum in sapphire and as a pump of the 1st stage (collinear Type I BBO). This configuration produces a carrier-envelope phase (CEP) stable idler [9] at 1.5 µm that we further use as seed (signal wave) in the subsequent OPA stages (see Fig. 2(a)). CEP stability of the 2nd stage output was verified by means of f-to-2f interferometry in the wavelength range from 690-830 nm (Fig. 2(b)).
Fig. 2. (a) Spectral properties of the final OPA stages. (b) f-to-2f interferogram reflecting CEP stability measured after the 2nd stage OPA.
865
700 (a) -200 0 200 time delay (fs)
740 720 700 (b) -200 0 200 time delay (fs)
0.4
1.0
1.5
1.0 53.5 fs
0.2 0.5
0.5
0.0 0.0
measured retrieved
1400 1500 wavelength (nm)
1.0 0.5
(d) 0.0
phase (π rad)
720
760
phase (π rad) intensity (a.u.)
740
intensity (a.u.)
760
wavelength (nm)
wavelength (nm)
The output of the 2nd stage is stretched in a grating stretcher to ~40 ps and a tunable higher-order phase correction is introduced by DAZZLER (Fig.1). The temporally stretched seed is amplified in two final OPA stages (3rd and 4th) using a 50-mJ picosecond pump pulse from the Nd:YAG system. The maximum pulse energy at 1.5 µm before the 60% efficient grating pulse compressor is ~10 mJ, as measured through a bandpass filter that blocks off the 3.6-µm idler wave. µJ-level 10-kHzrepetition-rate pulses after the 2nd OPA stage, pulse stretcher, and DAZZLER were compressed to ~50-fs with the grating compressor and measured with SHG FROG, as shown in Fig.3. Work is now in progress to compress and characterize the multi-mJ 20-Hz pulses at the output of the 4th OPA, the bandwidth of which supports virtually the same pulse duration as the 2nd stage. The output of the multi-mJ IR OPCPA system will be broadened in a noble gas, where we expect to reach up to 4 times higher pulse energies in comparison with a filament/hollow-fiber pumped at λ=0.8 µm.
-100
0.0 0 100 time (fs)
Fig. 3. a) SHG FROG characterization of the stretched and recompressed 1.5 µm pulses: (a) Measured and (b) retrieved SHG FROG trace. (c) Measured spectrum (black dashed), retrieved spectral intensity (black solid) and phase (grey dashed). (d) Retrieved temporal intensity (black solid) and phase (grey dashed) profile indicating a FWHM 53.5 fs pulse duration.
In conclusion, we have developed a prototype CEP-stable IR OPCPA for high field applications which draws on a straightforwardly scalable picosecond pump at the Nd/Yb fundamental wavelength and uses a femtosecond Yb front-end. Acknowledgements. This work is supported by the Austrian Science Fund (FWF), grants U33-N16 and F1619-N08. 1 2 3 4 5 6 7 8 9
866
A. Dubietis et al., J. Sel. Top. Quantum Electron. 12, 163 (2006), and references therein. T. Fuji et al., Opt. Lett. 31, 1103 (2006). C. Vozzi et at., Opt. Express 14, 10109 (2006); C. Vozzi et at., Opt. Lett. 32, 2957 (2007). P. Colosimo et al., Nature Phys. 4, 386 (2008); J. Tate et al., Phys. Rev. Lett. 98, 013910 (2007); K. Schiessl et al., Phys. Rev. Lett. 99, 253903 (2007); A. Gordon et al., Opt. Express 13, 2941 (2005); B. Shan et al., Phys. Rev. A 65, 011804(R) (2001). M. Meckel et al., Science 320, 1478 (2008); S. N. Yurchenko et al., Phys. Rev. Lett. 93, 223003 (2004); M. Spanner et al., J. Phys. B 37, L243 (2004). C. P. Hauri et al., Appl. Phys B 79, 673 (2004); C. P. Hauri et al., Opt. Lett. 32, 868 (2006). J. Seres et al., Nature 98, 433 (2005). D. Kraemer et al., Opt. Lett. 31, 981 (2006); D. Kraemer et al., JOSA B 24, 813 (2007). A. Baltuška et al., Phys. Rev. Lett. 88, 133901 (2002).
5-fs multi-mJ CEP-locked parametric chirpedpulse amplifier at 1 kHz S. Adachi1, 3, N. Ishii1, 3, H. Ishii1, 3, T. Kanai1, 3, A. Kosuge1, 3, Y. Kobayashi2, 3, D.Yoshitomi2, 3, K. Torizuka2, 3, and S. Watanabe1, 3 1
Institute for Solid State Physics, University of Tokyo, Kashiwanoha 5-1-5, Kashiwa Chiba 277-8581, Japan 2 National Institute of Advanced Industrial Science and Technology (AIST), 1-1-1 Umezono, Tsukuba 305-8568, Japan 3 CREST, Japan Science and Technology Agency, Sanbancho 5, Chiyoda-ku, Tokyo 102-0075, Japan [email protected] Abstract. We report an optical parametric chirped-pulse amplifier with 5.5-fs pulse duration, 2.7-mJ pulse energy at a 1-kHz repetition rate, pumped by a 450-nm pulse from a frequencydoubled Ti:sapphire laser.
Introduction The concept of optical parametric chirped-pulse amplification [1] (OPCPA) has been recognized to be promising for the generation of high-intensity few-cycle laser pulses. So far, sub-10 fs, multi-mJ to multi-tens-of-mJ OPCPA systems have been reported [2-4]. Recently we reported sub-7-fs, 1.5 mJ OPCPA system pumped with a 400-nm pulse from a frequency-doubled Ti:sapphire laser. However, the shorter pulse width (~ 5 fs) is essentially important to obtain an isolated attosecond pulse by a high-harmonic generation process as discussed in Ref [5].
Gain spectra with several pump wavelengths Even the generation of sub-5-fs pulses has been already demonstrated from noncollinear optical parametric amplifiers (NOPA), but with rather low pulse energies [6-8]. Figure 1 shows the calculated parametric gain bandwidths of a type-I BBO crystal (widely used for OPA) for several pump wavelengths. In the case of 400-nm pump wavelength, the signal spectral range of 520-750 nm [(a)] is the best for broadest amplification [6-8], where the phase mismatch ?k of OPA is widely diminished around the center wavelength. In contrast, the spectral range (a) has poor overlap with the spectrum from a typical broadband Ti:sapphire oscillator [(b)], which is used as a seed source in our OPCPA system. Therefore, in the previous report [9], we were obliged to modify the OPA configuration to match these two wavelength ranges, at the expense of broadest OPA gain bandwidth [(d)]. Meanwhile, there is another solution for this problem: Shifting the pump wavelength itself. From the calculations, the pump wavelength of 450-500 nm [(e), (f)] is desirable to amplify the seed from the Ti:sapphire oscillator [(b)]. Though we could choose any pump wavelength within 450-500 nm since Ti:sapphire has a very broad gain bandwidth (typically 650-1050 nm, corresponding frequency-doubled range of 325-525 nm), we set the pump wavelength at 450 nm (corresponding fundamental wavelength of 900 nm) in order to obtain a moderate Ti:sapphire gain.
867
Figure 1. Calculated parametric amplification ranges of a BBO crystal for several pump wavelengths. Noncollinear and phasematching angles for the calculations were optimized for each pump wavelength.
Experiment The overview of our OPCPA system is described elsewhere [9]. A master oscillator (Venteon OS version, Nanolayers) produced octave-spanning spectrum [10]. The spectral components at 570 and 1140 nm were utilized for the carrier-envelope phase (CEP) stabilization with an f-to-2f setup, while 600-1100 nm spectral component was sent to a stretcher as a seed pulse of OPCPA. We observed a CEP beat note with a SNR of ~ 35 dB in a 100-kHz resolution bandwidth, and the CEP of the oscillator was verified to be stabilized for > 1 hour by a home-built phase-lock loop. All of the optical elements (mirrors, polarizers, etc) in the Ti:sapphire pump laser system were replaced to be optimized at 900 nm. 5-mJ and 20-mJ fundamental radiations at 900 nm were generated from the pump system, and after the frequencydoubling 4-mJ (for pre-OPA) and 15-mJ (for power-OPA) pump pulses at 450 nm were obtained. The stretcher of the OPCPA consisted of a grating pair and a prism pair to give a negative chirp, which was compensated with a bulk material compressor (SF57, 25 cm) after parametric amplification. Its configuration was slightly modified from that in Ref [9] to compensate for broader spectral range (650930 nm). The duration of the stretched seed pulse was calculated to be ~ 50 ps. Figure 2. (left) Measured OPCPA spectra pumped with (a) 400-nm pump pulse and (b) (c) 450nm pump pulse. Figure 3. (right) Spectral and temporal profiles of recompressed OPCPA output pulse. (a) Solid curve, spectral intensity; dotted curve, spectral phase. (b) Solid curve, temporal intensity; dashed curve, transformlimited intensity.
868
Figure 2 displays the OPCPA spectra pumped at 450 nm for slightly different phase-matching angles [(b), (c)], as well as the previous result obtained with 400-nm pump pulse [(a)]. Now the spectral width became substantially broader enough to support a 5-fs pulse generation. Figure 3 shows spectral and temporal profiles of the recompressed OPCPA output pulse obtained with SPIDER measurement. The spectral phase was almost flat over the whole OPCPA spectral range of > 130 THz after the huge compression ratio of 104 (50 ps : 5 fs), enabled by an acousto-optic programmable dispersion filter (Dazzler, Fastlite) with an adaptive phase control. Consequently, the reconstructed pulse width was 5.5 fs, which is almost equal to the transform-limited pulse width [Fig. 3(b)]. The compressed output energy was 2.7 mJ with 15-mJ pump pulse ( ~ 17 % conversion efficiency). The evaluation of the CEP drift induced by the OPCPA was implemented with the second f-to-2f interferometer placed after the compressor. The CEP of the OPCPA output was stable over 30 sec except for a slow drift, which can be easily compensated by providing feedback to the relative delay in the first f-to-2f interferometer.
Conclusions We have demonstrated the OPCPA system with 5.5-fs pulse duration, 2.7-mJ pulse energy at a 1-kHz repetition rate, pumped by a 450-nm pulse from a frequencydoubled Ti:sapphire laser. This CEP-locked TW-class few-cycle laser system will offer an ideal light source for the experiments of ultrafast spectroscopy in the soft Xray in the attosecond timescale. 1 2 3 4 5 6 7 8 9 10
A. Dubietis, G. Jonusauskas, and A. Piskarskas, Opt. Commun. 88(4-6), 437-440 (1992). N. Ishii, et al., Opt. Lett. 30(5), 567-569 (2005). S. Witte, et al., Opt. Express 14(18), 8168-8177 (2006). F. Tavella, A. Marcinkevicius, and F. Krausz, Opt. Express 14(26), 12822-12827 (2006). R. Kienberger, et al., Nature 427(6977), 817-821 (2004). A. Shirakawa, I. Sakane, M. Takasaka, and T. Kobayashi, Appl. Phys. Lett. 74(16), 2268-2270 (1999). M. Zavelani-Rossi, et al., Opt. Lett. 26(15), 1155-1157 (2001). A. Baltuska, T. Fuji, and T. Kobayashi, Opt. Lett. 27(5), 306-308 (2002). S. Adachi, et al., Opt. Lett. 32(17), 2487-2489 (2007). O. Mucke, et al., Opt. Express 13(13), 5163-5169 (2005).
869
Sub-two-cycle pulses at 1.6 μm from an optical parametric amplifier D. Brida1, G. Cirmi1, C. Manzoni1, M. Marangoni1, S. Bonora1,2, P. Villoresi2, S. De Silvestri1, and G. Cerullo1* 1
National Laboratory for Ultrafast and Ultraintense Optical Science – INFM-CNR, Dipartimento di Fisica, Politecnico di Milano, Piazza L. da Vinci 32, 20133 Milano, Italy. 2 LUXOR - Laboratory for UV and X ray Optical Research – CNR-INFM, D.E.I. - Università di Padova, Italy. * E-mail: [email protected] Abstract. We demonstrate two optical parametric amplifier schemes, based on β-bariumborate and periodically-poled lithium tantalate respectively, generating ultrabroadband pulses in the 1-2 μm range. Using a deformable mirror compressor we obtain 8.5-fs pulses at 1.6 μm.
Light pulses with duration of just few optical cycles are important for a number of applications, ranging from time-resolved optical spectroscopy to high field science. Such pulses have been generated by a variety of techniques: directly from a laser oscillator, by spectral broadening in a fiber or by ultrabroadband optical parametric amplifiers (OPAs). OPAs seeded by a white-light continuum (WLC) support under suitable conditions ultrabroad gain bandwidths [1, 2]. In fact the phase matching bandwidth in an OPA depends on the group velocity (GV) mismatch between signal and idler, so that broadband gain requires matching the GVs of signal and idler. This condition occurs either for type I phase matching around the degeneracy point or in a non-collinear OPA (NOPA), in which the signal GV is matched to the projection of the idler GV along the signal direction. The broad gain bandwidths of OPAs have been exploited for few-cycle pulse generation in the visible [1] and around 800 nm [2], but so far the 1-2 μm spectral range has not been explored and the shortest pulses in this region have ≈15 fs duration. Here we report on two different schemes for the generation of ultrabroadband near-IR pulses: a degenerate OPA in β-barium borate (BBO) and a NOPA in Periodically Poled Stoichiometric Lithium Tantalate (PPSLT). Both schemes, when pumped at 800 nm, produce, in a simple single stage setup, pulses with μJ-level energy and spectra spanning the 1-1.7 μm range (for the NOPA) and the 1.2-2.1 μm range (for the degenerate OPA). Using an adaptive pulse shaper employing a Deformable Mirror (DM), we produce nearly transform-limited (TL) pulses with 8.5 fs duration at 1.6 μm, corresponding to less than two optical cycles. These are to our knowledge the shortest light pulses generated in this wavelength range. Fig. 1 shows the experimental setup of the broadband near-IR OPAs. The pump pulses are derived from a Ti:sapphire laser (80 µJ, 150 fs at 800 nm and 1 kHz). A ≈2 μJ fraction of the pump is focused in a 3-mm-thick sapphire plate to generate a WLC seed; the remaining energy pumps the nonlinear crystal. For the degenerate OPA we use a 3-mm-thick type I BBO crystal (θ=21°) in a nearly collinear configuration. The amplified signal has an energy of 2÷3 μJ and a spectrum covering the 1200-2100 nm wavelength range (dashed line in Fig. 2(a)). One can shift the amplified bandwidth to
870
the blue by using non-collinear phase matching in PPSLT [3], for which, in contrast to BBO, the GV of the idler is larger than that of the signal. For the IR NOPA we use a 1.2-mm-thick PPSLT crystal, with poling period Λ=20.9 μm and non-collinear angle α = 2°. We obtain again μJ-level pulses with a spectrum shown as solid line in Fig. 2(a), which is blue-shifted by ≈ 150 nm with respect to the degenerate OPA one.
Fig. 1. Experimental setup of the broadband near-IR OPAs based on BBO and PPSLT. It is possible to switch from one configuration to the other by just replacing the nonlinear crystal and varying the angle between pump and signal beams. VA, variable attenuator; BS, beam splitter, HW half-wave plate.
Both OPA schemes generate ultrabroadband near-IR pulses supporting sub-10-fs duration, corresponding to about two optical cycles in this wavelength range. Their spectral phase shows a strong high-order dispersion contribution, which is impossible to correct by using prism or grating pairs. To achieve dispersion compensation to all orders, we implemented an adaptive system based on a DM. The DM is placed in the Fourier plane of a 4f zero-dispersion pulse shaper [4] consisting of a Brewster-cut SF56 prism and a spherical gold mirror. The DM is a rectangular silver-coated membrane activated by 30 linear electrodes. We chose SF56 as prism material because it shows good angular wavelength dispersion, enabling to fill the DM nearly completely, while adding a low contribution to the pulse spectral phase. Furthermore, the use of a prism instead of a grating reduces the losses. The mirror response was calibrated in order to define the influence function matrix describing the effect of each electrode on the membrane shape. The mirror shape for optimum pulse compression was obtained by first measuring the pulse spectral phase with Second Harmonic Generation Frequency Resolved Optical Gating (SHG-FROG) and then introducing the additional phase required to compensate it. Such process was iteratively repeated acting on the residual phase after the previous correction. In such approach it is crucial to accurately map each wavelength on the mirror by measuring with a spectrometer the wavelengths corresponding to 5 different positions and interpolating the results with the angular dispersion function of the prism. This procedure also enables to reproduce the compression results on a day to day basis. Fig. 2(b) shows the SHG-FROG trace, measured with a 10-μm-thick BBO crystal, of the compressed pulses from the degenerate OPA, while Fig. 2(c) reports
871
the retrieved temporal intensity and phase profile. The measured 8.5 fs pulsewidth is very close to the TL duration (8.2 fs) and corresponds to less than two cycles of the 1.6 μm carrier frequency (5.3 fs period). Similar results are expected for the PPSLTbased NOPA.
Fig. 2. (a) Spectra of the degenerate OPA in BBO (dashed line) and the NOPA in PPSLT (solid line). (b) SHG-FROG field trace (128×128 pixels) for the degenerate OPA pulse compressed by the DM system. (c) Retrieved temporal intensity and phase profile of the compressed pulse (reconstruction error = 0.0107).
In conclusion, we have demonstrated two OPA configurations, based on BBO and PPSLT respectively, capable of generating ultrabroadband pulses in the 1-2 μm wavelength range, and achieved nearly TL 8.5 fs at 1.6 μm using a DM compressor [5]. The μJ-level pulses produced by our simple single-stage system are already suitable for time-resolved spectroscopy in the near-IR with unprecedented resolution. It should be straightforward, by adding one or two similar OPA stages, to scale the output energy by 2÷3 orders of magnitude, enabling the application of the energetic sub-two-cycle pulses to high harmonic generation. For such applications, it will also be possible to stabilize the Carrier-Envelope Phase (CEP) of the pulses by seeding the OPA with a CEP stable broadband pulse produced either by intrapulse difference frequency generation [6] or by WLC generation from the idler of an OPA in which pump and signal are derived from the same source [7]. References 1 2 3 4 5 6 7
872
A. Baltuška, T. Fuji, and T. Kobayashi, Opt. Lett. 27, 306-308 (2002). S. Witte, R. T. Zinkstok, A. L. Wolf, W. Hogervorst, W. Ubachs, and K. S. E. Eikema, Opt. Express 14, 8168-8177 (2006). G. Cirmi, D. Brida, C. Manzoni, M. Marangoni, S. De Silvestri, and G. Cerullo, Opt. Lett. 32, 2396-2398 (2007). E. Zeek, K. Maginnis, S. Backus, U. Russek, M. Murnane, G. Mourou, H. Kapteyn, and G. Vdovin, Opt. Lett. 24, 493-495 (1999). D. Brida, G. Cirmi, C. Manzoni, S. Bonora, P. Villoresi, S. De Silvestri, and G. Cerullo, Opt. Lett. 33, 741-743 (2008). C. Vozzi, G. Cirmi, C. Manzoni, E. Benedetti, F. Calegari, G. Sansone, S. Stagira, O. Svelto, S. De Silvestri, M. Nisoli, and G. Cerullo, Opt. Express 14, 10109-10116 (2006). C. Manzoni, D. Polli, G. Cirmi, D. Brida, S. De Silvestri, and G. Cerullo, Appl. Phys. Lett. 90, 171111 (2007).
Carrier envelope offset control of broad Raman sidebands by locking two pump laser frequencies to a single optical cavity T. Suzuki1,2 , M. Hirai1 , R. Tanaka1 , and M. Katsuragawa1,2 1 2
Department of Applied Physics and Chemistry, University of Electro-Communications, 1-5-1 Chofugaoka, Chofu, Tokyo 182-8585, Japan PRESTO, JST, 4-1-8 Honcho Kawaguchi, Saitama, Japan E-mail: [email protected]
Abstract. We generate broad Raman sidebands with zero carrier-envelope-offset by frequencylocking the pump lasers to a single optical cavity. It is shown in both spectral and temporal domains that the carrier-envelope-offset is controlled to discrete values.
Introduction In the last decade, there has been significant progress in generating high coherence in molecular vibrational and/or rotational transitions by means of adiabatic Raman excitation [1,2]. A molecular ensemble with such high coherence, in turn, strongly modulates the pump laser fields and generates high order Raman sidebands collinear to the pump beams. Since the generated sidebands are mutually coherent, ultrashort pulses can be constructed by synthesizing them [3]. The main feature of this ultrashort pulse generation scheme is that the ultrashort pulses are produced from two single-frequency lasers. Here, we demonstrate that the carrier envelope offset (CEO) of such Raman sidebands can be controlled to discrete values by locking the two pump laser frequencies to a single optical cavity, leading to the generation of ultrashort pulses with a constant carrier envelope phase (CEP). L-N2 cryostat
f = 850 mm
Prism Screen f = 250 mm Dual-wavelength injection locked pulsed Ti:Sa laser
Spectrum analyzer or biplanar photo tube
Pump (Nd:YAG, SH)
BBO crystal
-1,
Dual waveplate
0
2
-1
Ti:Sa
Isolator
Seed laser ( 0) Seed laser ( -1) or ct te De
Frequency locking
on cti ity fle cav e R m fro
High finesse cavity
Frequency locking to a cold cavity
Fig. 1. Schematic diagram of the experimental setup
873
Experimental Methods The Raman sidebands with zero CEO can be generated when the two pump laser frequencies (Ω−1 , Ω0 ) are set to Ω0 = nΔΩ and Ω−1 = (n − 1)ΔΩ where ΔΩ is the frequency difference between the two pump lasers and n is an integer. We realize this critical condition by locking the two pump laser frequencies to a single optical cavity and then choosing the appropriate longitudinal modes. The schematic of the experimental setup is illustrated in Fig. 1. The CEO control with the single optical cavity was examined for the following two cases. One is the case using the cavity of the pump laser itself and the other is the case using an external highfinesse cold cavity. The two-frequency pump pulses to drive the high Raman coherence are generated from the dual-wavelength injection-locked Ti:Sa laser, which consists of a single cavity. The two pump-laser frequencies, Ω−1 , Ω0 , are thereby necessarily restricted to integer multiples of the free-spectral-range (FSR) of the laser cavity. We set the two frequencies, Ω−1 , Ω0 , around the zero CEO condition by choosing the appropriate longitudinal cavity modes. The Ti:Sa laser cavity has small dispersion due to the Ti:Sa crystal and thus the FSR is different for each frequency components. Therefore the ratio of two longitudinal cavity mode numbers which gives the smallest CEO is not equal to the simple integer ratio of n − 1 to n. However, the CEO can be controlled by steps of the FSR, because the deviation of the FSR is small compared with the FSR itself. By using a cold cavity to lock the frequencies instead of the Ti:Sa cavity, the small internal dispersion of the cavity can be removed and zero CEO is realized with the ideal mode ratio of n − 1 to n. Once we lock two frequencies to the ideal modes of the cavity, the CEO remains zero, in principle, irrespective any drift in cavity length. Moreover, because our cold cavity has a finesse of ∼200 which is one order higher than the Ti:Sa cavity, it is expected that the accuracy of the frequency locking will be improved.
Fig. 2. Photograph of broad Raman sidebands generated by two pump pulses Ω0 and Ω−1 , and the second harmonic 2Ω−1 .
In order to evaluate the CEO, we further introduced the second harmonic of one of the pump lasers, 2Ω−1 , in addition to the pump lasers, Ω−1 , Ω0 . The Raman sidebands are generated from the pump pulses, Ω−1 , Ω0 , and simultaneously from the second harmonic, 2Ω−1 , giving the CEO information through an overlap of both the sidebands, which is well-known as the f − 2 f technique in a femtosecond laser comb. It should be noted that this scheme also provides us with an octave-spanning Raman sideband.
Results and Discussion Figure 2 shows a photograph of the generated Raman sidebands dispersed with a prism. It is clearly seen that the sidebands originating from the fundamental pump lasers, Ω−1 , 874
Ω0 , and those from the second harmonic, 2Ω−1 , are overlapped with each other. We picked up the sidebands at 513 nm, and introduced them into an optical spectrum analyzer. Figure 3a shows CEOs obtained from the frequency deviation between the pair of sidebands versus the selected longitudinal modes of the pump lasers. It is clearly shown that the CEO was controlled with steps of the FSR (1.24 GHz) and could be set to zero by choosing the appropriate longitudinal modes. We also carried out the same CEO measurement, but in the temporal domain, by employing a fast detection system with a response time better than 10 GHz. In the temporal domain, the CEO is observed as a beat between a pair of sidebands. Figure 3b shows the results obtained with the same pair of sidebands as in Fig. 3a. It is confirmed that the CEO is controlled consistently to that found in the frequency domain. (b) 4.0
2.0
2.0
Frequency difference (GHz)
Frequency difference (GHz)
(a) 4.0
0.0
-2.0 FSR (1.24 GHz)
-4.0
0.0
-2.0 FSR (1.24 GHz)
-4.0
1 mode
1 mode -6.0
-6.0 -4
-3
-2 -1 0 1 Relative mode number
2
3
-4
-3
-2 -1 0 1 Relative mode number
2
3
Fig. 3. CEOs measured in (a) spectral and (b) temporal domains. Ti:Sa pump laser cavity has a FSR of 1.24 GHz. The CEOs can be controlled with the discrete FSR steps.
We also examined the case of the cold cavity frequency locking. As shown in fig. 1, a high-finesse cold cavity was used for frequency locking. In the same way as before, we obtained in the temporal domain that the CEOs were controlled with steps of the FSR of the cold cavity (1.94 GHz). The difference of the FSR for the two frequency components is three orders less than that of the Ti:Sa cavity, and the controllability of the CEO is estimated to be as precise as 108 Hz. It means that the CEP is maintained at least for a few nanoseconds, which is enough to cover a train of ultrashort pulses over nanosecond pulsed envelope. 1 S. H. Harris and A. V. Sokolov, Phys. Rev. A55, R4019, 1997. 2 J. Q. Liang, M. Katsuragawa, F. Le Kien, and K. Hakuta, Phys. Rev. Lett. 85, 2474, 2000. 3 M. Katsuragawa, K. Yokoyama, T. Onose, and K. Misawa, Opt. Express 13, 5628, 2005.
875
Cancellation of the coherent accumulation in rubidium atoms excited by a train of femtosecond pulses T. Ban, D. Aumiler, H. Skenderovic, S. Vdovic, N. Vujicic and G. Pichler Institute of Physics, Bijenicka 46, HR-10 000 Zagreb, Croatia E-mail: ticijana@ifs, Abstract. In present experiments gradual change from the frequency comb excitation to pulse by pulse excitation of Rb atoms is observed. Shown results could lead to the development of a new method for system coherence monitoring.
Introduction Our recent work in single photon spectroscopy of rubidium [1,2] combines fixed comb lines, broad absorption (rubidium atoms at room temperature) and an additional cw scanning probe laser. Resonant excitation of the rubidium atoms by discrete frequency comb optical spectrum results in the comb-like velocity distribution of the excited state hyperfine level populations and velocity-selective population transfer between the Rb ground state hyperfine levels. A modified DFCS was developed which uses the fixed frequency comb for the Rb 52 S1/2 → 52 P1/2,3/2 excitation and the weak cw scanning probe laser for ground levels population monitoring. Observed modulations in the probe absorption are a direct consequence of the velocity-selective optical pumping (VSOP) induced by the frequency comb excitation. The fs pulse train excitation of Rb four 52 S1/2 → 52 P1/2 and six 52 S1/2 → 52 P3/2 level Doppler-broadened system was investigated theoretically in the context of the density matrix formalism. The analogous effect has also been observed at theoretically treated in the case of cesium atoms [3]. This work [4] improves the sensitivity of the detection by introducing the lock-in technique. This technique eliminates the Doppler background from the signal and results in direct monitoring of the modulation of the probe laser transmission. The structure and depth of the observed modulations are unique for each of the four Doppler broadened absorption lines and they reflect the structure of hyperfine levels and values of corresponding transition dipole moments. The enhanced sensitivity (over ten times) enables us to study various effects in frequency comb spectroscopy of Doppler broadened lines more thoroughly. In this work, we investigate how the frequency comb induced VSOP is influenced by the cw probe laser intensity. It turns out that in the strong probe regime the frequency comb excitation leads to the increase of the probe absorption and to the disappearance of the modulation structure. The disappearance of the modulation structure implies the disappearance of the VSOP. This imposes the conclusion that in the strong probe case, the atomic excitation is not driven by pulse train, but rather by pulse by pulse excitation. The physical mechanism behind this effect is the effective shortening of the atomic coherence relaxation time due to the strong probe laser. The numerical calculations of the density matrix evolution for this system, already developed in [1,2], are amended by introducing the probe laser electric field. Calculated probe laser absorption supports the experimental findings.
876
Experimental Methods A Tsunami (Spectra Physics) mode-locked Ti:sapphire laser with pulse duration of about 100 fs generates 5.8 THz broad optical frequency comb. The comb frequencies were not tunable and the frequency comb was kept constant during the measurements. The laser repetition rate fr, measured with a fast photodiode amounts 80 MHz. The fs laser beam, chopped with a SR540 mechanical chopper at 3.5 kHz repetition rate, was weakly focused onto the center of the glass cell containing rubidium vapor at room temperature. The 85,87 Rb 52 S1/2 hyperfine ground state populations were monitored with a cw diode laser (Toptica DL100, ECDL at 780 nm), which propagated anti-collinearly with the fs laser, intersecting it under a small angle in the center of the cell. The probe frequency was slowly scanned across the Doppler-broadened 85,87 Rb 52 S1/2 → 52 P3/2 hyperfine transitions at 3 GHz/s scanning rate. The probe laser transmission was simultaneously detected with two Hamamatsu Si photodiodes and fed into a digital oscilloscope (Tektronix TDS5140).
Results and Discussion Transmission across all four Doppler broadened 52 S1/2 → 52 P3/2 absorption lines at 780 nm for the fs laser tuned to 52 S1/2 → 52 P1/2 transition at 795 nm is shown in Fig.1. Two outer absorption lines result from the 87 Rb absorption, whereas the inner two come from 85 Rb absorption. The excited state 52 P3/2 hyperfine levels are not resolved due to the Doppler broadening. The modulations of the absorption profiles are the result of the fs pulse train excitation of Rb atoms. The signal from the second photodiode PD2, Fig. 1, was fed to the lock-in amplifier referenced to the mechanical chopper on the fs laser beam. The lock-in signal represents the change of the probe laser transmission induced by fs laser. The lock-in output is monitored on the digital oscilloscope. The advantage of the lock-in detection technique is obvious, since the broad Doppler background is eliminated. Additionally, the signal to noise ratio is greatly enhanced providing insight to finer details of the modulations. We investigated the change in the probe transmission due to the fs laser excitation, for different probe laser powers. By increasing the probe laser power the structure and depth of the modulations change significantly. First, the fs laser induced increase of the absorption for all velocity groups is observed in the strong probe case. Second, as the probe laser power is increased the modulation depth decreases and a broad background with negative change in trasmission for all velocity groups appears. The velocity selection due to the frequency comb excitation is therefore lost, and the atoms interact with individual pulses rather than with the pulse train. This result is supported by the theoretical calculation where in the strong probe case no accumulation of population and coherence is obtained.
877
Fig. 1. Measured probe transmission for the Rb vapor at room temperature in the case when the fs laser is tuned to 52 S1/2 → 52 P1/2 transition at 795 nm. The probe laser frequency is continuously scanned across all four Doppler broadened 52 S1/2 → 52 P3/2 absorption lines at 780 nm.
Conclusions We have presented enhanced sensitivity measurements of the velocity selective optical pumping (VSOP) of the Rb ground state hyperfine levels induced by the pulse train excitation. Using this approach we were able to directly measure the change of the probe laser transmission as a result of the resonant fs laser excitation of the Rb atoms. The most interesting result has been obtained in the case of the strong probe laser field. In this case, the velocity selection due to the frequency comb excitation is lost, corresponding to the interaction of atoms with the individual pulses rather than with the pulse train. Exploitation of quantum memory for information storage and data processing is one of the fundamental aspects in the experimental physics. In the experiments where the coherence of the system has to be controlled and maintained for as long a time as possible despite higher field intensities, our method could be applied for system coherence monitoring. 1 2 3
D. Aumiler, T. Ban, H. Skenderovic and G. Pichler, Phys. Rev. Lett. 95, 233001 (2005). T. Ban, D. Aumiler, H. Skenderovic and G. Pichler,Phys. Rev. A 73, 043407 (2006). N. Vujicic, S. Vdovic, D. Aumiler, T. Ban, H. Skenderovic and G. Pichler, Eur. Phys. J. D 41, 447 (2007). 4 T. Ban, D. Aumiler, H. Skenderovic, S. Vdovic, N. Vujicic and G. Pichler, Phys. Rev. A 76, 043410 (2007).
878
Part XI
Optics, Optoelectronics, Measurement, Diagnostics and Instrumentation
Sub-10-fs XUV Tunable Pulses at the Output of a Time-Delay-Compensated Monochromator L. Poletto1, P. Villoresi1, E. Benedetti2, F. Ferrari 2, S. Stagira 2, G. Sansone2, M. Nisoli2 1 2
CNR-INFM - D.E.I. - Università di Padova, Padova, Italy CNR-INFM - Dipartimento di Fisica, Politecnico di Milano, Piazza L. da Vinci 32, 20133 Milano, Italy E-mail: [email protected]
Abstract. Extreme-ultraviolet pulses, produced by high-order-harmonic generation, have been spectrally selected by a time-delay-compensated monochromator. Temporal characterization has been obtained using cross-correlation method: pulses as short as 8 fs, with high photon flux, have been measured.
Introduction Extreme ultraviolet (XUV) radiation produced by high-order harmonic generation (HHG) is attracting a large and rising interest for the ample range of possible applications. High brightness, femtosecond XUV pulses are important in various research fields, ranging from time-resolved spectroscopy to holography, microscopy and free-electron laser injection [1]. Various techniques have been implemented to select spectral portions of the XUV spectrum such as metallic filters and XUV dielectric multilayers; however, both of them are limited to a fixed spectral range and do not present any tunability. On the other hand, the use of a single diffraction grating causes severe broadening of pulse duration and significant attenuation. In this work we have measured the duration of the harmonic pulses and the corresponding photon flux after spectral selection by a time-delay-compensated monochromator (TCM) based on the off-plane diffraction mount, where the incident and diffracted wave vectors are almost parallel to the grating grooves. Such off-plane mount allows one to achieve a remarkable improvement in terms of tunability and throughput with respect to the classical diffraction mount [2]. We demonstrate that the developed monochromator allows one to preserve the XUV pulse duration after spectral filtering. Indeed, we have generated broadly tunable coherent XUV pulses, with duration as short as 8 fs and high photon fluxes at the output of the TCM [3]. Such source lends itself as an important tool for a number of applications of femtosecond XUV pulses ranging from atomic and molecular spectroscopy to solidstate physics.
Experimental Results The TCM is characterized by two equal sections, with two toroidal mirrors and a plane grating, as shown in Fig. 1. Since the grating has to be operated in parallel light, the first mirror of each section acts as the collimator and the second mirror as the focusing element. The first section generates a spectrally dispersed image of the harmonic source on the intermediate plane, where a slit is used for spectral selection of the harmonics. The selected spectral portion, composed by a single harmonic or a set of few harmonics, propagates toward the second section, which compensates for
881
both the temporal broadening and the spectral dispersion and generates a stigmatic image on its focal plane. The four mirrors are operated at equal grazing angle and unity magnification to minimize the aberrations. Wavelength scanning is achieved by rotating the gratings around an axis tangent to their vertices and parallel to the grooves. Toroidal mirrors
Intermediate slit
Toroidal mirrors HHG source
Ti:Sa laser
Grating 2
Grating 1
Fig. 1. Optical setup of the time-delay-compensated monochromator.
In this experiment, the laser pulses (25-fs duration, 800-nm central wavelength and 1-kHz repetition rate) are split in two parts using a drilled mirror. The inner part is focused on an Argon cell with static pressure for HHG. The XUV radiation propagates inside the TCM for spectral selection of single harmonics by the intermediate slit. An Argon jet is located at the output focal point of the monochromator. The outer annular part of the infrared (IR) beam is focused onto the same Argon jet, for the cross-correlation measurement with collinear geometry. The delay between the two pulses is controlled by a piezoelectric translator. The photoelectrons generated by single-photon absorption of the XUV pulses are collected by a time-of-flight spectrometer. The duration of the spectrally selected harmonic pulses is obtained by measuring the cross-correlation between the XUV and the 25-fs IR pulses. The harmonic XUV pulse ionizes a gas (Argon) in the presence of the IR field. When the two pulses overlap in time and space on the gas jet, sidebands appear in the photoelectron spectrum, spectrally shifted by the IR photon energy, determined by the absorption of one harmonic photon plus the absorption or the emission of one IR photon. The sideband amplitude as a function of the delay, between the XUV and IR pulses provides the cross-correlation signal [4]. Figures 2(a) and 2(b) show (dots) the temporal evolution of the amplitude of the first sideband vs delay in the case of XUV pulses obtained by selecting the 19th and 23rd harmonic, respectively. In order to obtain the XUV pulse durations, we have first calculated the evolution of the sideband amplitude vs delay assuming the measured IR intensity and duration, for different values of the XUV pulse duration. As shown in Fig. 2(a)-(b), the measured cross-correlation traces can be well fitted assuming an XUV pulse duration t =13±0.5 fs (FWHM) in the case of the 19th harmonic and t =8±1 fs (FWHM) in the case of the 23rd harmonic. The relative durations of the XUV and generating pulses turn out to be in good agreement with numerical simulations based on the nonadiabatic saddle-point method [5]. The experimental results clearly demonstrate that spectral selection of the XUV pulses obtained by using the TCM has been achieved preserving the harmonic pulse duration. We have then characterized the output of the TCM in terms of total photon yield, by placing an absolutely calibrated XUV photodiode at the TCM output. The TCM throughput
882
efficiency is higher than 10% in the 20-40 nm spectral region with a peak of 18% around 30 nm. Using a 230-mJ pump pulse, we have measured a photon flux at the output of 6.5´10 8 ph/s and 1.3´10 9 ph/s, in the case of the 19th and 23rd harmonic, respectively.
Intensity (arb. units)
1.0
(a)
1.0
H19 t = 13 fs
0.8 0.6
0.6
0.4
0.4
0.2
0.2
0.0 -60
-40
-20
0
20
Delay (fs)
40
(b)
H23 t = 8 fs
0.8
60
0.0 -60
-40
-20
0
20
Delay (fs)
40
60
Fig. 2. Amplitude of the first sideband in the case of the 19th (a) and 23rd (b) harmonics as a function of the delay between the XUV and IR pulses. The dots are the experimental results; the solid lines are calculated as explained in the text.
Conclusions In conclusion, coherent XUV pulses produced by HHG have been spectrally selected by a time-delay-compensated monochromator. Pulses as short as 8 fs have been measured at the output of the monochromator with high photon flux. Such high flux, combined with the very short duration, enables a number of novel and intriguing applications of femtosecond XUV pulses. 1 2 3 4 5
A. L'Huillier, D. Descamps, A. Johansson, J. Norin, J. Mauritsson, and C. –G. Wahlström, Eur. Phys. J. D 26, 91, 2003. L. Poletto, Appl. Phys. B 78, 1013, 2004. L. Poletto, P. Villoresi, E. Benedetti, F. Ferrari, S. Stagira, G. Sansone, M. Nisoli, Opt. Lett. 32, 2897, 2007. T. E. Glover, R. W. Schoenlein, A. H. Chin, and C. V. Shank, Phys. Rev. Lett. 76, 2468, 1996. G. Sansone, C. Vozzi, S. Stagira, and M. Nisoli, Phys. Rev. A 70, 013411, 2004.
883
First Step Towards a Femtosecond VUV Microscope: Zone Plate Optics as Monochromator for High-Order Harmonics. J´erˆome Gaudin1,2 , Stefan Rehbein1 , Peter Guttmann1 , Sophie God´e1 , Gerd Schneider1 , Philippe Wernet1 and Wolfgang Eberhardt1 1
Berliner Elektronenspeicherring Gesellschaft f¨ur Synchrotronstrahlung BESSY, Albert-EinsteinStrasse 15, D-12489 Berlin , Germany E-mail: [email protected] 2 European X-Ray Free Electron Laser XFEL - Deutsches Elektronen-Synchrotron, Notkestrasse 85, D-22607 Hamburg, Germany E-mail: [email protected] Abstract. We report the use of zone plate optics as a monochromator for the spectral selection of a single high-order harmonic of a femtosecond laser generated in a rare gas medium in the photon energy range from 30 up to 70 eV while keeping the pulse duration in the femtosecond range. This is our first step towards a VUV microscope with sub-micrometer spatial resolution and femtosecond time resolution.
Introduction Nowadays microscopy using soft X-Rays has became a routine method for high resolution imaging. By combining this technique with a light source delivering ultrashort pulses one would be able to perform time resolved imaging experiments with both highspatial and high temporal resolution. Such light sources in the VUV domain are now available: High-order Harmonics (HHs) of femtosecond lasers. These type of sources can be optimized to routinely deliver high photon flux and ultrashort pulses which make them suitable for imaging experiments and first results have been recently published [1,2]. But so far the use of zone plates (ZPs) with HH-based sources has been limited to objective lenses with HHs of around 100 eV. The main limitation is that the transmission of the standard material for the ZP substrate, silicon nitride, is very low for energies below about 90 eV. To overcome this transmission problem we used a Si foil as substrate. We show that ZP optics provide a sufficient monochromaticity to select only a particular harmonic and that the transmission properties enable to record CCD images in a comparably short acquisition time for photon-energies of a few ten eV.
Experimental Set-up The planed microscope is shown in Fig.1 . The results presented in this article concern only the condenser ZP (parts numbered in red in the Fig.1, see also [3] for a more detailed description). HH generation is driven by an infra-red laser (1.5 mJ, 40 fs per pulse at 1kHz, photon energy = 1.57 eV) which is focused inside a glass capillary (2 cm long, 300 µm inner diameter). The IR beam is then blocked by two 150 nm thick aluminum filters (not shown in Fig.1). At a distance of 1455 mm from the capillary, the condenser ZP is mounted on a two-axis motorized stage. The ZP substrate is a 180 nm thin Si foil. The ZP pattern is made of a polymer with a thickness of 120 nm.
884
This polymer has been chosen as it is known to be stable under VUV irradiation. The total transmission can to be estimated in the order of 5% to 10% in the energy range considered. The ZP consists of N = 2500 zones with a diameter D = 2 mm diameter. The maximum temporal broadening is given by t = Nλ /2c (for the harmonic 33 (51.8 eV) t≈100 fs). A pinhole of d=10 µm diameter is used as an order sorting aperture (OSA). The OSA is mounted on a 3-axis linear stage allowing for a variation of the distance to the ZP with an accuracy of 1 µm. All the measurements are performed with a back illuminated soft X-Ray CCD camera.
6 1
3
4
5
2 1 Fig. 1. The full field VUV microscope, only elements numbered in red have been implemented yet. Top left insert: Image of the 4 main harmonics in Argon. Numbers in the corners indicate the corresponding harmonic order.1: focusing lens - 2: glass capillary for HH generation - 3: condenser ZP - 4: order sorting aperture -5: micro-zoneplate - 6: X-Ray CCD
Results In order to select one harmonic we take advantage of the fact that ZPs are chromatic optics. Hence the different harmonics are focused at different points. Since f increases with decreasing wavelength (i.e. increasing harmonic order) the diameter of the corresponding circle on the CCD is smaller the higher the harmonic order. Adding an OSA allows us to block the -1 order. Moreover, if it is placed exactly at the focus of one harmonic only this harmonic is transmitted. This provides the monochromatization and we obtain CCD images showing only one circle. Images taken for the 4 most intense harmonics as generated in Argon are presented in Fig.1. If one now varies the distance between the ZP and the pinhole, and records an image for each step one obtains a full emission spectrum. Spectra measured for the generation gases Argon and Neon are shown in Fig.2. This allowed us to test the ZP in the photon energy range from 30 to 70 eV. Using the zone plate formula f(λ )=D· drn /m·λ where m stands for the diffraction order, together with the thin lens formula we convert the ZP/OSA distance to photon energy. The graph also shows the centre energies as calculated according to the fundamental laser energy of 1.57 eV. The calculated values are in
885
a good agreement with experimental peak maxima for low harmonics. The mismatch for the highest harmonics can be explained by the blue shift of the fundamental laser while propagating in an ionized medium [4]. This effect is more important in Neon as we focused the IR beam more tightly by using a shorter focal lens, enhancing HH generation efficiency but also enhancing ionization. The typical energy resolution amounts to E/∆E = 42 (FWHM). Finally, the short acquisition time of 10 s (60 s) for Argon (Neon) should be noted. h a r m o n ic o r d e r
h a r m o n ic o r d e r
8
1 0
7
1 0
6
1 0
5
2 7
2 9
3 1
3 3
3 5
3 3
3 7
s ig n a l
1 0
2 5
1 0
7
1 0
6
3 5
3 9
3 7
4 1
4 5
4 3
In te g ra te d C C D
In te g ra te d C C D
s ig n a l
2 3
3 5
4 0
4 5
5 0
5 5
6 0
5 0
5 5
h v (e V )
6 0
6 5
7 0
h v (e V )
Fig. 2.HH spectrum obtained in Argon (left) and in Neon (right)
Conclusions In conclusion, we have demonstrated that ZPs can be used as monochromating optics for HH in the VUV range. Moreover this first successful step opens the way to the development of a full field VUV microscope. In fact by adding a zone plate (element number 5 in Fig. 1) to the existing experiment we will be able to set-up a full field microscope. This so called micro ZP will be the microscope objective. It will image the focal spot of the condenser ZP. The temporal broadening induced will be negligible. The spatial resolution expected is in the order of drn of the condenser ZP. Such a tabletop microscope should allow to perform time resolved imaging experiments with a submicrometer/sub-picosecond resolution. 1
I.R. Fr¨uke, J. Kutzner, T. Witting, H. Zacharias and Th. Wilhein, Eur. Phys. Lett. 72, 915, 2005 2 R.L. Sandberg, A. Paul, D.A. Raymondson, S. H¨arich, D.M. Gaudiosi, J. Holtsnider, R. I. Tobey, O. Cohen, M.M. Murnane and H.C. Kapteyn, Phys. Rev. Lett. 99, 098103, 2007 3 J. Gaudin, S. Rehbein, P. Guttmann, S. God´e, G. Schneider, Ph. Wernet and W. Eberhardt, J. Appl. Phys. 104, 033112, 2008. 4 C.G. Wahlstr¨om, J. Larsson, A. Persson, T. Starczemski, S. Svanberg, P .Sali`eres, Ph. Balcou and A. L’Huillier, Phys. Rev. A 48, 4709, 1993.
886
Measurement of Electron Pulse Duration by Attosecond Streaking P. Reckenthaeler1,2*, M. Centurion1, V.S. Yakovlev2, M. Lezius1, F. Krausz1,2 and E.E. Fill1 1
Max-Planck-Institut für Quantenoptik, Hans-Kopfermann-Strasse 1, D-85748 Garching, Germany 2 Department für Physik der Ludwig-Maximilians-Universität München, Am Coulombwall 1, D-85748 Garching, Germany * E-mail: [email protected] Abstract: We propose a new method to measure the duration of ultrashort electron pulses using the principle of laser assisted Auger-decay.
Introduction The quest for ever shorter electron pulses is motivated by a number of exciting applications, including ultrafast electron diffraction, crystallography and microscopy [1-3], electron imaging [4, 5], the generation of ultrashort X-ray pulses [6, 7] and pumping of X-ray lasers [8]. A key problem in most applications is to measure the duration of ultrashort electron pulses. Several methods have been investigated and applied for this purpose including streak cameras [9], interferometry of coherent transition radiation [9, 10], radio-frequency zero-phasing [11], terahertz radiation diagnostics [12], electro-optic encoding [13] and ponderomotive interaction of a laser with the electron pulse [14]. None of these methods, however, has proved to be applicable in a wide range of parameters: Streak cameras and radio-frequency zero-phasing cannot be used for pulses shorter then a few hundred fs, the coherent transition radiation method as well as the terahertz radiation and electro-optic methods require large number of electrons per pulse to provide sufficient signal and the ponderomotive interaction requires high laser intensities and is limited by the duration of intense laser pulses.
Proposed Experiment We present a new method for measuring the duration of electron pulses, which holds promise for being applicable to an unprecedentedly broad range of parameters: from femtoseconds to attoseconds, from electronvolts to megaelectronvolts, and from millions of electrons per bunch to single-electron pulses [15]. It is similar to that known as the “attosecond-streak camera” [16,17], an ingenious tool for investigating ultrafast processes. So far the latter has been used for measurement of soft x-ray pulses, characterizing laser fields and investigating ultrafast processes in atoms [18]. The underlying idea is based on the fact that and electron pulse impinging on a solid target will generate, through impact ionization, a pulse of Auger electrons with duration equal to that of the incident pulse convolved with the duration of the Auger decay. The Auger decay can be much faster than the duration of the pulses and is easily accounted for. The energy of an Auger electron created in the presence of a
887
laser field is altered by the field. The spectrum is dependent on the relative time delay between the laser and the electron pulse. This can be used to establish crosscorrelation between electron and laser pulses and thus measure the duration of the electron pulse. Calculation of the spectra following a quantum-mechanical approach yields side bands spaced by the laser photon energy. Such spectra have been seen in the so-called laser-assisted Auger decay [19].
Numerical Results In the following, we present results of the calculations for two different cases, viz. relatively long laser and electron pulses (of order 100 fs), and very short pulses, viz. a few-cycle laser pulse with an as-electron pulse. The "streaked" spectra, generated by plotting the modified Auger spectra as functions of delay between laser and electron pulse, are shown in Fig. 1. The very different parameter ranges illustrate the broad applicability of the method. In the long-pulse case (electron pulse duration longer than the laser cycle), the electron spectrum is broadened, since the electron energies are swept through a large number of laser cycles. In the short-pulse regime (electron pulse duration shorter than laser cycle), the electron spectrum at any particular delay will be broadened and shifted. An Auger line with an appropriate decay time must be chosen for optimum results in the two regimes. The following parameters were used for the calculation of the long-pulse case: Electron pulse duration 80 fs, laser pulse duration 50 fs. Both pulses were assumed to be Gaussian in time. For the Auger transition, the Oxygen KLL line with energy of 500 eV is chosen. The natural width of this line is 0.15 eV corresponding to a decay time of 4.4 fs. The laser intensity is assumed to be 1.6 x 1012 W/cm2 at a wavelength of 800 nm. The maximum energy shift is calculated to ±20 eV. The long pulse case is shown in Fig. 1a. As the overlap between the two pulses increases, more and more side bands appear on both sides of the main Auger line. From such a correlation measurement, the duration of the electron pulse can easily be retrieved if the laser pulse envelope is known.
Fig. 1: Calculated streaked electron spectra. Time zero is defined as the temporal delay at which both pulse maxima coincide. The colour scale is normalized to the intensity of the original Auger line. a) Spectra for long-pulse conditions: Electron pulse duration 100 fs; laser pulse duration 80 fs. Auger line applied oxygen KLL at 500 eV. b) Spectra for short-pulse conditions: Electron pulse duration 800 as; laser pulse duration 5 fs. Auger line Ti KLL at 4.06 keV.
The short-pulse case (Fig. 1b) is calculated with the parameters: electron pulse duration 800 as, a few-cycle cosine-laser pulse with duration of 4 fs at a wavelength of 800 nm. A laser intensity of 1.6 x 1012 W/cm2, was assumed. Note that the laser pulse must be carrier-envelope stabilized, otherwise the temporal resolution would only be given by its duration. For the Auger transition, the titanium KLL Auger line at
888
4.06 keV with a natural width of 0.94 eV and a corresponeing decay time of 702 as is chosen. Here, the maximum energy shift of the electrons is ± 57 eV. Features quite different from the long-pulse case are now observed: The main change in the spectrum is an energy shift which occurs in synchronism with the laser vector potential. Slight broadening results from the combined smearing of electron pulse duration and Auger decay time over the laser cycle. For successful realization of the experiment, the number of Auger electrons generated must be high enough to produce a spectrum in an appreciable amount of time. We calculated the number of Auger electrons per pulse electron taking into account the cross-section for the generation of a K-hole, the fluorescence yields for oxygen and titanium, the number density of the solid and the escape depth of the Auger electrons from the solid. The result is, that with an electron gun running at 1,000 electrons per pulse at a kHz repetition rate or with 1 electron per pulse at a MHz repetition rate, and an electron spectrometer with an acceptance angle of about 1 sterad, a complete spectrum can be recorded in less than an hour.
Conclusion In conclusion, we present a new method of electron pulse duration measurement. By drawing on laser assisted Auger electron emission induced by impact ionization, the method can be used for determining electron pulse durations in a broad range of parameters of electron energy and pulse duration. Electron energies may range from a few keV up to highly relativistic ones in the GeV range and pulse durations from a few 100 fs to sub fs. The method also works independently of the number of electrons per pulse. 1 2 3 4 5 6 7 8 9 10 11 12 13 14 15 16 17 18 19
A. H. Zewail, Ann. Rev. Phys. Chem. 57, 65 (2006). J. R. Dwyer, C. T. Hebeisen, R. Ernstorfer, M. Harb, V. B. Deyirmenjian, R. E. Jordan, and R. J. Dwayne Miller, Phil. Trans. R. Soc. A 364, 741 (2006). V. A. Lobastov, R. Srinivasan, and A. H. Zewail, PNAS 102, 7069 (2005). Y. Okano, Y. Hironaka, K. G. Nakamura, and K. Kondo, Appl. Phys. Lett. 83, 1536 (2003). C. G. Serbanescu and R. Fedosejevs, Appl.Phys. B 83, 521 (2006). H. Schwoerer, B. Liesfeld, H.-P. Schlenvoigt, K.-U. Amthor, and R. Sauerbrey, Phys. Rev. Lett. 96, 014802 (2006). R. Schoenlein, W. Leemans, A. Chin, P. Volfbeyn, T. Glover, P. Balling, M. Zolotorev, K. Kim, S. Chattopadhyay, and C. Shank, Science 274, 236 (1996). D. Kim, C. Toth, and C. P. J. Barty, Phys. Rev. A 59, R4129 (1999). T. Watanabe, M. Uesaka, J. Sugahara, T. Ueda, K. Yoshii, Y. Shibata, F. Sakai, S. Kondo, M. Kando, H. Kotaki, and K. Nakajima, Nucl. Instrum. Meth. in Phys. Res. A 437, 1 (1999). H. C. Lihn, P. Kung, C. Settakorn, H. Wiedemann, and D. Bocek, Phys. Rev. E 53, 6413 (1996). D. X. Wang, G. A. Krafft, and C. K. Sinclair, Phys. Rev. E 57, 2283 (1998). J. van Tilborg, C. B. Schroeder, C. V. Filip, C. Toth, C. G. R. Geddes, G. Fubiani, E. Esarey, and W. P. Leemans, Phys. Pasmas 13, 056704 (2006). I. Wilke, A. M. MacLeod, W. A. Gillespie, G. Berden, G. M. H. Knippels, and A. F. G. van der Meer, Phys. Rev. Lett. 88, 124801 (2002). C. T. Hebeisen, R. Ernstorfer, M. Harb, T. Dartigalongue, R. E. Jordan, and R. J. Dwayne Miller, Opt. Lett. 31, 3517 (2006). P. Reckenthaeler, M. Centurion, V. S. Yakovlev, M. Lezius, F. Krausz and E. E. Fill, Phys. Rev. A 77, 042902 (2008). J. Itatani, F. Quere, G. L. Yudin, M. Y. Ivanov, F. Krausz, and P. B. Corkum, Phys. Rev. Lett. 88, 173903 (2002). M. Kitzler, N. Milosevic, A. Scrinzi, F. Krausz, and T. Brabec, Phys. Rev. Lett. 88, 173904 (2002). R. Kienberger, et al., Nature 427, 817 (2004). J. M. Schins, P. Breger, P. Agostini, R. C. Constantinescu, H. G. Muller, G. Grillon, A. Antonetti, and A. Mysyrowicz, Phys. Rev. Lett.73, 2180 (1994).
889
Nanoscale Spatial Effects of Pulse Shaping Daan Brinks1, Fernando D. Stefani1, and Niek F. van Hulst1,2 1
ICFO – Institut de Ciencies Fotoniques, Mediterranean Technology Park, 08860 Castelldefels (Barcelona), Spain. E-mail: [email protected] 2 ICREA - Institució Catalana de Recerca i Estudis Avançats, 08015 Barcelona, Spain E-mail: [email protected] Abstract. Commonly used pulse-shaping techniques create a coupling between the spatial and temporal characteristics of the shaped femtosecond laser pulse. Consequently, measured apparent responses to shaped pulses that might seem produced by time domain molecular dynamics could in fact be due to spatial changes in the electric field.
Introduction
Shaping ultrashort laser pulses has opened the possibility of influencing molecular dynamics through interaction with light. To steer molecular dynamics one requires to shape fs pulses in the time-frequency domain. Most commonly used pulse shapers are based on spatial light modulators (SLM)[1] and Acousto-Optic Programmable Dispersive Filters (AOPDF)[2]. In both techniques, the spatial and temporal characteristics of the shaped pulse are coupled. Since detection of the effect of shaped pulses necessarily takes place in a limited spatial region, i.e. a focus or the overlap volume of a pump and a probe beam, this coupling can influence the outcome of experiments. It is therefore important to know the magnitude of this spatio-temporal coupling.
Spatio-temporal coupling
Figure 1: Illustration of the spatial effect of pulse shaping in an AOPDF. The angle between the input and output beam and the different positions of refraction in the AOPDF crystal create a chirped output beam.
In an AOPDF, the transit time of light through a birefringent crystal is controlled by polarization rotation through acousto-optic interaction at a certain position inside the crystal. Different travelling times for different frequency components cause a phase modulation of the pulse; the efficiency of the acousto-optic interaction for different frequency components causes an amplitude modulation[3].
890
Since the Poynting vectors of the optical and the acoustic waves are aligned and not their k-vectors[4], the acousto-optic interaction causes a slight change in direction; i.e. each frequency component that undergoes the acousto-optic interaction exits the crystal at a different position on the output facet. This effect is depicted in figure 1. This phenomenon complicates the interpretation of experiments in several ways: 1. Different positions across the beam profile will be dominated by different frequencies, which can influence the outcome of multi-photon processes. 2. Different frequency components will have different k-vectors in a focus, influencing the efficiency of interaction with (transition) dipoles depending on their respective orientation. 3. Spectral amplitude modulation causes the transverse beam profile to change. As a result every temporal pulse shape will also have a different transverse spatial profile.
Spatial shaping effects in a focus
In order to quantify the magnitude of the spatio-temporal coupling and its possible effects on experiments, we perform simulations of shaping operations as they are typically used in investigations of molecular dynamics. We simulate common shaping configurations and calculate the spatial effect of compensating chirp by phaseshaping and creating phase-locked pulse pairs and multipulse-sequences by amplitude modulation.
0 fs
2
-10000 fs
2
Figure 2: Frequency-space beam profile showing the frequency content versus the spatial dimension in which the frequencies have been dispersed for shaping (x-axis). Shown are (a) an unshaped pulse and (b) a pulse with -10000 fs2 chirp added. The pulse chirped in time is clearly also chirped in space.
In figure 2 the coupling between temporal and spatial chirp after shaping is illustrated. The plots display the intensity of the frequency components of the laser pulse as a function of the position along the shaping coordinate x. Figure 2a shows a Gaussian laser pulse of 10 fs centered at 800 nm and with 10000 fs2 chirp (as can be acquired in the optics of a typical optical setup). Figure 2b shows the same pulse after adding -10000 fs2 chirp compensation in the shaper. Clearly, the shaped pulse presents different spectral components at different positions across the beam. In Coherent Control and Multidimensional Spectroscopy experiments, multi-pulse sequences are the basic constituents of the pulse shapes used to probe and control molecular dynamics [1,5]. Therefore, we took the pulse of figure 2b and simulated the shaping into a multi-pulse sequence and the subsequent imaging in a focus.
891
Figure 3: Spatial difference profiles in the focal plane of a 1.4 NA objective. (a): difference between the focal profile of a phase locked double pulse and a single chirped 10 fs pulse. (b): difference between the focal profile of a phase locked quadruple pulse a single chirped 10 fs pulse.
In figure 3 we compare the intensity distribution of a focused multi-pulse sequence to a focused single pulse, all in the focal plane of a 1.4 NA objective. The normalized differences (i.e. (Imulti-pulse – Isingle pulse)/ (Imulti-pulse + Isingle pulse)) are plotted as a function of position in the diffraction limited focus. The single pulse is in all cases a 10 fs pulse, as in figure 2b. In figure 3a the multi-pulse consists of two such pulses with a 10 fs delay and phase locked at 0 rad. In figure 3b the multi-pulse consists of four pulses with an interpulse delay of 10 fs and again phase locked at 0 rad. Clearly changing the shape of the double pulse creates local changes in the intensity of up to 40%. This effect grows to 60% in the case of the quadruple pulse. Any molecule exposed to a field varying locally like this will respond accordingly. It is worth noting that the variations are not only very strong, but depending on the position in the focus can be very local: positions in the focus that are 10 nanometer apart already experience field intensities developing very differently with varying pulse shapes.
Conclusions
To our best knowledge, current pulse shaping techniques, used in for instance coherent control, create a coupling between the temporal and spatial characteristics of the shaped pulses. Manifest effects include spatial chirp and different spatial intensity profiles for different temporal pulse shapes. The spatio-temporal coupling will influence the outcome of experiments, as not all molecules exposed to the shaped pulses experience the same changes in field strength for varying pulse-shapes. Interestingly, the development of the local field as a function of temporal pulse shape can differ profoundly for positions in a tight focus that are separated by no more than 10 nm. 1 2 3 4 5
J. Herek et al., Nature 417, 533, 2002. V.I. Prokhorenko et al., Science 313, 1257, 2006. F. Verluise et al., Opt. Lett 25, 575, 2000. F. Verluise et al., J. Opt. Soc. Am. B, 17, 138, 2000. N.F. Scherer et al., J. Chem Phys. 95, 1487, 1991
892
Designer Femtosecond Pulse Shaping Using Grating-Engineered Quasi-Phasematching in Lithium Niobate Łukasz Kornaszewski1, Markus Kohler1, Usman K. Sapaev2, Derryck T. Reid1 Ultrafast Optics Group, School of Engineering and Physical Sciences, Heriot-Watt University, Edinburgh, EH14 4AS, UK 2 Laser-Matter Interaction Laboratory, NPO Akadempribor, Academy of Sciences of Uzbekistan, Tashkent 700125, Uzbekistan E-mail: [email protected] 1
Abstract. The generation of tailored femtosecond pulses with fully engineered intensity and phase profiles is demonstrated using second-harmonic generation of an Er:fibre laser in an aperiodically-poled lithium niobate crystal in the undepleted pump regime. Second harmonic pulse-shapes including Gaussian, stepped, square and multiple pulses have been characterised using cross-correlation frequency-resolved optical gating and shown to agree well with theory.
Introduction Designer pulse shaping requires the independent manipulation of the spectral intensity and phase of an optical waveform. Aperiodically-poled quasi-phasematched (QPM) crystals have already been used to compress pulses produced by second-harmonic generation (SHG) [1], and to create sub-ps pulse sequences and shaped multi-ps pulses [2]. This concept can be extended to shape individual fs pulses by controlling the local duty cycle and the period in a QPM grating, as we previously showed theoretically [3, 4]. We now present the first experimental validation of this approach. Our earlier work [3, 4] described a theoretical model containing a crystal transfer function that took account of each domain size and its position in the crystal, evaluated using the analytical formula [3, 4]: n1 d ijk 11n exp [i k Q n ] ∑ 21m exp [ i k Q m] 1 E CRYS = k m=1
{
}
where: is SHG/cnSHG, dijk is the absolute value of the nonlinear coefficient; k () is the magnitude of the wavevector mismatch in the process, and Qm is the end position of domain m in the grating which contains a total of n domains, as described in [4]. Using the crystal transfer function it is easy to calculate the SH pulse, EOUT(t), generated from an input pulse, EIN(t), by a grating characterised by ECRYS() [5]: E OUT t= F
1
{F [ E
2 IN
t ] E CRYS }
2
where F and F-1 are direct and inverse Fourier-transform operators, respectively.
Crystal design and experiment Following the procedure outlined in [4] we used a simulated annealing algorithm to find the appropriate grating designs for 9 different target SH pulses. The design assumed an unchirped 150 fs Gaussian fundamental pulse centred at 1530 nm which
893
was similar, but not identical to the actual pulses used in the experiment. The 4 mm long APPLN crystal was phasematched for an input wavelength of 1530 nm and had 9 gratings of differing lengths. The design (inset, Fig. 1) was optimised for the outputs: short / long Gaussian pulses (1, 2), stepped / square pulses (3, 4), triangular pulses (5), identical but oppositely chirped pulses (6, 7), and multiple (8, 9) pulses.
Fig. 1. Experimental configuration. Pulses from the Er:fibre oscillator are compressed and then focused into the APPLN crystal. After collimation, the second-harmonic (SH) and the fundamental wave (FW) light enter a Michelson interferometer. A dichroic mirror (DM) acts as a beamsplitter, and the nonlinear mixing occurs in a 100 µm-thick BBO crystal. FW+SH denotes the sum-frequency beam resulting from nonlinear mixing in the BBO crystal. L1 and L2 are lenses with focal lengths of 100 mm and 15 mm respectively. Inset: schematic of the APPLN crystal, and the target SH profiles designed assuming 150 fs Gaussian input pulses. The first few domains are shown schematically for grating 6.
Fig. 1 shows the experimental arrangement. The compressed output pulses from an Er:fibre oscillator were focused into the SHG crystal, after which the fundamental and SH pulses were mixed in a BBO crystal and the XFROG traces recorded (Fig. 2).
Fig. 2. XFROG traces for all 9 gratings, with insets showing the expected pulse shapes.
894
Results and discussion Retrieved results are presented in Fig. 3. The measurements indicated qualitative agreement between experiment and theory, but to allow a quantitative comparison we used equations (1) and (2), along with the complex field amplitude of the Er:fibre pulses determined from the XFROG measurements, to evaluate the expected SH pulse shapes, also shown in Fig. 3.
Fig. 3. XFROG measurements of the second-harmonic pulses (solid thick line — intensity; open circles — phase), compared with the shapes calculated using the fundamental pulses (dotted thin line – intensity; filled circles – phase).
The agreement between the experimental and calculated SHG pulses is generally good, in terms of both their amplitude and phase. However, for shorter grating designs there is a significant difference probably due to a relatively long region of non-poled material which does not contribute to the pulse shaping but rather only adds unwanted dispersion to already created pulse.
Conclusions In conclusion, the use of fully grating-engineered crystals for femtosecond pulse shaping is a simple and robust alternative to an adaptive optics system, and our results show the strong potential of this technique. Extension of the technique to the highdepletion regime will widen its applicability. 1 2 3 4 5
M. A. Arbore, A. Galvanauskas, D. Harter, M. H. Chou and M. M. Fejer, Opt. Lett. 22, 1341 (1997). G. Imeshev, A. Galvanauskas, D. Harter, M. A. Arbore, M. Proctor, and M. M. Fejer , Opt. Lett. 23, 864 (1998). D. T. Reid, J. Opt. A 5, S97 (2003). U. K. Sapaev and D. T. Reid, Opt. Expr. 13, 3264 (2005). M. A. Arbore, O. Marco and M. M. Fejer, Opt. Lett. 22, 865 (1997).
895
Direct Measurement of Spectral Phase for Ultrashort Laser Pulses Based on Intrapulse Interference Bingwei Xu, Vadim V. Lozovoy, Yves Coello, and Marcos Dantus Michigan State University, Department of Chemistry. East Lansing, Michigan 48824, USA E-mail: [email protected] Abstract. We present a method for the direct spectral phase measurement of ultrafast laser pulses. The second-derivative of the unknown spectral phase is revealed by the experimental 2D-contour plot and can be measured without mathematical manipulation.
Introduction Pulse characterization and compression are of critical importance in ultrafast laser science and technology. Among the available spectral phase characterization techniques, the early development of Yamada and coworkers [1], and the development of FROG [2] and SPIDER [3] represent milestones in the field. There are also optimization algorithm approaches to achieve pulse compression with or without spectral phase measurements [4, 5]. Ideally, a spectral phase measurement should be simple, direct and relatively insensitive to noise. Here, we report on such a method, based on a simple chirp scan. It has long been known that nonlinear optical (NLO) processes are sensitive to the second derivative of the phase. If we introduce a reference function f"(ω) to measure the second derivative of the unknown phase φ"(ω), NLO processes are maximized at frequency ω when the equation φ"(ω)-f"(ω)=0 is satisfied [6, 7]. Since f"(ω) is known, the value of φ"(ω) can be easily obtained. Based on this observation, our group has been using the multiphoton intrapulse interference phase scan (MIIPS) method, which typically uses a sinusoidal function for f(ω) [6, 7]. In this contribution, we successively impose a set of quadratic phases (chirp) instead of sinusoidal phases on the ultrashort pulses and record the corresponding nonlinear spectra. The secondderivative of the unknown spectral phase is directly revealed by the experimental 2Dcontour plot resulting from the single chirp scan. An accurate measurement can be extracted from the experimental data without any mathematical treatment or approximation [8]. Given that chirp can be introduced using standard passive optics such as a prism, grating or grism-pair arrangement, this method can be conveniently implemented without the need of an adaptive pulse shaper.
Experimental Methods For the experiments carried out with an adaptive pulse shaper we used a folded allreflective grating-based system containing a grating, a long focal length spherical mirror, and a 640-pixel dual-mask spatial light modulator (SLM-640, CRi Inc.). After the pulse shaper, the laser was focused onto a thin KDP crystal and the second harmonic generation (SHG) signal was directed to a spectrometer. The detailed setup of the experiments described here can be found in publications [8, 9].
896
Results and Discussion Measurements with an adaptive pulse shaper Transform-limited pulses were first obtained by measuring and compensating the spectral phase of the system using the sinusoidal MIIPS method [6, 7]. A 3000 fs3 cubic spectral phase function was then introduced to the pulses via the pulse shaper, and measured with a simple chirp scan. The resulting 2D contour plot is shown in Fig. 1(a). Quantitative results are shown in Fig 1(b) together with the spectrum of the laser. A sinusoidal spectral phase function was also introduced and measured. The experimental trace is shown in fig. 1(c). Fig. 1(d) shows the introduced (black curve) and measured second derivative of the phase (dots). Note that in both cases the experimental traces directly reveal the measured φ″(ω). The accuracy of this method, especially for complex phases, improves with an iterative measurement-compensation routine [8, 9].
Fig. 1. Measurements of φ″(ω). (a) Experimental trace corresponding to a cubic phase defined by φ(ω)=1/6*3000fs3*(ω-ω0)3. (b) shows the introduced (black curve) and measured functions (crosses), the dashed curve corresponds to the spectrum of the laser. (c) Experimental trace corresponding to a sinusoidal phase defined by φ(ω)=5π sin[7fs*(ω-ω0)]. (b) The introduced function is shown in black. Black and grey dots correspond to the functions measured after a single scan and after one iteration, respectively.
Once φ″(ω) is obtained, as demonstrated before, double integration can be used to calculate φ(ω). Fig. 2(a) shows complex spectral phase measurements. For these experiments, two independent pulse shapers were used. One pulse shaper introduced the desired spectral phase, while the other was used to measure it using MIIPS. The agreement between the introduced and measured phases illustrates the performance of the method for the case of complex spectral phases.
897
Fig. 2. (a) Complex spectral phase measurements. The black and grey curves correspond to the introduced and measured phases. The dotted curve corresponds to the spectrum of the pulses (b) Phase measurement without an adaptive pulse shaper. The figure shows the experimental trace obtained after a chirp scan with a grating compressor. The linear feature corresponds to a cubic spectral phase.
Measurements without an adaptive pulse shaper Here, we demonstrate spectral phase measurements without the use of an adaptive pulse shaper. Instead, different amounts of linear chirp were introduced to amplified pulses using the built-in compressor in the regenerative amplifier by varying the spacing between the grating pair. Fig. 2(b) shows the experimental trace obtained after the chirp scan. The linear φ″(ω) dependence indicates the presence of a cubic phase distortion, also known as third-order dispersion (TOD). No effort was made here to eliminate the measured TOD.
Conclusions In conclusion, a new MIIPS implementation based on a simple chirp scan was presented. The corresponding experimental trace directly yields the second derivative of the unknown spectral phase, without any mathematical treatment. Acknowledgements. We gratefully acknowledge funding for this research from the National Science Foundation, Major Research Instrumentation grant CHE-0421047. 1 2 3 4 5 6 7 8 9
898
K. Naganuma, K. Mogi, and H. Yamada, IEEE J. Quantum Elect. 25, 1225, 1989. R. Trebino, and D. J. Kane, J. Opt. Soc. Am. A. 10, 1101, 1993. C. Iaconis, and I. A. Walmsley, Opt. Lett. 23, 792, 1998. T. Baumert, T. Brixner, V. Seyfried, M. Strehle, and G. Gerber, Appl. Phys. B-Lasers O. 65, 779, 1997. D. Yelin, D. Meshulach, and Y. Silberberg, Opt. Lett. 22, 1793, 1997. V. V. Lozovoy, I. Pastirk, and M. Dantus, Opt. Lett. 29, 775, 2004. B. Xu, J. M. Gunn, J. M. Dela Cruz, V. V. Lozovoy, and M. Dantus, J. Opt. Soc. Am. B 23, 750, 2006. V. V. Lozovoy, B. Xu, Y. Coello, and M. Dantus, Opt. Express 16, 592, 2008. Y. Coello, V. V. Lozovoy, T. C. Gunaratne, B. Xu, I. Borukhovich, C. Tseng, T. Weinacht, and M. Dantus, J. Opt. Soc. Am. B 25, 140, 2008.
Two Dimension Spatial Light Modulator with an Over-Two-Octave Bandwidth for High-Powered Monocycle Optical Pulses K. Hazu, T. Tanigawa, N. Nakagawa, Y. Sakakibara, Sh. b. Fang, T. Sekikawa, and M. Yamashita Department of Applied Physics, Hokkaido University, and Core Research for Evolutional Science and Technology, Japan Science and Technology Agency, Kita-13, Nishi-8, Kita-ku, Sapporo, 060-8628 Japan E-mail: [email protected] Abstract. We carried out feedback chirp compensation using a two-dimension spatial light modulator operating in a wavelength range from 260 to 1100 nm, which is useful for the application to ultrabroadband and high-powered optical pulses.
1. Introduction A liquid crystal spatial light modulator (LC-SLM) is a very powerful tool for chirp compensation and shaping of ultrashort optical pulses. LC-SLMs have enabled phase control over the octave broad spectrum to generate monocycle pulses by the feedback (FB) phase control technique using a 4-f phase compensator [1-3]. On the other hand, pulse shaping has been reported for quantum coherent control, high-speed optical information processing and material characterization. Furthermore, the interest in the application for coherent control in high-field physics is increasing and the technique to generate a few mJ pulses with a duration in the 5 fs region has been proposed. However, LC-SLMs have a bandwidth limitation in the UV region because of UV absorption by the liquid crystal and low damage threshold.
2. Structure of 2D-SLM The fabricated UV-LC-SLM has vertically two-channels (a large height of 9.8 mm for two dimension (2D) with 200 µm channel gaps: we also fabricated a one-channel 2D·UV-LC-SLM with the same height) and horizontally 648-pixels (98 µm width: 5 µm pixel gap). A new nematic LC of a mixture of cyclohexane derivatives with fluorine substituents (20 µm thickness) operating in the range from 260 to 1100 nm was sandwiched between two fused-silica glasses (0.5 mm thickness), deposited indium tin oxides (100 nm thickness) and parallel-oriented organic films [4]. The maximum values of phase modulation are 31.7, 13.1 and 7.0 rad at 305, 600 and 1100 nm, respectively [4], which are all larger than 2π rad.
3. Optical damage of 4-f phase compensator with 2D·UV-LC-SLM We investigated optical damage characteristics of the 2D·UV-LC-SLM using an amplified Ti:sapphire laser system (30 fs, 1 mJ, 800 nm, 1 kHz). The result showed that the UV-LC-SLM has no damage under the direct irradiation of the laser pulse with photon density of 29 GW/cm2 for longer than 10 hours. We calculated the photon density of the 1-mJ, 30-fs input pulse in a conventional 4-f phase compensator. For the 4-f compensator consisting of a pair of
899
concave mirrors (f=300 mm) and gratings (300 grooves/mm), the photon density at one pixel of the UV-LC-SLM was 378 GW/cm2. This is much higher than the damage threshold (29 GW/cm2) of the UV-LC-SLM. To solve this damage problem, we introduced cylindrical mirrors (f=1000 mm) instead of concave mirrors, and a fused silica prism instead of gratings into the 4-f folded phase compensator with the 2D·UVLC-SLM [Fig. 1]. The prism has more transmission efficiency than gratings over the broadband, and is able to avoid spatial overlapping of the second-order diffracted light. The photon density in the new cylindrical 4-f phase compensator was 22 GW/cm2 at one pixel of the 2D·UV-LC-SLM under the same condition (30 fs, 1 mJ), which is smaller than the damage threshold. Ignoring losses of optical components, the maximum input pulse energy to be allowed is 1.3 mJ for the new 4-f phase compensator.
4. Feedback chirp compensation experiment in the near-infrared (NIR) region A feedback (FB) chirp compensation experiment was carried out using an amplified Ti:sapphire laser system (40 fs, 220 µJ, 780 nm, 1 kHz). A 23-mm-long fused silica glass was used as a dispersion medium yielding a strong chirp. The spectral phase was characterized by the modified spectral phase interferometry for direct electric-field reconstruction (M-SPIDER) [2]. The delay time, the spectral shear and the spectral shift were 990 fs, 23.6 rad·THz and 2.35 rad·PHz, respectively. Figure 2(a) shows the spectral phase before and after FB compensations. The chirped pulse before FB compensation has a spectral phase variation over 30 rad in the spectral range from 720 to 860 nm and a duration of 118 fs [Fig. 2(b)]. The group delay dispersion (GDD) obtained by curve fitting was 846 fs2 at 800 nm. Those agree with the GDD of the dispersion glass and other optical components. After FB compensation, the spectral phase is almost flat within the smaller GDD than 20 fs2. The corresponding temporal intensity profile shows a duration of 24 fs, which is very close to the transform-limited pulse duration of 18 fs.
5. Feedback chirp compensation experiment in the UV region We carried out a FB chirp compensation experiment in the UV region. We used a nonlinear optical crystal BBO (0.1 mm, type I) to generate UV light (24 µJ, 392 nm, 1 kHz). The spectral phase was characterized by the same method above. The delay time, the spectral shear and the spectral shift were 557 fs, 15.7 rad·THz, 2.41 rad·PHz, respectively. Figure 3(a) shows the spectral phase before and after FB compensations. The chirped pulse before FB compensation has a duration of 270 fs [Fig. 2(b)]. The GDD was 2418 fs2 at 392 nm. Those agree with the GDD of the prism material in the new 4-f phase compensator and other optical components. After FB compensation, the fitted GDD was 730 fs2. The corresponding temporal intensity profile shows a duration of 34 fs, which is close to the transform-limited pulse duration of 21 fs.
6. Conclusions We have investigated optical damage characteristics of the 4-f phase compensator with the UV-LC-SLM. The allowed maximum input pulse energy of 1.3 mJ has been obtained. We also carried out FB chirp compensations in the NIR region and the UV region using the same 2D·UV-LC-SLM. In the NIR region, after FB compensation,
900
the spectral phase becomes almost flat and its temporal duration is 24 fs (TL: 18 fs). In the UV region, GDD approaches zero and its temporal duration is 34 fs (TL: 21 fs). These results suggest that the 2D·UV-LC-SLM enables us to generate subfemtosecond optical pulses by applying it to induced phase and self-phase modulated output, and to generate high-powered optical pulses by applying it to output from an angularly-dispersed, non-colinear optical parametric amplifier. Prism
(fused silica)
3.5 mm
SLM
Cylindrical mirror (f=1000 mm)
2D-LC-SLM(648-pixel) 9.8 mm
8 mm
Fig. 1 The 4-f folded configuration consists of a cylindrical mirror, a prism and a 2D·UV-LCSLM. 20
-200
-100
0
100
spectrum before FB after FB
0
-10
200
-400 -200 0 200 TIME ( fs)
-20 380
WAVELENGTH (nm)
I(t) I FB(t)
INTENSITY (arb. unit)
10
INTENSITY (arb. unit)
I(t) IFB(t)
(b)
PHASE (rad)
PHASE (rad)
spectrum before FB after FB
INTENSITY (arb. unit)
(a)
400
390 400 WAVELENGTH (nm)
410
Fig. 2 NIR region: (a) The intensity spectrum and the spectral phase before and after FB compensations. Inset shows the temporal intensity profiles before and after FB compensations. UV region: (b) The intensity spectrum and the spectral phase before and after FB compensations. Inset shows the temporal intensity profiles before and after FB compensations.
References 1 M. Yamashita, H. Shigekawa and R. Morita, “ Mono-cycle photonics and optical scanning tunnelling microscopy”, M. Yamashita et al. eds. (Springer Verlag, Berlin, 2005). 2 M. Hirasawa, N. Nakagawa, K. Yamamoto, R. Morita, H. Shigekawa, and M. Yamashita, Appl. Phys. B 74, S225 (2002). 3 M. Yamashita, K. Yamane, and R. Morita, IEEE J. Sel. Top. Quantum Electron. 12, 213 (2006). 4 K. Hazu, T. Sekikawa, and M. Yamashita, Opt. Lett. 32, 3318 (2007).
901
Vector Pulse Shaper Assisted Short Pulse Characterization Andreas Galler1 and Thomas Feurer1 1
Institute of Applied Physics, University of Bern, CH-3012 Bern, Siwtzerland E-mail: [email protected]
Abstract. We demonstrate that shaper-assisted pulse characterization is able to imitate most standard pulse characterization methods. If a polarization shaper is used even more complex schemes, such as SPIDER, can be realized.
Introduction In most pulse shaping experiments the shaping of linearly polarized light pulses and their characterization is performed in two separate optical arrangements. Only a handful of experiments have been reported where the pulse shaper was used as an integral part of the diagnostic setup. Those are the multi-photon intra-pulse interference phase scan method [1], the shaper-assisted collinear SPIDER [2], and time-domain interferometry with an acousto-optic modulator [3]. Here, we show that a pulse shaper may be used to imitate most standard short-pulse characterization arrangements. Thus, shaped optical waveforms can be characterized exactly where it is needed since the corresponding nonlinear element can be mounted instead of the sample of interest. In addition, the unique capability of the pulse shaper to control the carrier envelope phase can be used to reduce the number of required sample points in interferometric measurements. Our main objective is to replace a standard short pulse characterization setup by a single nonlinear element and to use the pulse shaping apparatus, first, to create the desired shaped waveform and, second, to produce two replica of the shaped waveform or modified versions thereof and to scan the delay between them. Both operations are linear and can be performed simultaneously by the same device. We show results for scanning second order autocorrelation measurements, for FROG, for STRUT, and for triple correlation measurements. While pulse shaping techniques for linearly polarized light are well established, only recently attention has been devoted to the development of programmable vector pulse shaping techniques [4]. Here, we demonstrate a novel design with full phase and amplitude control of both polarization components of the shaped vector waveform. The design is common path and requires only a single spatial light modulator. Additionally, we demonstrate that such a vector pulse shaper may be used to imitate a SPIDER measurement.
Experiment and Simulation All experiments were performed with a Ti:Sapphire oscillator (KML) delivering pulses with an energy of approximately 1nJ, a bandwidth of 80 nm at a centre wavelength of 820 nm. Linearly polarized laser pulses were phase- and amplitude modulated in a standard pulse shaping apparatus consisting of a double display spatial light modulator (Jenoptik SLM640-d) in the symmetry plane of a 4f zero-dispersion compressor. They were then focused to a suitable nonlinear crystal by an additional lens. Alternatively, the pulses were sent to a polarization shaper. It consisted of three main sections; a 902
polarization beam splitter, a two lens telescope to adjust the angles of incidence on the grating, and a conventional folded 4f spectral filtering arrangement. A schematic of the complete setup is shown in Fig. 1. The polarization of the incoming beam was linear and tilted by 45 deg with respect to the reference frame. After passing through the specially designed Wollaston prism the two orthogonal polarizations were angularly separated. The magnification of the following two lens telescope was adjusted such that the angle between the two beams impinging on the grating was about 3 deg. Then, both spectra in the Fourier plane of the actual pulse shaper covered about one half of the modulator’s active area. This setup allows reaching virtually every point on the Poincar´e sphere.
Fig. 1. ) The polarization shaper consists of a Wollaston prism, a two-lens telescope, and a folded zero dispersion compressor. Amplitude modulation of a given frequency component is realized by, first, the SLM changing its state of polarization and, second, the Wollaston prism directing its undesired polarization component into a different direction.[5]
Results and Discussion In order to imitate a specific characterization scheme it is necessary to identify the transfer function which produces the required pulse sequence, for example two replica with a given time delay between them. The two replica are inherently parallel, which under normal circumstances would lead to an interferometric-type measurement. With the shaper we are able to delay the slowly varying envelope only, which is realized by selecting γ = 0 in the transfer function i 1 h i(ω −(1−γ )ωc )τ /2 e (1) + e−i(ω −(1−γ )ωc )τ /2 . M (ω ) = 2
Varying τ within reasonable limits then results in an intensity-like autocorrelation. When the nonlinear response to the waveform produced by eq.(1) with γ = 0 is recorded an interferometric autocorrelation is obtained. If the nonlinear response is spectrally resolved an interferometric FROG is obtained [6]. This scheme can easily be extended to three pulses with two delay times to adjust, which together with third harmonic generation would yield a so-called triple correlation [7].The following transfer function ½ ω < ω0 − ∆2ω ∨ ω > ω0 + ∆2ω A , (2) M(ω ) = i ωτ A + (1 − A)e ω0 − ∆2ω ≤ ω ≤ ω0 + ∆2ω produces a waveform which results in an interferometric version of a STRUT measurement [8,9] and an example of such a measurement is shown in Fig. 2. 903
Fig. 2.) (a) Shaper assisted STRUT measurement of a laser pulse with a GVD of 3000 fs2 . The position of the maximum correlation signal is indicated by the black line. (b) Comparison between the extracted phase and the phase retrieved from a reference FROG.
By using the vector pulse shaper one can realize pulse sequences where the subpulses have orthogonal polarizations. This is a necessary prerequisite to imitate a SPIDER measurement [10]. We use a type II BBO for the second harmonic generation and mix the two orthogonally polarized waveforms. The ordinary beam consists of two replica of the pulse to be characterized and the extraordinary beam is a chirped pulse. Thus, the transfer function has two parts, £ ¤ (2) (3) Mo (ω ) = 12 eiωτ /2 + e−iωτ /2 and Me (ω ) = eiϕ , with the inter pulse delay τ and the quadratic phase ϕ (2) .The first replica of the ordinary waveform is mixed with a different frequency than the second replica and upon recording the generated second harmonic signal a standard SPIDER trace is recovered. 1 2 3 4 5 6 7 8 9 10
V. V. Lozovoy, I. Pastirk, and M. Dantus, ”Multiphoton intrapulse interference 4; Characterization and compensation of the spectral phase of ultrashort laser pulses,” Opt. Lett. 29, 775 (2004). B. von Vacano, T. Buckup, M. Motzkus, ”In situ broadband pulse compression for multiphoton microscopy using a shaper-assisted collinear SPIDER,” Opt. Lett. 31, 1154 (2006). A. Monmayrant, M. Joffre, T. Oksenhendler, R. Herzog, D. Kaplan, and P. Tournois, ”Time-domain interferometry for direct electric-field reconstruction by use of an acousto-optic programmable filter and a two-photon detector,” Opt.Lett. 28, 278 (2003). T. Brixner, G. Gerber, ”Femtosecond polarization pulse shaping,” Opt. Lett. 26, 557 (2001). M.Ninck, A.Galler, T.Feurer, T.Brixner, ”Programmable common-path vector field synthesizer for femtosecond pulses,” Opt.Lett. 32, 3379 (2007). G. Stibenz, G. Steinmeyer, ”Inteferomteric frequency-resolved optical gating,” Opt. Express 13, 2617 (2005). T. Feurer, S. Niedermeier, R. Sauerbrey, ”Measuring the temporal intensity of ultrashort laser pulses by triple correlation,” Appl. Phys. B 66, 163 (1998). J.L.A. Chilla, O.E. Martinez, ”Direct determination of the amplitude and the phase of femtosecond light pulses,” Opt. Lett. 16, 39 (1991). A. Galler, T. Feurer, ”Pulse shaper assisted short laser pulse characterization,” Appl. Phys. B 90, 427 (2008). C. Iaconis, I.A. Walmsley, ”Spectral phase interferometry for direct electric-field reconstruction of ultrashort optical pulses,” Opt. Lett. 23, 792 (1998).
904
Femtosecond Spectral Interferometry with Attosecond Accuracy by Correction for Spectrometer Resolution Asymmetry Michael K. Yetzbacher, Trevor L. Courtney, William K. Peters and David M. Jonas1 1
Department of Chemistry and Biochemistry, University of Colorado, Boulder, Colorado 80309-0215, USA E-mail: [email protected]
Abstract. Asymmetry in the line spread function of the spectrometer causes delay dependent phase shifts. Fourier deconvolution with the complex-valued optical transfer function allows accurate spectral phase recovery.
Introduction Spectral Interferometry [1] is unique in its ability to measure time delays and constant spectral phase differences between ultrafast pulses [2]. Such measurements are essential to distinguish, for example, a photon echo from a nonlinear free induction decay. Spectral interferomtery is widely used in 2D spectroscopy, SPIDER [3] and OCT. Accuracy of the algorithms for spectral interferometry has been established for dispersion measurements, but has not, to our knowledge, been experimentally tested for linear or constant phase terms. The line-spread function (LSF) of the measuring instrument (including spectrometer and pixellated detector) distorts spectra. The measured spectral interferogram dispersed by a grating spectrometer is given by discrete sampling of I (λ ) = {| E1 (λ ) |2 + | E2 (λ ) |2
(1) + | E1 (λ ) E2 (λ ) | cos[φ1 (λ ) − φ2 (λ )]} ⊗ LSF (λ ) The convolution is written in the wavelength domain as the pixel spacing is approximately constant in λ. The optical transfer function (OTF(ξ)) is the Fourier transform of the LSF. Dorrer et al. [1] developed an algorithm to correct for the influence of the modulation transfer function MTF=|LSF|. When a typical spectrometer is used at other than the (coma-free) design wavelength, it displays an asymmetric LSF [4] which will shift the fringes to one side and generate a complex valued OTF [5].
Experiment Pulses from a Ti:sapphire laser with a ~40 nm wide spectrum were sent into a MachZehnder interferometer which was PZT stabilized using feedback from the difference between the two outputs for a CW laser (632.8 nm, optical period T). Stabilization, measured out of loop with an additional CW laser, was λ/250 at 594.5 nm over 0.1 s. Pulse-pairs with relative delays nT (integer n) were coupled through single-mode fiber into a 0.34m Czerny-Turner spectrograph with a 300 grooves/mm grating. The LSF is measured by adapting an algorithm used to measure the point-spread function (PSF) of the Hubble Space Telescope [6]. Figure 1 shows the measured LSF for the CCD spectrograph, formed by interleaving dithered atomic line spectra.
905
Fig. 1. The instrument response (LSF) determined with an atomic line at 795 nm. The low wings are asymmetric.
Fig. 2. The phase delay error e(ω ) = [δ∆φ (ω )/ω ] − nT recovered from a difference of two spectral phases. Inflection points are spaced with a period of 15 pixels, the stepand-repeat CCD fabrication interval.
Analysis Although there is no unique way to determine the center for an asymmetric function, this poses no difficulty if the same center is used for calibrating and deconvolution. Line centers from another atomic emission lamp are retrieved to within 1/71 of a pixel. Spectra are deconvolved with the LSF by complex division in the wavelength conjugate (ξ) domain. After deconvolution and removal of the first two terms of eq. (1), the data has the form | E1 (λ ) E2 (λ ) | cos[φ1 (λ ) − φ2 (λ )] . Spectral interferograms are processed according to the algorithm in refs. [1,2]. The recovered spectral phase, ∆φ (ω ) = φ1 (ω ) − φ2 (ω ) , is corrected for interferometer dispersion by subtracting a reference spectral phase, ∆φr (ω ) , determined in the same interferometer. This difference, δ∆φ (ω ) = ∆φ (ω ) − ∆φr (ω ) , is then least-squares fit to a line δ∆φ (ω ) = φ0 + ωτ to retrieve the ω=0 intercept phase shift, φ0 , and the slope delay, τ .
Results and Conclusions The magnitude of the Fourier transform of the measured LSF was compared with the fringe amplitude for several interferograms. These two measurements of the MTF agree past the Nyquist limit of 8 ps. Figure 2 shows the phase delay error over the central 80 nm. The phase delay error compares favourably with the ±35 as error reported for 2D spectral shearing interferometry [7]. Figure 3 shows intercept phase shifts with and without LSF deconvolution. Compared to a systematic intercept phase shift error of 1 mrad per fs delay in prior work[2], errors are 20x smaller after deconvolution. These errors can be reduced 3x more by pseudo-deconvolution using the wavelength dependent LSF (the width varies by 10% across the spectrograph).
906
Fig. 3. Delay dependence of the intercept phase shift retrieved with LSF deconvolution (squares) and without (crosses).
Fig. 4. Slope delay error (τ-nT) at several delays without accounting for the LSF (crosses) and after deconvolution of the LSF (squares).
Without correcting for the LSF, slope delay and intercept phase errors are strongly correlated. Figure 4 shows the slope delay error in attoseconds. Note the mirror image similarity to Fig. 3. Interferograms with large time delays are more strongly affected by the LSF. Surprisingly, correction for the MTF alone removes less than half the slope delay error for delays greater than 1.5 ps. Pseudo-deconvolution (using the LSF for each wavelength to deconvolve the entire spectrum for restoration of the undistorted spectrum at that wavelength) reduces the phase delay error significantly, but still leaves the systematically spaced inflection points that likely arise from random CCD step-and-repeat fabrication errors of 100 nm after every 15 pixels (pixel widths are 25 microns). A fine wavelength re-calibration procedure to remove the linear time delay dependence of the phase delay at each pixel can reduce these errors, leading to rms phase delay errors of ±2.4 as, consistent with the measured interferometer stabilization error over 0.1 s and the 1s interferogram acquisition time. With this fine re-calibration, the maximum slope delay error is 20 as. Using SPIDER, delay errors of 10 as can lead to pulse duration errors of 1 fs for single cycle pulses [7]. Thus, our data shows that SPIDER measurements of pulse duration can be significantly affected by the phase of the complex-valued OTF. Further, asymmetry in the instrument response can appear in both dimensions for an imaging spectrometer. Therefore, spatially encoded arrangements [8,9] may be subject to similar distortions by the two-dimensional complex-valued OTF. 1 2 3 4 5 6 7 8 9
C. Dorrer, N. Belabas, J. Likforman, and M. Joffre, J. Opt. Soc. Am. B, 17, 1795, 2000. A. W. Albrecht, J.D. Hybl, S.M. Gallagher Faeder, and D.M. Jonas, J. Chem. Phys. 111, 10934, 1999. P. Baum, and E. Riedle, J. Opt. Soc. Am. B, 22, 1875, 2005. J. Reader, J. Opt. Soc. Am., 59, 1189, 1969. E. Hecht, Optics, 2nd ed., (Addison-Wesley, Reading, 1990). J. Anderson, and I.R. King, Pub. Astronomical Soc. Pacific, 112, 1360, 2000. J.R. Birge, R. Ell, and F. Kartner, in Ultrafast Phenomena XV, P. Corkum, D. Jonas, R.J.D. Miller, A.M. Weiner, eds. Springer, Berlin, 2007. E.M. Kosik, A.S. Radunsky, I.A. Walmsley, and C. Dorrer, Opt. Lett., 30, 326, 2005. P. Bowlan, P. Gabolde, A. Shreenath, K McGresham, R. Trebino, and S. Akturk, Optics Express 14, 11892, 2006.
907
Spatial phase control and applications of highorder harmonics C. Valentin1, J. Gautier1, E. Papalazarou1, Ch. Hauri1, G. Rey1, Ph. Zeitoun1, S. Sebban1, V. Hajkova2, J. Chalupsky2, L. Vysin3 and L. Juha2 1
Laboratoire d’Optique Appliquée, ENSTA-Ecole Polytechbique-CNRS UMR 7639, Chemin de la Hunière, F 91761 Palaiseau Cedex, France E-mail: [email protected] 2 Institute of Physics, Na Slovance 2, Cz 18221 Prague, Czech Republic 3 Faculty of Biomediacal Engineering, Zikova 4, Cz 16636 Prague, Czech Republic Abstract. We present experimental results of control of high-order harmonic wave-fronts. We have reached a spatial phase with average distortions of /7 at 32 nm when controlling the fundamental laser beam wave front. We apply our results to experiments requiring tight focusing conditions.
Introduction The tremendous progress of gas high-order harmonic generation (HHG) based ultrashort light source developed in the Extreme Ultra-Violet (EUV) and soft x-rays, during the past few years, opened up a whole range of applications. Assets such as the short pulse duration, the high spatial and temporal coherences as well as its compactness make HHG a handy laboratory tool with respect to other EUV sources (i.e. synchrotron radiation, EUV-Free Electron Lasers). In this proceeding, we will show how to control EUV beam wave-front using a deformable mirror on the path of the fundamental laser beam. The aim of this experiment is to improve the EUV beam focusing, thus to achieve high intensities. We will then show application experiments of such our harmonic source as digital in-line holography and EUV-induced ablation on PMMA, in attosecond time scale.
EUV spatial phase shaping In order to shape the EUV spatial phase, we need to measure the wave-fronts of our harmonic source. For this, we have developed an EUV Hartman sensor in collaboration with Imagine Optic [1]. We have first calibrated this sensor using highorder harmonics generated in argon =32nm) with /50 accuracy (/120 at 13 nm [2]). Moreover, we have previously demonstrated that the root-mean-square (r.m.s.) values for EUV wave-front distortions do not depend dramatically on the generation parameters (as pressure in the gas, iris diameter) [3]. In standard generation conditions, IR and EUV beams present a strong astigmatism with r.m.s. distortion values of/5 for the IR beam and /4 for the EUV beam (at = 32 nm). For this experiment, we have studied the correlations between the IR and EUV spatial distortions, measuring wave-fronts by a Shack-Hartmann (SH) sensor (HASO, Imagine Optic) for the IR beam and the Hartmann sensor for the EUV beam. We placed a deformable mirror (BIM31, CILAS) on the IR beam before the generation setup. The IR beam is provided by a Ti:Sapphire laser system (810nm, 6 mJ, 35 fs, 1 kHz,). The beam diameter was measured to be 36 mm at 1/e2. The deformable mirror (DM) is used to control the IR beam wave-front: We could either
908
flatten the spatial phase, either impose a specific aberration, thanks to the servo-loop between the SH sensor and the DM. All the measurements were performed when the iris was fully open. Then we generated harmonics around H25 in argon ( = 32 nm) and we measured the EUV wave front in the propagation axis. For these measurements, the iris was closed to a diameter of 15 mm to ensure best phasematching. We present only the results obtained with astigmatism at 0°, corresponding to the 5th Zernike polynomial (cf figure 1). b)
a)
0.3
50
mm
0
-0.2 0
mm
50
Fig. 1. a) Correlations between IR and EUV r.m.s. spatial distortions measured with respect to wavelength , b) EUV wave-front distortions measured by the Hartmann sensor.
Figure 1a shows that IR and EUV r.m.s. distortions measured with respect to wavelength are correlated. The harmonic beam appeared to have less spatial phase distortions than the measured IR ones, because of an apertured at 15 mm IR beam. Moreover, harmonic flux and wave-front depend on IR wave-front but in a non-trivial manner. We measured the flattest EUV wave-front to be /7 r.m.s. distortions (figure 1b) at a wavelength of 32 nm, leading to twice the diffraction limit according to Marechal’s criterion. Very tight focusing is then possible with 40% of the encircled energy in the focal spot.
Applications of high-order harmonic based source We have performed two application experiments using our high-order harmonic beamline: Digital in-line holography and ablation of PMMA. For both applications, we needed to focus efficiently the harmonic beam composed of 5 successive odd orders around 25 (H21-H29). For this purpose, we used a multilayer off-axis parabolic mirror (f=65mm, =5°). Multilayer structure has been calculated in order to keep the attosecond structure of the pulse train [4]. 250 25
a)
b)
mm
µm
0
mm
25
0
µm
250
Fig. 2. a) Raw hologram recorded in 10 seconds, b) reconstructed image of the tip from the hologram, the insert is a zoom on the tip showing a resolution of 740 nm.
909
The first proof-of-principle holography experiment has demonstrated a resolution of 620 nm with a harmonic beam composed of one order (H25) [5]. Using all the harmonics generated in the Argon cell (H21-H29), we have measured the dimensions of a tip placed after the focusing parabola with a resolution of 740 nm, keeping the attosecond structure (cf figure 2). The second experiment, performed in collaboration with L. Juha’s group demonstrates that ablation of poly (methyl methacrylate) PMMA with high-order harmonic beam is possible. The figure 3 shows the AFM image of the ablation hole with dimensions 2 µm x 3 µm obtained with 6 104 laser shots. The total focused harmonic flux is then over the damage threshold of PMMA at 32 nm (2 mJ/cm2) [6]. We measured simultaneously the harmonic beam wave-front with r.m.s. distortions of /6 after the parabola (figure 3b). The shape of the crater corresponding to the focal spot can be compared to the point Spread Function (PSF) calculated by the Hartmann sensor software (figure 3c). The two measurements are well correlated.
a)
5 b)
0.3
1
µm
µm
-5-5
5 c)
µm
5 -0.3
-5
µm
5 0
Fig. 3. a) AFM (Atomic force microscopy) image of ablation hole in PMMA performed by an EUV beam around 32 nm, b) wave-front measurement after the offaxis parabolic mirror showing r.m.s. distortions of /6 at 32 nm, c) calculated PSF given by the Hartmann sensor software.
Conclusions In summary, strong correlations between IR beam and generated harmonic beam have been highlighted. Controlling the EUV wave-front using a deformable mirror for IR beam is then possible. Improvements are foreseen for both application experiments: Getting more flux and controlling the spatial phase will allow single shot experiments. They pave the way to studies of bio-molecules with sub-micron and attosecond resolutions. 1 2 3 4 5 6
910
Imagine Optic : www.imagine-optic.com. P. Mercère et al., Opt. Lett. 28, 1534, 2003.. J. Gautier et al., Euro. Phys. J. D, DOI: 10.1140/epjd/e2008-00123-2, 2008. A.-S. Morlens et al., Opt. Lett.31, 1558, 2006. A.-S. Morlens et al., Opt. Lett.31, 3095, 2006. J. Chalupsky et al., Opt. Exp. 15, 6036, 2007.
A New Generalized Projections Algorithm Geared Towards Sub-100 Attosecond Pulse Characterization Justin Gagnon1, Vladislav S. Yakovlev1,2, Eleftherios Goulielmakis1, Martin Schultze1 and Ferenc Krausz1,2 1 2
Max-Planck-Institut für Quantenoptik, D-85748 Garching, Germany Department für Physik, Ludwig-Maximilians-Universität München, D-85748 Garching, Germany Email: [email protected]
Abstract: We developed a new algorithm for characterizing attosecond pulses from streaked spectra. We compare our algorithm to the current one used for attosecond characterization, and show that it is better suited for sub-100 attosecond pulses.
Introduction Since their first experimental demonstration, extreme-ultraviolet (XUV) attosecond pulses have become increasingly shorter due to improvements in technology. This progress calls for advances in methods used to characterize these pulses. Currently, the standard characterization technique is attosecond streaking, first proposed by J. Itatani et al. [1]. When electron spectra are "streaked" by an infrared (IR) laser field, and are measured for several delays between the IR and XUV pulses, they contain all the information about the XUV and IR fields. Characterization techniques previously developed for femtosecond light pulses [2,3] were adapted for the need of attosecond metrology [4,5]. One of the most general and versatile retrieval algorithms is FROG CRAB [4]. While this algorithm was demonstrated for ~130 as pulses [6], we found that several underlying approximations and assumptions need to be revised in order to accurately characterize ever shorter and broadband attosecond XUV pulses. In this paper, we propose an improved implementation of the attosecond FROG retrieval that is free from several important drawbacks of the original algorithm. We show that our improvements are necessary in order to rely on the retrieved IR and XUV fields, and we establish the range of reliability of the attosecond FROG technique Electron spectra recorded at different delays form a streaking spectrogram. The streaking spectrogram can be described, in atomic units, by [7]
,
(1)
where EX(t) is the electric field of an XUV pulse and A(t) is the vector potential that describes the infrared (IR) streaking pulse. In order to process the spectrogram S(p,) with a FROG algorithm, there cannot be terms inside the integral that depend both on momentum and time. An inspection of (1) reveals two such terms: d(p + A(t)) and pA(t). We remove the momentum dependence of these terms by making the susbstitution p p0, where p0 is the central momentum of the unstreaked electrons.
Methods The Principal Components Generalized Projections Algorithm (PCGPA) [8] has long been the standard used for the inversion of FROG traces and has recently been applied to attosecond streaking spectrograms [4]. However, this algorithm is not optimized for attosecond streaking. First of all, by its very construction the PCGPA requires the spectrogram to satisfy periodic boundary conditions with respect to the delay. Although this condition holds true for
911
conventional FROG applications, the expression for S(p,) shows that the spectra do not necessarily satisfy periodic boundary conditions. Furthermore, the PCGPA requires the spectrogram to be interpolated so that its energy and delay steps, , obey the sampling relation (in atomic units). For a typical attosecond streaking range of = 70 eV, this implies a nominal delay step of ~ 60 as, which may be difficult to achieve experimentally. In practice, delay steps of ~100 as or greater are used in order to speed up the acquisition time, but most importantly to minimize the drift of experimental conditions. Our findings demonstrate that interpolating the spectrogram along the delay axis dramatically affects the quality of the retrieval. To alleviate these shortcomings, we have developed a new version of generalized projections which retains the overall robustness of other algorithms, but obviates the need to interpolate the spectrogram along the delay axis, and does not assume periodic boundary conditions for the gate. Generalized projections algorithms employ the strategy of enforcing alternating constraints between the time and frequency domains [8] in order to find a pulse and gate pair that reproduce the measured spectrogram. A signal matrix is initially computed from time-sampled guesses for the pulse P and gate G. In our algorithm, the signal matrix S is obtained by shifting the pulse elements with respect to the gate elements by a time interval equal to the delay step, according to the prescription Sj,i=PjGj+L(i-1). The inclusion of the integer “L”, which is the number of time samples contained in a delay step, makes it possible to avoid interpolating the spectrogram along the delay axis during the preprocessing stage.
Results To compare our algorithm to the PCGPA, a streaking spectrogram was simulated using a 750 nm laser field with a peak intensity of 2.4×1013 W/cm2, and a FWHM duration of 4 fs. The spectrogram's 65 spectra were calculated at delay intervals of 100 as. The attosecond XUV pulse consisted of a sequence of two pulses, which is a typical feature of attosecond pulses obtained through spectral filtering of high harmonics [9]. The FWHM duration of the stronger pulse was 85 as, and weaker one had a duration of 77 as. Figure 1 shows a comparison between the pulses retrieved by (A) our algorithm, and (B) the PCGPA. The same parameters were used for both algorithms. However, due to the coarse delay step of 100 as, the spectrogram that was fed to the PCGPA had to be interpolated to 256 delays points to satisfy the sampling constraint , whereas our algorithm only made use of the 65 known spectra. The PCGPA was unable to correctly retrieve the pulses and the vector potential because it was given three interpolated spectra for every know spectrum.
Fig. 1. Retrieval of a train of two pulses (A) and vector potential (B) using our algorithm and the PCGPA
Although our algorithm avoids the interpolation of the spectrogram, another source of error arises due to the substitution p p0 in (1). To investigate the effect this approximation has on the retrieved XUV and IR fields, we gave our algorithm a synthetic spectrogram calculated without making this substitution. To reproduce realistic experimental conditions, we again
912
chose a double XUV pulse structure, separated by a half-cycle of the IR field. The XUV pulses were given phases with higher-order terms. The most intense one had a duration of 90 as, whereas the weaker one was more extended in time, with a temporal intensity profile modulated on the attosecond time scale. The theoretical dipole transition matrix element of Neon was used1, and the IR streaking field was modeled with a 3.5 fs Gaussian pulse with a peak intensity of 2.4×1013 W/cm2, and a central wavelength of 750 nm. The spectrogram was again calculated at delay intervals of 100 as. The results of the retrieval are shown in Fig. 2.
Fig. 2. Retrieved XUV pulse (A) and vector potential (B) from a spectrogram calculated without the approximation p p0, and using the dipole transition matrix element of Neon.
As a consequence of replacing the momentum p with the central electron momentum p0, the retrieved attosecond XUV pulses and IR vector potential exhibit slight temporal distortions, but most importantly the relative phase between the attosecond pulses is off by approximately 90 pA(t) p0A(t) in expression (1), which entails an error in the phase acquired by the electrons during their interaction with the laser field [10].
Conclusion We have shown that our implementation of the attosecond FROG retrieval technique is a reliable and accurate method for characterizing ever shorter attosecond pulses. Constant improvements in technology will allow for temporally resolving physical phenomena on scales approaching the atomic unit of time ( 24 as). Acknowledgements. This work was supported by the DFG Cluster of Excellence: MunichCentre for Advanced Photonics. The authors are grateful to X. Gu and F. Krausz. 1 2 3 4 5 6 7 8 9 10
J. Itatani et al., Phys. Rev. Lett., Vol. 88, 173903, 2002. R. Trebino and D. J. Kane, J. Opt. Soc. Amer. A, Vol. 10, 1101, 1993. C. Iaconis and I. A. Walmsley, IEEE J. Quantum Electron., Vol. 35, 501, 1999. Y.Mairesse and F. Quéré, Phys. Rev. A, Vol. 71, 0011401(R), 2005. F. Quéré et al., Phys. Rev. Lett., Vol. 90, 073902, 2003. G. Sansone et al., Science, Vol. 314, 443, 2006. Markus Kitzler et al., Phys. Rev. Lett., Vol. 88, 173904, 2002. D. J. Kane, IEEE J. Quantum Electron., Vol. 35, 421, 1999. M Schultze, et al., New J. Phys., Vol 9, 243, 2007. J Gagnon, E Goulielmakis and V. S. Yakovlev, Appl. Phys. B, Vol. 92, 25, 2008.
1The dipole transition matrix element from the ground state of neon was calculated in the Hartree-Fock (HF) approximation. The continuum states were modeled as a linear combination of the s and d waves forming a frozen-core HF solution propagating in the direction of observation (courtesy of Y. Komninos).
913
Characterization of Mid-Infrared Pulses by TimeEncoded Arrangement Kevin F. Lee1,2, Adeline Bonvalet1,2, and Manuel Joffre1,2 1
Laboratoire d’Optique et Biosciences, Ecole Polytechnique, Centre National de la Recherche Scientifique, 91128 Palaiseau, France 2 Institut National de la Santé et de la Recherche Médicale, U696, 91128 Palaiseau, France E-mail: [email protected] Abstract. We characterize mid-infrared pulses using upconversion to the visible regime by mixing with two collinear time-delayed replicas of an 800 nm chirped pulse. The phase is encoded as a function of the time-delay.
Introduction Mid-infrared laser light is an important tool for studying molecules. To characterize the relative phases of the frequency components within a laser pulse, there are techniques such as modified-SPIDER (spectral phase interferometry for direct electric-field reconstruction) [1] and Zero-Additional Phase (ZAP) SPIDER [2], which have previously been adapted to the mid-infrared [3] by upconverting to the visible region, where standard charged-coupled device (CCD) detectors can be used. Here, we adapt two-dimensional spectral shearing interferometry (2DSI) [4] to the mid-infrared with upconversion to the visible. To differentiate from related 2D techniques such as spatially encoded arrangement for SPIDER (SEA-SPIDER) [5,6], we will call our method time-encoded arrangement SPIDER (TEA SPIDER) [7]. TEA SPIDER has two key advantages over ZAP SPIDER, the ability to calibrate the frequencies involved, and a simpler collinear optical arrangement.
TEA SPIDER In this experiment, we want to characterize a 3.9 µm pulse, which was generated by taking the difference frequency of the signal and idler from an optical parametric amplifier driven by an 800 nm regenerative amplifier. We normally make measurements in the infrared using chirped-pulse upconversion (CPU). This involves focusing the mid-infrared light in a nonlinear crystal along with a strongly chirped 800 nm pulse which, in our case, is uncompressed output from the regenerative amplifier. Sum-frequency generation gives a visible pulse near 664 nm, the spectrum of which we measure for each laser shot using a spectrometer and a CCD. To perform a TEA SPIDER measurement, we add a Michelson interferometer to the 800 nm beam, giving us two collinear replicas at 800 nm, with a time delay controlled by a motorized delay stage in one arm of the interferometer. If either of the interferometer arms is blocked, the experiment returns to being a chirped pulse upconversion detection scheme. With both arms unblocked, the two pulses will cause spectral interference that is a function of the time delay. The desired phase information is carried in this interference pattern.
914
Fig. 1. TEA SPIDER trace of a chirped mid-infrared pulse recorded by the spectrometer as the pulse delay between the 800 nm pulses is varied. The tilt of the fringes away from the horizontal corresponds to the group delay of the pulse at that frequency, as shown by the approximate scale on the right. A large pulse delay corresponding to a 2.8 THz shear frequency was chosen to improve the visibility of the fringe tilt.
An example of a raw TEA SPIDER trace is shown in Figure 1. If there was no dispersion in the mid-infrared pulse, the fringes would be horizontal. The pulse measured in Figure 1 was purposely chirped by passing the beam through a 2 cm piece of CaF2, resulting in tilted fringes. The two-dimensional phase Φ(ω,τ) of a TEA SPIDER trace is written as [4]: Φ(ω,τ) = Ω1τ + φ(ω - Ω1) - φ(ω - Ω2) ≈ Ω1τ + Ω dφ/dω
(1)
where Ω1 and Ω2 are the frequencies of the quasi-monochromatic part of the 800 nm fields with which the mid-infrared pulse is mixing, with Ω1 associated with the moving arm. The difference of these two frequencies is the shear frequency Ω = Ω2 Ω1. φ is the phase of the mid-infrared pulse, and τ is the time delay between the 800 nm pulses. Noting that a fringe corresponds to a constant value of Φ, the group delay can be found by multiplying the delay τ by Ω1/ Ω, giving the group delay axis on Figure 1. Looking more carefully at the TEA SPIDER trace, we can convert this to the relative phase of the different frequency components. The analytic signal can be retrieved from the raw data by doing a Fourier transform with respect to time, zeroing components at negative frequencies, and transforming back. The result is the phase derivative which can be integrated to retrieve the desired phase. This phase retrieval and the measured upconverted spectrum, was used to create the reconstructed pulse shown in Figure 2.
915
Fig. 2. The pulse intensity (—) and phase (---) of a chirped mid-infrared pulse as reconstructed by using TEA SPIDER. The data used here was an average of 1000 laser shots, with a shear frequency of 0.5 THz. The TEA SPIDER trace also provides a measure of the frequency component of the adjustable pulse with which the mid-infrared pulse is mixing. The fringe spacing is Ω1τ, so it gives directly the frequency Ω1, assuming a good knowledge of the translation stage delay. For very good precision, one might add a beam from a helium-neon laser to the interferometer, and use the resulting beating to continuously measure the translation stage motion, effectively using the helium-neon wavelength as a reference for the 800 nm beam, and subsequently the mid-infrared spectrum.
Conclusions When using mid-infrared pulses with chirped pulse upconversion, TEA SPIDER is the most convenient way to characterize the mid-infrared pulse. With the addition of an interferometer to the chirped pulse, one can calibrate the frequencies involved, and measure the relative phases in the mid-infrared beam. 1 2 3 4 5 6 7
916
M. Hirasawa, N. Nakagawa, K. Yamamoto, R. Morita, H. Shigekawa, and M. Yamashita, Appl. Phys. B. 74, S225, 2002. P. Baum and E. Riedle, J. Opt. Soc. Am. B. 22, 1875, 2005. K.J. Kubarych, M. Joffre, A. Moore, N. Belabas, and D.M. Jonas, Opt. Lett. 30, 1228, 2005. J.R. Birge, R. Ell, F.X. Kartner, Opt. Lett. 31, 2063, 2006. E.M. Kosik, A.S. Radunsky, I.A. Walmsley, C. Dorrer, Opt. Lett. 30, 326, 2005. A.S. Wyatt, I.A. Walmsley, G. Stibenz, G. Steinmeyer, Opt. Lett. 31, 1914, 2006. K.F. Lee, K.J. Kubarych, A. Bonvalet, M. Joffre, J. Opt. Soc. Am. B. 25, A54, 2008.
Intensity and phase measurements of the spatiotemporal electric field of focusing ultrashort pulses Pamela Bowlan,1 Ulrike Fuchs,2 Pablo Gabolde,1 Rick Trebino1 and Uwe D. Zeitner2 1
Georgia Institute of Technology, School of Physics, 837 State St NW, Atlanta, GA 30332, USA 2 Fraunhofer-Institut für Angewandte Optik und Feinmechanik, Albert-Einstein-Str. 7, 07745 Jena, Germany Email: [email protected]
Abstract: We demonstrate a spectral interferometer with NSOM probes for measuring focusing ultrashort pulses with high spatial and spectral resolution. We measure a 0.44 NA focus and, for the first time to our knowledge, we observe the forerunner pulse.
Introduction Nearly all ultrashort pulses are used at a focus, where their intensity is highest. And because the quality of many experiments depend on the quality of the focus, it is important to be able to measure the pulse at the focus. But this has remained a difficult challenge since the origin of the field of ultrafast optics. Focused pulses can easily have extremely complex spatiotemporal structure when lens aberrations are present[1, 2]. As a result, simply making one measurement vs. time (or frequency) and another vs. space is not a sufficient characterization of the pulse; a complete spatiotemporal measurement must be made to characterize a focused pulse. And because pulses are routinely focused to spot sizes less than a few microns, a technique with sub-micron spatial and high temporal resolution is needed. Recently we demonstrated a technique for measuring the spatiotemporal field of a focusing ultrashort pulses, which we call SEA TAPDOLE[3]. SEA TADPOLE is a high spectral resolution and experimentally simple version of spectral interferometry. It uses fiber optics to introduce the unknown and reference beams into the device. The fiber can also be used to spatially resolve the unknown focusing beam, yielding spatial resolution equal to the fiber mode size. This is done by scanning the fiber that collects the unknown beam so that an interferogram is measured at every x, y, and z along the focusing beam’s cross section. Then E(t) can be found for each fiber position which gives us the spatiotemporal field of the focusing pulse or E(x,y,z,t). In our original setup, we used single-mode optical fibers with a mode size of 5.6 µm, which limited our measurements to foci with NAs less than 0.12. Here we replace the fiber that samples the focusing beam with single-mode fiber with a nearfield scanning optical microscopy (NSOM) probe at its tip[4]. The NSOM probe is essentially an aperture on the end of the fiber, which can be as small as 30 nm in diameter. Using NSOM fibers, SEA TAPDOLE can measure even tightly focused pulses, so that previously unmeasured complex effects, such as the fore-runner pulse, can now be observed for the first time.
917
Experimental Results We built our SEA TADPOLE using an NSOM fiber probe that was purchased from Nanonics and it had an aperture diameter of 500nm. The fiber used in this probe was the same as the fiber that we use to collect the reference beam. Fig. 1 shows the experimental setup that we used.
Fig. 1. Experimental setup for scanning SEA TADPOLE: The reference and unknown pulses enter the device via equal-length, single-mode optical fibers. The unknown pulse is collected with an NSOM fiber which has an aperture diameter that is smaller than the focused spot size. The NSOM probe is scanned in x (y) and z directions, so that multiple measurements of the field can be made at different positions in space. Once inside SEA TADPOLE, in the horizontal dimension, the light is collimated and then spectrally resolved at the camera using the grating and the cylindrical lens. In the vertical dimension, the light emerging from the two fibers crosses at a small angle and makes horizontal spatial fringes at the CCD camera.
We focused our Ti:Sa oscillator pulse with a bandwidth of 26 nm (FWHM) and a spot size of 4 mm (FWHM) using an aspheric lens purchased from New Focus with a focal length of 8 mm, yielding a NA of 0.44. A 4X telescope was used to increase the spot size of the beam before the aspheric lens. We placed a 25µm pin hole at the focus of the telescope to remove any aberrations that may have been introduced by the first lens (the higer NA one) in the telescope. The second lens in the telescope had an NA less than 0.01 and therefore its aberrations were negligible. We put a beam splitter after the telescope to obtain a reference pulse and loosely focused it into a singlemode optical fiber. The tightly focused pulse was sampled with an identical fiber, which had the NSOM probe on its end. The fiber was scanned in x and z, so that E(x,ω) was measured at nine different values of z. To check our experimental results, we performed a simulation using the method described in reference [5]. Because we did not focus through a microscope side, as this lens is designed to do, there is a little spherical aberration present in this focus, though the main distortion is chromatic aberration. To determine the aberration parameters of this lens for our simulation, we did ray tracing using the program OSLO and the lens coefficients that were supplied by New Focus. Figure 2 shows the results of our simulations and experiments.
918
Fig. 2. Experimental (bottom) and simulation (top) results: Each box shows E(x,t) as a function of x and t at a distance z from the geometric focus where the color indicates the instantaneous frequency. A pulse preceding the main pulse is observed.
The data shows E(x,t) where the color indicates the instantaneous frequency and shows that the pulse is chirped which is due to the glass in the lens. The color also varies with transverse position (x) due to the chromatic aberrations. The spot size of the redder colors is bigger than that for the bluer colors before the focus, because blue focuses before red. Due to the combination of chromatic aberration and overfilling a lens, there is an additional pulse that is before the main pulse that appears before the focus[1]. This pulse is often referred to as the “forerunner pulse”. The spot size of the (at one t) additional pulse is smaller than 1µm in some places which illustrates our sub-micron spatial resolution. Acknowledgements.Pamela Bowlan acknowledges support from the NSF fellowship IGERT-0221600 and would like to thank the OSA for permission to use content from several of its publications which are listed below. 1 2 3 4 5
Z. Bor and Z. L. Horvath, "Distortion of femtosecond pulses in lenses. Wave optical description," Opt. Commun. 94, 249-258 (1992). U. Fuchs, U. D. Zeitner and A. Tuennermann, "Ultra-short pulse propagation in complex optical systems," Opt. Express 13, 3852-3861 (2005). P. Bowlan, P. Gabolde and R. Trebino, "Directly measuring the spatio-temporal electric field of focusing ultrashort pulses," Opt. Express 15, 10219-10230 (2007). P. Bowlan, U. Fuchs, R. Trebino and U. D. Zeitner, "Measuring the spatiotemporal electric field of tightly focused ultrashort pulses with sub-micron spatial resolution," Opt. Express 16, 13663-13675 (2008). M. Kempe and W. Rudolph, "Femtosecond pulses in the focal region of lenses," Phys. Rev. A 48, 4721-4729 (1993).
919
Polarization, ionization and spatial gates in single attosecond pulse generation Valer Tosa1, Carlo Altucci2, and Raffaele Velotta2 1
National Institute for R&D Isotopic and Molecular Technologies, Donath 71-103, 400293 Cluj-Napoca, Romania, E-mail: [email protected] 2 CNISM-Dipartimento di Scienze Fisiche, Università di Napoli “Federico II”, via Cintia,26 80126, Napoli, Italia E-mail: [email protected], [email protected]
Abstract. We show that in polarization-gating technique, ionization dynamics and threedimensional propagation effects act as additional filters in single attosecond pulse generation. We propose a novel laser field configuration generating single harmonic bursts from multicycle laser pulses, allowing the use of laser pulses up to 25 fs to yield a single XUV pulse.
Introduction High Harmonic Generation (HHG) in gas can be understood by the so-called threestep model [1]. In the first step the electron tunnels through the Coulomb potential barrier lowered by the slowly varying laser electric field, then gains kinetic energy moving in the laser field and finally returns to the vicinity of the ionic core where recombines and emits a burst of light shorter than one femtosecond. From this description it is clear that a single attosecond pulse can only be achieved if one is able to select the HHG emission within half an optical cycle (o.c.) of the laser field. Using multi-cycle pulses to obtain a single attosecond pulse would represent a great step forward, due to their typical high energy per pulse and their commercial availability even with stabilised CEP. In this paper we propose a new approach to achieve isolated attosecond pulses from multi-cycles driving laser sources delivered by commercial laser systems, which combines a polarisation gating generated on the leading edge of the driving pulse with a subsequent ionisation gating whose role is to prevent the emission of additional attosecond pulses from the rest of the pulse. Such a method proves to be effective with Fig. 1 Scheme of the proposed method to generate single pulse duration up to 25 fs attosecond pulse from multi-cycle laser pulses. and can be used with intensity as high as 1015 Wcm-2 or potentially higher. A scheme of principle of our method is reported in
920
Figure 1, where two copies of a linearly polarised, multi-cycles laser pulse are differently chirped and superimposed with crossed polarisation in a single Michelson interferometer. The frequency modulation (chirping) of one of the two beams (continuous line) is achieved when the beam crosses a piece of suitable material (G), whereas the adjustable delay between the two arms is realised by the translational stage in the left arm. Eventually, the output of the interferometer is a superposition of two crossed linearly polarised beams, one of which is chirped and delayed. This leads to an electric field with a time varying polarisation; as sketched in Figure 1 (output arm).
Results and Discussion Appropriate values of the experimental parameters can realise linear polarisation gates which can be correlated with the evolution of the ionization. This is shown in the Figure 2(a) where the ellipticity =Ey(t)/Ex(t) is reported for a laser pulse with p=20 fs. The two pulses are in phase (=0) only for approximately half a cycle thus realising the single emission, which is the main requirement for the generation of an attosecond pulse. The main recollision events leading to single attosecond pulse occur in the time window centred at t-3.5 o.c, since in the previous window centred at t-6.3 o.c. the pulse intensity is still too low for an effective HHG. Moreover, in the earlier window the two fields are in phase with their maximum rather than with their minimum, the second condition being more effective in HHG [2]. The central part of Fig. 2. (a) Ellipticity (dashed line), electron fraction (solid the pulse is intense thick line), and single dipole emission (solid thin line), as enough to fully ionise the function of time for a pulse of 20 fs and CEP=0.4 rad, medium, while the delay other conditions as described in text. (b) Chirped (Ex, solid between the pulses line) and chirp-free (Ey, dotted line) squared electric fields. assures that the ellipticity is quite high (0.7) at the peak of the combined pulse. The above two factors impede further contributions to the atomic polarization from the rest of the pulse. Such a description is supported by the single dipole emission, (see Figure 2(a)) calculated within the Strong Field Approximation [3,4] generalized for a field with timedependent ellipticity [5]. Non-adiabatic, three-dimensional propagation of fundamental and harmonic fields through to the gas target has been accounted for by extending the model [6] for linear polarization to the case of time-dependent polarization. Medium ionization has been calculated by using the non-adiabatic Ammosov-Delone-Krainov model [7]. Throughout our simulations we have assumed a pulse peak intensity at focus of 81014 W·cm-2, a 1-mm long Ar jet of 3.3103 Pa local pressure placed at 1.7 mm in a diverging beam of 3.5 mm confocal parameter.
921
The filtering action of field propagation, which is essential to the formation of a single, short and wellshaped attosecond pulse, is illustrated in Figure 3 where on-axis, single dipole emission is shown versus time for three different medium lengths. These data elucidate the physical mechanism leading to the formation of a single attosecond pulse at the end of the medium, strongly guided by transient phasematching of all the emitted elemental HHG contributions. The single dipole emission is characterised by a complicated multi-peak structure in time, varying both along the propagation and the radial directions, but only the first main peak always occurs at the same time, i.e. t–3.5 o.c., thus implying that only this Fig. 3. (top) Single dipole emission versus time for harmonic field contribution three different Ar medium lengths. (bottom) Harmonic experiences the same phase near and far field illustrating the effect given by a 0.7 at each medium slab. mm iris placed at 0.5 m distance. For clarity, far field is In conclusion, in our shifted on intensity axis. method a polarization gating still induces a train of bursts in the single atom response. A strong ionization suppress bursts developed during the second half of the pulse, while the propagation acts as an additional, spatial filter both in axial and in radial direction, giving rise to a strong single attosecond pulse. Our calculations show that the scheme is robust against carrier to envelope as well as intensity fluctuations, and that pulses up to 25 fs can be used to generate a single isolated attosecond pulse. The experimental scheme is easy to implement in a typical ultrafast laser facility, and therefore represents a vital step to enable the access to attoscience to a wide scientific community. 1 2 2 3 4 5 6
922
P.B. Corkum, F. Krausz, Nature Physics 3, 381, 2007. C. Altucci, V. Tosa, R. Velotta, Phys. Rev A 75, 061401(R), 2007. P.B. Corkum, Phys. Rev. Lett. 71, 1994, 1993. M. Lewenstein, P. Balcou, M. Yu. Ivanov, A. L’Huillier, P.B. Corkum, Phys. Rev. A 49, 2117, 1994. P. Antoine, A. L’Huillier, M. Lewenstein, Phys. Rev. A 53, 1725, 1996. V. Tosa, V., H.T. Kim, I.J. Kim, C.H. Nam, Phys. Rev. A 71, 063807 2005. M.V. Ammosov, N.B. Delone, V.P. Krainov, Sov. Phys. JETP 64, 1191, 1986.
All dispersive mirrors compressor for femtosecond lasers V. Pervak1, C. Teisset2, A. Sugita2, F. Krausz1,2, A. Apolonski1,2 1
Ludwig-Maximilians-Universitaet Muenchen, Am Coulombwall 1, D-85748 Garching, Germany E-mail: [email protected] 2 Max-Planck-Institut für Quantenoptik, Hans-Kopfermann-Str. 1, D-85748 Garching, Germany Abstract. We report on the development of highly dispersive mirrors for chirped-pulse amplifiers (CPA). The designed mirrors are potentially capable of replacing the prisms in the existing CPA compressors making them more compact and stable.
Introduction High-energy femtosecond laser systems deal with dispersion of the materials in pulse compressors, in spectral broadening stages, either inside the laser oscillator cavity or outside. In all cases, the amount of the (absolute) dispersion to be compensated usually grows with the pulse energy. In the case of negative-dispersion oscillators, or chirped-pulse Ti:Sa oscillator CPO, such a monotonic dependence had already been proven theoretically and experimentally [1-5] providing in such a way a stable soliton-like intracavity pulse. In kHz systems the compressor size and its intrinsic dispersion grows with energy because of both the size of the dispersive components and the propagation distance through the components to be used. In high-energy femtosecond Ti.Sa oscillator-amplifier systems, usually prism or grating compressors are in use because they allow compensating large material dispersion of the order of (2-5)x104 fs2 in a wavelength bandwidth of interests (~40 nm), resulting in sub-60 fs pulses. For a broader spectral range, uncompensated third-order dispersion (TOD) becomes so large that the pulse can not be compressed down to the targeted duration (usually 20-30 fs) and additional high-dispersive (hereafter: chirped) mirrors must be added. An alternative to this approach can be an all-chirped-mirror (CM) compressor. “Standard” CMs with the group delay dispersion (GDD) of the order of -50-100 fs2 cannot be used in compressors of such type or/and in high-energy (of µJ-level) oscillators due to a large number of bounces required. To make the problem clear, let us make a rough estimate of the throughput of the compressor equipped with “standard” CMs. For the dispersion to be compensated of the order of 2x104 fs2, the number of necessary bounces must be as high as 200. For a typical CM reflectance of 0.995, it leads to a throughput of only 0.995200 = 0.37. After 500 bounces which one needs to compensate a material dispersion of 5x104 fs2 with 100 fs2-CMs, the throughput of less than 10% becomes completely unacceptable. A second, even more important obstacle, is that the initial pulse will be completely destroyed after such amount of bounces due to accumulated deviations of the realized GDD curve from the targeted one. Meantime, based on the progress in the CM development [6-13], one can expect CMs with dispersion values of ~103 fs2 and reflectivity of some 0.995 for a spectral range of at least 50 nm around a central wavelength of 800 nm. Such mirrors, being realized, could replace prisms and gratings in the compressors giving
923
thus a compact cheap device with a stable output beam free of residual TOD. The absolute value of the CM dispersion can be even higher as the spectrum becomes narrower, as it happens in a case of Yb:YAG disk oscillator [14], where mirrors with the dispersion around 103 fs2 per bounce were successfully demonstrated for the spectral width of several nm. Due to the fact that CMs have higher losses and lower damage threshold in comparison to Bragg (= high) reflectors, CMs with GDD of the order of up to 104 fs2 are desirable for that spectral range in order to decrease the total amount of chirped mirrors in the oscillator cavity.
High dispersive chirped mirrors As a first step in direction of highly dispersive mirrors (HDCM) for CPA, we demonstrate the usability of highly dispersive mirrors for high-energy femtosecond oscillators, namely for i) CPO [3] and ii) an Yb:YAG disk oscillator. In both these oscillators GDD to be compensated is around 2x104 fs2, of the order of the nominal dispersion present in the CPA compressor. By definition, HDCM is characterized by a high group delay of different spectral components. Because the delay is proportional to the optical thickness of the layers involved, HDCM has thick layers and a big total multilayer structure thickness. From that we can formulate the conditions of manufacturing HDCMs: we need a very stable deposition process allowing us to deposit a thick multilayer structure with high accuracy. The total amount of GDD of a HDCM compressor needed for obtaining chirp-free high-energy pulses out of a Ti.Sa CPO is of the order of 2.5x104 fs2 at 800 nm and this value is achievable with only 20 bounces of the HDCM shown in Figure 1 (a).
Fig. 1 (a,b). The calculated GDD and reflectivity of HDCMs. Left: HDCM for Ti:sapphire CPO, right: for Yb:YAG oscillator.
We have now to prove that such amount of bounces will not deteriorate the incident test chirp-free pulse in terms of its duration and energy. In the analysis, the main part of the dispersion was taken away and only residual GDD ripples were included. For virtual compression experiment, we used an incident 60-fs pulse realized in Ti:Sa CPO [3]. The reflected pulse does not become longer when the ripples are absent or small. Calculations show that after 20 bounces i) the exiting pulse preserves its incident duration and ii) the main pulse contains >95% of the energy from the initial value. Based on the analysis above, we hope for efficient compression of highly-chirped pulses exiting the CPO. In a Yb:YAG oscillator, HDCM will provide enough negative dispersion for keeping high-energy soliton pulse inside the oscillator cavity, Figure 1 (b). A HDCM compressor was successfully applied for compressing 2 ps chirped pulse out of Ti:Sa CPO down to 65 fs pulse at 5.3 MHz repetition rate. The group delay dispersion of the HDCM is -1300 fs2 per 924
reflection @800 nm that represents the highest negative dispersion value in the 40 nm wavelength bandwidth, realized so far. The HDCM at 1030nm has a nominal dispersion of ~ -2000 fs2.Three HDCMs inside the Yb:YAG disk oscillator allowed us to generate stable 6 µJ 800 fs pulses. The low amount of HDCMs allowed us to minimize the losses in the optical part of the oscillator.
Conclusions We demonstrate that the required CPA dispersion of the order of 104 - 105 fs2 can possibly be introduced by a set of high-dispersive chirped multilayer dielectric mirrors offering several advantages including simplicity, alignment-insensitivy, and the potential for increased efficiency. As a first step toward an all-HDCM CPA compressor, we have shown 2 sets of HDCMs with both bandwidth and the main value of the dispersion comparable to what one expect in CPA lasers. The mirrors were manufactured and successfully tested in µJ-level laser oscillators. Acknowledgements. This work was supported by the DFG Cluster of Excellence “Munich Centre for Advanced Photonics” (www.munich-photonics.de). 1
F. Krausz, M.E.Fermann, Th.Brabec, P.F.Curley, M.Hofer, M.H.Ober, Ch.Spielmann, E.Wintner, A.J.Schmidt, IEEE J.Quant. Electron. 28, 2097-2122, 1992. 2 A. Fernandez, T. Fuji, A. Poppe, A. Fürbach, F. Krausz, and A. Apolonski, Opt. Lett. 29, 1366-1368, 2004. 3 S. Naumov, A. Fernandez, R. Graf, P. Dombi, F.Krausz, and A. Apolonski, New J. Phys. 7, 216, 2005. 4 V. L. Kalashnikov, E. Podivilov, A.Chernykh, S.Naumov, A.Fernandez, R.Graf, A.Apolonski, New J. Phys. 7, 217, 2005. 5 A.Fernandez, A.Verhoef, V.Pervak, G.Lermann, F.Krausz, A.Apolonski, Appl.Phys. B 87 395-398, 2007. 6 R. Szipöcs, K. Ferencz, C. Spielmann, and F. Krausz, Opt. Lett. 19, 201–203, 1994. 7 V. Pervak, S. Naumov, G. Tempea, V. Yakovlev, F. Krausz, A. Apolonski, Proc. SPIE Vol. 5963, pp. 490-499, 2005. 8 V. Pervak, A. V. Tikhonravov, M. K. Trubetskov, S. Naumov, F. Krausz, A. Apolonski, Appl. Phys. B. 87, 5-12, 2007. 9 G. Steinmeyer, G. Stibenz, Appl. Phys. B 82, 175-181, 2006. 10 G. Steinmeyer, Appl. Opt. 45, 1484–1490, 2006. 11 N. Matuschek, L. Gallmann, D. H. Sutter, G. Steinmeyer, U. Keller, Appl. Phys. B 71, 509-522, 2000. 12 F. X. Kärtner, N. Matuschek, T. Schibli, U. Keller, H. A. Haus, C. Heine, R. Morf, V. Scheuer, M. Tilsch, and T. Tschudi, Opt. Lett. 22, 831–833, 1997. 13 G. F. Tempea, B. Považay, A. Assion, A. Isemann, W. Pervak, M. Kempe, A. Stingl, and W. Drexler, in Biomedical Optics, Technical Digest (CD) (Optical Society of America, 2006), paper WF2. 14 S. V. Marchese, T. Südmeyer, M. Golling, R. Grange, and U. Keller, Opt. Lett. 31, 27282730, 2006.
925
Optical Mapping of Attosecond Ionization Dynamics by Few-Cycle Light Pulses A.J. Verhoef1, A. Mitrofanov1, E.E. Serebryannikov2, D. Kartashov1, A.M. Zheltikov2, and A. Baltuška1 1
Photonics Institute, Vienna University of Technology, Gusshausstrasse 27-387, A-1040, Vienna, Austria E-mail: [email protected] 2 Physics Department, International Laser Center, M.V. Lomonosov Moscow State University, Moscow, Russia Abstract. Few-cycle light pulses are used to map ultrafast ionization dynamics in the time and frequency domain by all-optical means. Tunneling ionization encodes an attosecond phase mask, suggesting a promising method for attosecond shaping of high-intensity optical fields.
Propagation of high-intensity ultrashort light pulses through ionizing gas-phase media involves a complicated ultrafast dynamics of the laser field in space, time, and frequency domain. Filament formation [1] and compression of laser pulses to fewcycle pulse widths [2] are among the most interesting scenarios of such ultrafast pulse-propagation dynamics. Ionization induced by few-cycle laser fields is, on the other hand, an intriguing ultrafast physical process, whose significance for ultrafast science and technologies is still to be fully realized. In the presence of a highintensity optical field, electrons are released from atoms on an attosecond time scale. Moreover, in the tunneling regime, this process displays a strong sensitivity to the carrier–envelope phase (CEP) of a few-cycle light pulse [3]. In a recent experiment [4], attosecond steps in the ion yield, measured by time-of-flight spectrometry, have been resolved by using an isolated XUV attosecond pump pulse and a few-cycle optical probe pulse. In this work, we focus on the possibility of all-optical mapping of attosecond ionization dynamics by few-cycle light pulses and show that this dynamics encodes an attosecond phase mask that projects itself a) as a frequency sweep and b) as a gradual phase velocity increase of the traveling optical pump field. This information can only be probed in the optical frequency range and cannot be accessed with XUV pulses because of the plasma dispersion n p = (1 − ω p2 ω 2 )1 2 , where np is the plasma refractive index, ωp is the plasma frequency, and ω is the radiation frequency. For potential attosecond spectroscopy applications, another advantage of gaining an attosecond response directly with an optical pulse rather than with an XUV pulse is related to the ω−3 dependence of transition cross-sections and the very low yield ~10-7–10-8 of isolated attosecond XUV pulses. In this work we propose to use the ratio of spectral intensities of low-order harmonics as a measure of tunneling ionization rate. The relation between a step-wise change of the free electron plasma density (Fig. 1a), resulting from a twice-peroptical-cycle tunnel ionization (TI), and the appearance of a low-order harmonic spectrum was first introduced by Brunel [5] and explained in terms of tranverse plasma current. This harmonic radiation emission mechanism is distinctly different from the Corkum model for higher-order harmonic generation [6] that comprises a
926
1.0
Ipu(t)
Strong pump
(a)
0.8
(b)
0.015
ne, Δn [arb. units]
Field intensity, dne/dt [arb. units]
sequence of TI, quasi-free electron acceleration, and re-collision with the parent ion. Since the Brunel mechanism in gas does not rely on a re-collision, this type of harmonic emission, in contrast with the higher-order harmonics [6], is a direct frequency-domain reflection of the TI dynamics. The Brunel type of harmonic emission has been tackled experimentally and theoretically [7, 8], but the contribution of these harmonics has never been disentangled experimentally from other harmonic generation mechanisms, i.e. the Corkum type and the Kerr-nonlinearity type. In this work we employ crossed-beam geometry with a strong linearly polarized few-cycle pump pulse and a weak cross-polarized chirped probe pulse for a background-free detection of the Brunel type of emission (Fig. 1b). This emission process can be readily understood in terms of cross-phase modulation impinged on the probe pulse by a step-like temporal phase (=refractive index change) mask. Therefore, the intensity of the emitted Brunel harmonics scales linearly with the probe pulse intensity and does not depend on the polarization state of the probe pulse.
0.6
0.010
0.4
0.005 0.2
Gas target
Weak probe
Spectrometer
Dn(t)
Ipr(t+t)
0 order
Attosec. Phase mask
t
3w0 5w0 7w0 (2N+1)w0
N ? integer
0.000 0.0
-4
-3
-2
-1
0
1
2
3
4
Time [fs]
Fig. 1. (a) Field intensity (dashed line), electron density ne(t) and refractive index change (bold line), and time derivative of the electron density (thin line) calculated as functions of time for Keldysh parameter γ=1. (b) Schematic of an all-optical measurement of the attosecond phase mask. The observed orders of Brunel harmonics are odd harmonics of the pump and probe frequency ω0.
To assess whether or not practical information on TI dynamics can be extracted from a cross-correlation measurement of Brunel harmonics (Fig. 1b), we developed a full 3-D code to study propagation effects. Fig. 2a shows the dependence between the extent of the harmonic spectrum and the rise time of the phase steps, confirming the intuitively clear conclusion that a faster tunneling dynamics translates into a wider spectrum by causing steeper phase steps. 0.6
IFH/ITH
0.4
0
(a)
-2 θ=T/9
-4 -6 -8 0
Input θ=T/2 5
θ=T/4
ω/ω0
10
15
Spectral intensity (arb. units)
log[spectral intensity] (arb. units)
0.5
0.3 -2
10
10
13
14
10
10
15
14
2
14
2
Ip=3×10 W/cm
-3
10
Ip=2×10 W/cm
-4
10
0.10
(b)
2
Input intensity (W/cm )
0.2
0.15
0.20
0.25
0.30
0.35
Wavelength (μm)
Fig. 2. Numerical simulations. (a) Spectrum of an input ~2-cycle pulse and Brunel harmonic spectra for various phase step rise times θ given as a fraction of the cycle duration T. (b) Spectra of optical harmonics from a 5-fs Gaussian-shape pulse in neon at 300 mbar after 1 mm propagation. Inset shows the spectral intensity ratio of the 5th vs. the 3rd harmonic.
927
Fig. 2b shows the effect of propagation suggesting that for propagation distances in gas on the order of 1 mm the phase mask can survive without being significantly washed out, that would lead to a structure-less spectrum. Our numerical simulations also show that for our experimental conditions the contributions of “direct” crossphase modulation harmonics due to χ, χ, χ,… and cascaded χ nonlinearities can be neglected. Measured and calculated spectro-temporal cross-correlation maps for the 5th and 7th Brunel harmonics in krypton are presented in Fig. 3. Corresponding frequencyunresolved cross-correlation traces obtained by integrating the spectro-temporal maps along the frequency axis for three harmonic orders are shown in Fig. 3e. The intensity ratios of individual harmonics are still to be determined because in the present setup different detectors were used to measure each of the harmonic orders. 15
(a) Delay (fs)
Delay (fs)
0 -5 -10 -15 95
0.5
(b)
10
100
105
0
110
-15 130
115
H5 Experiment 135
0 -5
-15 95
150
155
160
(d)
10
5
-10
145
15
(c) Delay (fs)
Delay (fs)
10
140
Wavelength (nm)
Wavelength (nm) 15
100
105
5
110
Wavelength (nm)
115
0.3
0.2
0 -5 -10
H7 Simulation
H3 H5 H7
0.4
-5 -10
H7 Experiment
(e)
5
Normalized signal (arb. units)
5
-15 130
0.1
H5 Simulation 135
140
145
150
Wavelength (nm)
155
160
-100
-50
0
50
100
150
Delay (fs)
Fig. 3. Measured and simulated cross-correlations of low-order harmonics in a Kr jet. The jet thickness was 2 mm and the krypton pressure 400 mbar. The intensity of the 5-fs 760-nm 180-µJ pump pulse in the interaction region was Ip=5×1014 W/cm2. A 1-µJ pulse replica was stretched to ~20 fs and used as a probe. The pump–probe noncollinearity angle was 20 mrad.
In conclusion, we present the first to our knowledge cross-correlation detection of Brunel harmonic radiation in gas which provides direct information on tunneling ionization dynamics. This all-optical technique is much simpler than photo-electron and ion detection and is particularly promising for studying the TI dynamics in larger molecules (unsuitable for cold targets) and possibly bulk solids (unsuitable for photoelectron spectroscopy). Acknowledgements. This work is supported by the Austrian Science Fund (FWF), grants U33-N16 and F1619-N08. We are grateful to Misha Ivanov, Olga Smirnova, and Paul Corkum for stimulating discussions. 1 2 3 4 5 6 7 8
928
A. Braun et al., Opt. Lett. 20, 73 (1995); J. Kasparian and J.-P. Wolf, Opt. Express 16, 466 (2008). C.P. Hauri et al., Appl. Phys. B 79, 673 (2004); Opt. Express 13, 7541 (2005). G. L. Yudin and M.Yu. Ivanov, Phys. Rev. A 64, 013409 (2001). M. Uiberacker et al.Nature 446, 627(2007). F. Brunel, J. Opt. Soc. Am. B 7, 521 (1990). P.B. Corkum, Phys. Rev. Lett. 71, 1994 (1993). N. Burnett, C. Kan, P.B. Corkum, Phys Rev A 51, R3418 (1995). C.W. Siders, et al., PRL 87, 263002 (2001).
Polarization, Phase and Amplitude Control and Characterization of Ultrafast Laser Pulses Philip Schlup1 , Omid Masihzadeh1 , Lina Xu2 , Rick Trebino2 , and Randy A. Bartels1 1
2
Colorado State University, Department of Electrical and Computer Engineering, Fort Collins CO 80523, USA E-Mail: [email protected] School of Physics, Georgia Institute of Technology, Atlanta GA 30332, USA E-Mail:[email protected]
Abstract. We demonstrate complete control over the polarization, phase and amplitude state of an ultrafast laser pulse using a single, linear spatial light modulator, and introduce a selfreferenced method for characterization the polarization state.
The shaping of ultrafast pulses using programmable pulse shapers [1] has become ubiquitous and indispensable for experiments in diverse fields including physics and biochemistry. Since in many physical systems the full vector field, rather than just the spectral phase, plays a critical role, control over amplitude and phase [2, 3], as well as polarization [4–7], have been independently demonstrated. Simultaneous control over all aspects of the field has recently become feasible [8–10]. By contrast, there have been few techniques reported for the characterization of such polarization-shaped fields. The POLLIWOG method [11] requires a well-characterized reference pulse that may be cumbersome to obtain in practice. We here demonstrate and characterize a polarization-amplitude-phase pulse shaper [10] that offers benefits over published alternatives as it operates in a near-common path geometry, minimizing relative phase instability between the polarization components, while using only a single, linear liquid crystal spatial light modulator (SLM) element. In addition, we introduce a new, self-referenced method for fully characterizing the polarization state of an arbitrary ultrashort pulse [12]. It can be used with any well-established measurement technique that yields intensity and phase information for single polarization components, and involves combining three or more such measurements at different angles of an analyzing polarizer. The shaper is configured as a folded 4- f Martinez stretcher (Fig. 1) using a Si prism as dispersive element. An imaging telescope maps the angular separation between the two orthogonal polarization components passing through a Wollaston polarizer onto the prism so that the components are spatially separated at the SLM (Boulder Nonlinear Systems, Lafayette, CO). The resulting near-common-path, common-optic design has a high relative phase stability. Over-sampling of the spatial modes at the Fourier plane allows us to control the amplitude individual frequency components via spatial diffraction by applying a rapidly-oscillating phase grating, independent of a slowlyvarying phase mask [13]. We insert a λ /2 plate to obtain parallel p polarizations within the shaper to maximize the prism transmission and match the LC-SLM response. The resulting relative group delay yields spectral-interferometry fringes after an analyzing polarizer at the output of the shaper, with which we calibrate the wavelength distribution and pixel phase responses for each polarization [10]. A compensating plate (C in Fig. 1) is inserted to restore close temporal overlap between the shaped pulses. Coarse
929
4f C
Wollaston λ/2 M1
M2
M1
Prism
M3
Modulator
Figure 1: Schematic layout. M1 , M3 , concave, M2 flat mirrors; C compensating plate. adjustments are possible by observing the interference fringes between the two pulses after a polarizer at 45◦ . We determined the remaining delay from a type-II SHG crossFROG measurement, applying linear phases to each polarization to effect the requisite variable relative delay [12]. To characterize the polarization-shaped pulses, we write the arbitrary field in the ˜ frequency domain as E(Ω) = E˜x (Ω)ˆx + rE˜y (Ω)e−i(Ωτ +θ ) yˆ where r, τ and θ are the relative amplitude, delay, and phase offset between the E˜x and E˜y components. Power measurements of the E˜x and E˜ y components and normalizing the reconstructed fields R 2 dΩ = 1 determines the amplitude ratio r = (P /P )1/2 . The es˜ such that |E(Ω)| y x tablishment of τ and θ , to which self-referenced single measurements are insensitive, requires an additional measurement at angle η relative to the x-axis that takes the form rη E˜η (Ω)e−i(Ωτη +θη ) = cos η E˜x (Ω) + r sin η E˜y (Ω)e−i(Ωτ +θ ) . Solving for the relevant variables yields # " rη E˜η (Ω)e−i(Ωτη +θη ) − cos η E˜x (Ω) . (1) −i (Ωτ + θ ) = ln r sin η E˜y (Ω) The imaginary part represents a straight line over frequency Ω, from which the slope τ and intercept θ may be extracted. Manipulating the real part of Eq. (1), we write the −iφ (Ω) and find an˜ ˜ fields in terms of amplitudes and spectral phases as E(Ω) = A(Ω)e other straight-line for τη and θη :¤Ωτη + θη = φx (Ω)− φη (Ω)−cos−1 [Γ(Ω)]. £ 2 2 expression 2 2 ˜ ˜ Here Γ = rη Aη + cos η Ax − r2 sin2 η A˜ 2y /2rη cos η A˜ 2η A˜ 2x . The complete polarization characterization is thus in principle possible from just three measurements using existing methods. We term this approach tomographic ultrafast retrieval of transverse linear E-fields, or TURTLE. In experiments, we find the additional redundant information contained in FROG traces leads to more robust reconstructions using a simple fitting procedure. An example is shown in Fig. 2. The orthogonal components E˜x and E˜y [Fig. 2(a), (b)] were measured and reconstructed using the standard algorithm. FROG traces at angles η = ±45◦ [Fig. 2(c), (d)] were measured and then curve-fit, using the reconstructed E˜ x and E˜ y , with respect to the values of τ and θ . A simplex algorithm minimized the leastsquares difference between simulated and measured traces. In the “phantom” traces of Figs. 2(a)-(d), the left half plane shows the measured FROG trace compared to the reconstruction in the right half. Figure 2(e) shows the three-dimensional reconstruction of the field. We found good agreement between the retrieved delay (τ = 55 fs) and that posted to the SLM mask (50 fs) over a wide range of delays between 5 and 200 fs, as shown in Fig. 2(f). Although independent power measurements for the determination of r are indispensable in some instances (e.g., amplitude-only shaping), simulations 930
0
0 400−400 0 400 t (fs) t (fs)
−1
−400
200 400 0 −200ime (fs) T
150
200 τ (fs)
200 1 0.5 0 −0.5 −1 1
|τ| (fs)
E y (arb.)
(d)
(e)
.)
−400
(c)
(b)
rb
−50 0 50
(a)
(a Ex
Ω (mrad/fs) Ω (mrad/fs)
−50 0 50
0
−200 −2
100
0 2 θ (rad)
50 0
0
50
100 150 τShaper (fs)
200
Figure 2: TURTLE retrieval of arbitrary shaped pulse. (a) E˜x and (b) E˜ y reconstructed FROG traces. (c), (d) Fitted traces to find parameters in Eq. (1). (e) Retrieved temporal field. (f) Measured τ as function of delay applied to shaper; inset, error surface. show that highly-structured FROGs arising from complex amplitude-and-phase shaped pulses can obviate the need for these measurements. We note that the lack of spurious structure in the 45◦ -projection FROGs of Fig. 2(c), (d) indicates good relative phase stability (¿ π rad) between the polarization components over our ∼15-min acquisition time of high-resolution FROG traces. In summary, a single, high-resolution linear LC-SLM was used to achieve complete control over the phase, amplitude, and polarization state of an ultrafast pulse. The individual polarization components are shaped on different sections of the shaper. The near-common-path, common-optic geometry ensures a stable relative delay and phase between the polarization components. We introduced a self-referenced technique to determine the polarization state of an arbitrary ultrafast pulse and used it to characterize the shaped pulses. The technique uses as few as three measurements by existing methods that measure the intensity and phase of a single pulse polarization Acknowledgements. The authors wish to thank Carmen Menoni and Cameron Moore for the loan of the OSA. The authors gratefully acknowledge support from NSF CAREER Award ECS-0348068, ONR Young Investigator Award, the Beckman Young Investigator Award, and Sloan Research Fellowship support for R.A.B. 1 2 3 4 5 6 7 8 9 10 11 12 13
A. M. Weiner, Review of Scientific Instruments Vol. 71, 1929, 2000. J. C. Vaughan, T. Hornung, T. Feurer, and K. A. Nelson, Optics Letters Vol. 30, 323, 2005. J. W. Wilson, P. Schlup, and R. A. Bartels, Optics Express Vol. 15, 8979, 2007. T. Brixner, and G. Gerber, Optics Letters Vol. 26, 557, 2001. L. Polachek, D. Oron, and Y. Silberberg, Optics Letters Vol. 31, 631, 2006. C. G. Slater, D. E. Leaird, and A. M. Weiner, Applied Optics Vol. 45, 4858, 2006. T. Suzuki, S. Minemoto, T. Kanai, and H. Sakai, Physical Review Letters Vol. 92, 133005, 2004. S. M. Weber, F. Weise, M. Plewicki, and A. Lindinger, Applied Optics Vol. 46, 5987, 2007. M. Ninck, A. Galler, T. Feurer, and T. Brixner, Optics Letters 32, 3379, 2007. O. Masihzadeh, P. Schlup, and R. A. Bartels, Optics Express Vol. 15, 18025, 2007. W. J. Walecki, D. N. Fittinghoff, A. L. Smirl, and R. Trebino, Optics Letters Vol. 22, 81, 1997. P. Schlup, O. Masihzadeh, L. Xu, R. Trebino, and R. A. Bartels, Optics Letters Vol. 33, 267, 2008. J. W. Wilson, P. Schlup, and R. A. Bartels, Optics Express Vol. 15, 8979, 2007.
931
Silicon-Chip-Based Single-Shot Ultrafast Optical Oscilloscope Mark A. Foster1, Reza Salem1, David F. Geraghty1, Amy C. Turner2, Michal Lipson2, and Alexander L. Gaeta1 1 2
School of Applied and Engineering Physics, Cornell University, Ithaca, NY 14853 School of Electrical and Computer Engineering, Cornell University, Ithaca, NY 14853 E-mail: [email protected]
Abstract. We demonstrate a single-shot ultrafast optical oscilloscope using a four-wavemixingbased parametric temporal lens integrated on a CMOS-compatible silicon photonic chip. Experimentally, we demonstrate waveform measurement with a 100-ps record length and sub750-fs resolution.
Introduction Measurement of optical waveforms on the sub-picosecond time scale is of great interest for next generation highspeed optical communications and for studies of ultrafast chemical and physical phenomena [1]. Current electronic oscilloscopes are capable of single-shot measurements of waveforms with slightly better than 100-ps resolution. Optical techniques such as frequency resolved optical gating [2] and spectral phase interferometry for direct electric-field reconstruction [3] can provide single-shot measurements of optical waveforms with fewfemtosecond resolution but with record lengths limited to a few picoseconds. To fill this gap in measurement capabilities, researchers have explored techniques using the space-time duality of electromagnetic waves [4-10]. To date the implementations of these techniques for sub-picosecond resolution typically require second-order nonlinear materials and wavelength conversion to non-standard wavelength bands. Furthermore, simultaneous 100-ps record lengths and sub-ps resolution has yet to be demonstrated. Here we demonstrate a single-shot ultrafast optical oscilloscope (UFO) with a 100-ps record length and a resolution considerably better than 750-fs. The implementation relies upon our newly developed parametric time-lens based on the thirdorder nonlinear process of four-wave mixing (FWM). This parametric process yields an output that is generated at a wavelength near those of the pump and input waves. This enables the frequencies of all the interacting waves to be in the S-, C-, and Ltelecommunication bands, which allows for manipulation of all the waves using the wellestablished instrumentation and components available for these bands. Furthermore since all materials posses a third-order nonlinear moment, the FWM based time-lens can be implemented in any material platform including the CMOS-compatible SOI photonic platform used here.
Time-to-Frequency Conversion The temporal processing technique we employ is based on the well-known feature of spatial imaging systems that an object positioned at the front focal plane of a lens will produce the two-dimensional Fourier transform of the object at the back focal plane. Extending the spatial Fourier transform processor to the temporal domain in which diffraction and the spatial lens are replaced by dispersive propagation and the
932
application of a quadratic temporal phase generated here using FWM, respectively, yields a device that converts the spectral (temporal) profile of the input to the temporal (spectral) profile of the output. Since in the temporal domain the Fourier transform of the waveform is its optical spectrum, a measurement of the spectrum at the Fourier plane will provide the temporal profile of the incident waveform and this process is termed time-to-frequency conversion [4]. Since this spectrum does not change during propagation from the lens to the second focal plane, propagation through the second dispersive path is unnecessary and can be removed in a time-tofrequency converter.
Silicon-Chip-Based Ultrafast Optical Oscilloscope Here, we demonstrate a silicon-based single-shot UFO in which the time-to-frequency conversion is implemented using a FWM-based parametric time-lens in a 1-cm-long CMOS-compatible embedded SOI nanowaveguide with a cross-sectional size of 300 nm by 750 nm. The strong optical confinement of these silicon structures allows for highly efficient nonlinear processes and for engineerable group-velocity dispersion (GVD) that can yield conversion bandwidths greater than 150 nm [11]. We have experimentally observed conversion bandwidths approaching 500 nm. The pump and input waves are generated from an optical parametric oscillator (OPO) that produces 150-fs pulses at a 76MHz repetition rate. The pulses from the OPO are spectrally separated into a 280-fs pump pulse with 15-nm of bandwidth centered at 1545 nm and a variable bandwidth signal pulse centered at 1575 nm. The input waveform is passed through a dispersive element consisting of a 240-m length of single-mode optical fiber. For this length to correspond to the focal length of the FWM time-lens, the input wave is mixed with a pump wave, which has been passed through twice the dispersive length of the input signal. The pump wave is amplified using an erbium-doped fiber amplifier (EDFA) and FWM is carried out. The resulting FWM generated spectrum is measured with an optical spectrum analyzer (OSA) for multi-shot measurements or with a monochromator and IR-camera for single-shot measurements. We experimentally characterize the record length and temporal resolution of our system by sending in input waveforms of various complexity. First we inject a two narrowband 900-fs pulses and vary their temporal separation to observe the record length of the device. As shown in Fig. 1(b), we are able to measure the two pulses separated by a record length of 100 ps. To characterize the resolution of the FWM-based UFO, we decrease the signal pulsewidth to 400 fs and then reduce the separation of the pulses. As shown in Fig. 1(c), we are able to clearly differentiate the two 400-fs pulses when separated by 750 fs. To demonstrate the ability of the FWM-based UFO to measure complex pulses, we amplify the short input pulse in an EDFA and induce nonlinear spectral broadening in the amplifier and subsequently pass this pulse through 20-m of optical fiber. The resulting pulses are then measured using the UFO and using cross-correlation with the 280-fs pump pulse. The measurements of three different pulses and the respective cross-correlations are shown in Fig. 2(a)-(c). To further demonstrate the utility of the FWM-based UFO, we reduce the amount of fiber in the dispersive paths to 75 m and 150 m. This new system has a smaller record length but a better temporal resolution due to the reduced effect of TOD. Using this system, we measure the beating between two spectrally separated pulses that are temporally overlapped as shown in Fig. 2(d). To demonstrate the single-shot capability of the device, we replace the OSA used to measure the output spectrum of the UFO with a single-shot spectrometer and we measure a single-shot optical waveform composed of two pulses separated by 3.5 ps as shown in Fig. 2(e).
933
Fig. 1. (a) Collective measurements of a single 900-fs pulse scanned across the 100-ps record length of the system. (b) Measurements of two 900-fs pulses of variable separation. (c) Experimental characterization of the temporal resolution.
Fig. 2. (a)-(d) Comparison of a variety of complex pulses measured using the ultrafast optical oscilloscope and a crosscorrelator. (e) Single-shot measurement of a waveform using the ultrafast oscilloscope and comparison with a multishot cross-correlation. 1
C. Dorrer, “High-speed measurements for optical telecommunication systems,” IEEE Select. Topics Quantum Electron. 12, 843-858 (2006). 2 D. J. Kane and R. Trebino, “Single-shot measurement of the intensity and phase of an arbitrary ultrashort pulse by using frequencyresolved optical gating.” Opt. Lett. 18, 823825 (1993). 3 C. Dorrer, B. de Beauvoir, C. Le Blanc, S. Ranc, J. P. Rousseau, P. Rousseau, J. P. Chambaret, and F. Salin, “Single-shot real-time characterization of chirped-pulse amplification systems by spectral phase interferometry for direct electric-field reconstruction.” Opt. Lett. 24, 1644-1646 (1999). 4 M. T. Kauffman, W. C. Banyal, A. A. Godil, and D. M. Bloom, “Time-to-frequency converter for measuring picosecond optical pulses,” Appl. Phys. Lett. 64, 270-272 (1994). 5 C. V. Bennett, R. P. Scott, and B. H. Kolner, “Temporal magnification and reversal of 100 Gb/s optical data with an upconversion timemicroscope,” Appl. Phys. Lett. 65, 2513-2515 (1994). 6 L. K. Mouradian, F. Louradour, V. Messager, A. Barthelemy, and C. Froehly, “Spectrotemporal imaging of femtosecond events,” IEEE J. Quantum Electron. 36, 795-801 (2000). 7 C. V. Bennett, B. D. Moran, C. Langrock, M. M. Fejer, and M. Ibsen, “Guided-wave temporal imaging based ultrafast recorders,” Conference on Lasers and Electro-Optics, OSA Technical Digest Series (CD) (Optical Society of America, 2007), paper CFF1. 8 B. H. Kolner, “Space-time duality and the theory of temporal imaging,” IEEE J. Quantum Electron. 30, 1951-1963 (1994). 9 J. Van Howe and C. Xu, “Ultrafast optical signal processing based upon space-time dualities,” J. Lightwave Technol. 24, 2649-2662 (2006). 10 C. V. Bennett and B. H. Kolner, “Principles of parametric temporal imaging—Part I: System configurations,” IEEE J. Quantum Electron. 36, 430-437 (2000). 11 M. A. Foster, A. C. Turner, R. Salem, M. Lipson, and A. L. Gaeta, “Broad-band continuous-wave parametric wavelength conversion in silicon nanowaveguides,” Opt. Express 15, 12949-12958 (2007).
934
Time-resolved off-axis digital holography for characterization of ultrafast phenomena in water T. Balciunas1, A. Melninkaitis1, G. Tamosauskas2, and V. Sirutkaitis1 1
Laser Research Centre, Vilnius University, Vilnius LT-10223, Lithuania E-mail: [email protected] 2 Department of Quantum Electronics, Vilnius University, Vilnius LT-10222, Lithuania Abstract. We present the application of time-resolved off-axis digital holography for the investigation of refractive index/ transmission properties of laser-induced plasma filaments in water. Time evolution of both amplitude and phase contrast images of the self-focused beam in water was characterized with temporal resolution better than 50 fs. The images reveal the picture of the early dynamics of plasma.
Introduction The ability to excite matter with ultrashort light pulses and probe its subsequent evolution on the femtosecond time scale has opened up complete new realms of science. However, up to now the in situ characterization of laser-induced refractive index and transmission change with spatial and temporal resolution was a challenging task. We attempt to combine the merits of a “traditional” time-resolved pump–probe technique and holographic phase-contrast imaging to explore light-induced nonlinear changes in the transparent media and to monitor their time evolution [1]. Several other techniques such as time-resolved shadowgraphy [2], scattering-based imaging [3], inline digital holography [4], or conventional holography [5] have been applied for similar purposes. Despite their merits, they all have serious limitations. The advantages of digital holography, compared with conventional holography, are that (a) no wet processing of the data medium is necessary, (b) retrieving the holographic information is possible practically in real time, and (c) phase measurement is quantitative. We employed off-axis experimental design that is superior to in-line holography [4] due to the fact that it does not impose size limitations on the objects to be observed and works well with highly absorptive as well as transparent samples. To our knowledge, this study is the first attempt to implement off-axis digital holography with femtosecond time resolution.
Method Digital holography employs a digital image sensor to record the interferogram, and, subsequently, numerical algorithms described in [5] are used for the reconstruction of the original hologram. The recorded interferogram is usually referred to as a “digital hologram”. The optical layout of the experiment is virtually that of a Mach–Zehnder interferometer. The transparent sample to be characterized is the source of amplitude and phase modulation of the transmitted light, resulting in a so-called object wave O. At the exit of the interferometer the interference between the object wave O and the reference (unperturbed) wave R creates the hologram which is later reconstructed numerically using a computer.
935
The general layout of the experimental setup is shown in Fig. 1. Femtosecond pulses from an amplified Ti:Sapphire laser system were split into two parts: the first was focused in the water cell to generate light filaments, whereas the second one was used to pump the NOPA. The duration of the initial pulses was τ =130 fs. Pump pulses were focused into the water cell using a lens. The cell containing water was placed in one of the interferometer arms for probing.. These pulses were further used for probing in a Mach–Zehnder interferometer. The area around the waist of the pump beam in the water cell was selected for imaging. A delay line was used for changing the delay of the probe pulse. Ti:Sapphire laser
BS P λ/2
M
M
M BS
FS L1
C C D
BS
M R M1
M λ = 800 nm τ = 130 fs NOPA Topas white λ = 550 nm τ = 30 fs
M M
L2 S Fig. 1. Experimental setup of ultrafast digital holography. P - polarizer, λ/2 - half-wave plate, M - mirror, BS – beam splitter, R - delay line, L1 and L2 - lenses, S - water cell, CCD - digital camera, FS - dispersive element.
Results and Discussion Ultrashort pulse propagation in water was observed by registering the interferogram of 30 fs probe pulse imaged on the CCD camera. Phase-contrast and amplitudecontrast images reconstructed numerically from the interferograms are shown in Fig. 2. Very early pulse propagation dynamics can be clearly seen in the phasecontrast images. During the light filamentation, two competing phenomena the optical Kerr effect and plasma generation occur. Their respective contributions carry opposite signs [4]. The total effective refractive index can then be expressed as: n = n 0 + n 2 × I + Δn p (2) where the first term corresponds to the intrinsic refractive index of water, the second term characterizes the influence of the optical Kerr effect (n2 denotes the Kerr coefficient), and the third term is the refractive index change due to the generation of plasma. Due to the high temporal resolution: τprobe
