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Optics Express

Optics Express

  • Editor: C. Martijn de Sterke
  • Vol. 19, Iss. 20 — Sep. 26, 2011
  • pp: 18754–18773
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Supercontinuum generation in quasi-phasematched waveguides

C. R. Phillips, Carsten Langrock, J. S. Pelc, M. M. Fejer, I. Hartl, and Martin E. Fermann  »View Author Affiliations


Optics Express, Vol. 19, Issue 20, pp. 18754-18773 (2011)
http://dx.doi.org/10.1364/OE.19.018754


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Abstract

We numerically investigate supercontinuum generation in quasi-phase-matched waveguides using a single-envelope approach to capture second and third order nonlinear processes involved in the generation of octave-spanning spectra. Simulations are shown to agree with experimental results in reverse-proton-exchanged lithium-niobate waveguides. The competition between χ(2) and χ(3) self phase modulation effects is discussed. Chirped quasi-phasematched gratings and stimulated Raman scattering are shown to enhance spectral broadening, and the pulse dynamics involved in the broadening processes are explained.

© 2011 OSA

1. Introduction

The generation of coherent light in the infrared from mode-locked lasers is of considerable interest for applications including frequency comb generation, spectroscopy, and few-cycle pulse generation [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

4

4. J. Price, T. Monro, H. Ebendorff-Heidepriem, F. Poletti, P. Horak, V. Finazzi, J. Leong, P. Petropoulos, J. Flanagan, G. Brambilla, X. Feng, and D. Richardson, “Mid-IR supercontinuum generation from nonsilica microstructured optical fibers,” IEEE J. Sel. Top. Quantum Electron. 13, 738–749 (2007). [CrossRef]

]. Absolute frequency calibration is usually achieved by self-referencing, for example via 1f–2f or 2f–3f interferometry [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

,2

2. T. Fuji, J. Rauschenberger, A. Apolonski, V. S. Yakovlev, G. Tempea, T. Udem, C. Gohle, T. W. Hänsch, W. Lehnert, M. Scherer, and F. Krausz, “Monolithic carrier-envelope phase-stabilization scheme,” Opt. Lett. 30, 332–334 (2005). [CrossRef] [PubMed]

]. To perform self-referencing and to reach spectral regions not accessible through suitable or well-developed laser gain media, nonlinear-optical methods are usually required. Optical parametric oscillators (OPOs) can be used to span the mid-IR, but are complex and additional steps must be taken for self-referencing and, in some cases, to ensure stability [5

5. C. R. Phillips and M. M. Fejer, “Stability of the singly resonant optical parametric oscillator,” J. Opt. Soc. Am. B 27, 2687–2699 (2010). [CrossRef]

]. Much attention has also been given to χ (3)-based supercontinuum generation in optical fibers [3

3. J. M. Dudley, G. Genty, and S. Coen, “Supercontinuum generation in photonic crystal fiber,” Rev. Mod. Phys. 78, 1135–1184 (2006). [CrossRef]

]. Due to the high nonlinearity and engineerable dispersion available in fibers, spectra spanning multiple octaves can be achieved. This process is limited by the transparency window of the fiber; extending supercontinuum generation to non-silica fibers transparent in the mid-IR is an active area of research [4

4. J. Price, T. Monro, H. Ebendorff-Heidepriem, F. Poletti, P. Horak, V. Finazzi, J. Leong, P. Petropoulos, J. Flanagan, G. Brambilla, X. Feng, and D. Richardson, “Mid-IR supercontinuum generation from nonsilica microstructured optical fibers,” IEEE J. Sel. Top. Quantum Electron. 13, 738–749 (2007). [CrossRef]

].

Compared to fibers, relatively little attention has been given to supercontinuum generation via χ (2) processes in quasi-phasematched (QPM) media, even though very high nonlinearities are readily available in QPM waveguides [6

6. C. Langrock, S. Kumar, J. McGeehan, A. Willner, and M. M. Fejer, “All-optical signal processing using χ(2) nonlinearities in guided-wave devices,” J. Lightwave Technol. 24, 2579–2592 (2006). [CrossRef]

], and supercontinuum generation has been demonstrated experimentally [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

, 2

2. T. Fuji, J. Rauschenberger, A. Apolonski, V. S. Yakovlev, G. Tempea, T. Udem, C. Gohle, T. W. Hänsch, W. Lehnert, M. Scherer, and F. Krausz, “Monolithic carrier-envelope phase-stabilization scheme,” Opt. Lett. 30, 332–334 (2005). [CrossRef] [PubMed]

]. In contrast to methods employing bulk QPM media, QPM waveguides enable highly nonlinear interactions when pumping with commercial mode-locked lasers (including fiber lasers) with few-nJ pulse energies. Additionally, nonlinear interactions have been demonstrated in QPM waveguides using GaAs [7

7. X. Yu, L. Scaccabarozzi, J. S. Harris, P. S. Kuo, and M. M. Fejer, “Efficient continuous wave second harmonic generation pumped at 1.55 μm in quasi-phase-matched AlGaAs waveguides,” Opt. Express 13, 10742–10748 (2005). [CrossRef] [PubMed]

], which has transparency up to 17 μm and therefore offers the possibility of generating a supercontinuum across the mid-IR. In order to reach this goal, a detailed understanding and quantitative modeling of the nonlinear dynamics involved is first required. Progress has been made recently in modeling broadband χ (2) interactions [8

8. M. Conforti, F. Baronio, and C. De Angelis, “Nonlinear envelope equation for broadband optical pulses in quadratic media,” Phys. Rev. A 81, 053841 (2010). [CrossRef]

], but a complete analysis including waveguide effects, competing nonlinearities, and the role that chirped QPM gratings can play in enhancing spectral broadening has not yet been performed. In order to reach the full potential of QPM waveguides for continuum generation, these effects must be modeled so that appropriate QPM gratings and, where necessary, dispersion-engineered waveguides can be designed.

2. Numerical model

In this section, we introduce a model to simulate ultra-broadband nonlinear interactions between multiple waveguide modes using an analytic-signal formalism for forward-propagating waves, similar to that described in Refs. [8

8. M. Conforti, F. Baronio, and C. De Angelis, “Nonlinear envelope equation for broadband optical pulses in quadratic media,” Phys. Rev. A 81, 053841 (2010). [CrossRef]

, 10

10. M. Kolesik and J. V. Moloney, “Nonlinear optical pulse propagation simulation: from Maxwell’s to unidirectional equations,” Phys. Rev. E 70, 036604 (2004). [CrossRef]

13

13. S. Wabnitz and V. V. Kozlov, “Harmonic and supercontinuum generation in quadratic and cubic nonlinear optical media,” J. Opt. Soc. Am. B27, 1707–1711 (2010).

], generalized to describe waveguide interactions with multiple nonlinear optical effects and multiple guided modes. The model automatically accounts for all of the conventional χ (2) interactions including second harmonic generation (SHG), sum frequency generation (SFG), and difference frequency generation (DFG) including intrapulse DFG. All χ (3) interactions can be accounted for as well, but we consider only self and cross phase modulation (SPM and XPM), and SRS. Our analysis is suited to weakly-guided modes such as in RPE LiNbO3 waveguides, but could be generalized to describe tightly-confined modes.

First, the spectrum of the total electric field is expanded in terms of a sum of waveguide modes
E˜(x,y,z,ω)=12nBn(x,y,ω)A˜n(z,ωωref)exp[i(βrefωref/vref)z],
(1)
where Ẽ is a Fourier component of the real-valued electric field E(x,y,z,t) (tilde denotes a frequency-domain field quantity) and the optical frequency ω > 0. The real-valued transverse spatial mode profiles of the bound modes are given by Bn and the z-dependent envelopes by Ãn. We assume that Ãn(z, ω < 0) = 0 (i.e. Ãn are analytic signals). We have also defined a reference propagation constant β ref, a reference group velocity v ref, and a carrier frequency ω ref; appropriate choices for these reference quantities are discussed below. From Ref. [14

14. G. Imeshev, M. M. Fejer, A. Galvanauskas, and D. Harter, “Pulse shaping by difference-frequency mixing with quasi-phase-matching gratings,” J. Opt. Soc. Am. B 18, 534–539 (2001). [CrossRef]

], the nonlinear polarization can be expressed in the frequency domain as
P˜NL(ω)/ɛ0=χ(2)(ω,ω',ωω')E˜(ω')E˜(ωω')dω'+χ(3)(ω,ω',ω",ωω'ω")E˜(ω')E˜(ω")E˜(ωω'ω")dω'dω".
(2)

Since the magnitude of the second-order nonlinearity has a spatial dependence due to non-idealities of the waveguide fabrication process [9

9. R. V. Roussev, “Optical-frequency mixers in periodically poled lithium niobate: materials, modeling and characterization,” Ph.D. thesis, Stanford University (2006), http://nlo.stanford.edu/system/files/dissertations/rostislav_roussev_thesis_december_2006.pdf.

], we have introduced a normalized susceptibility χ¯(2)=|χ(2)(x,y)/χ0(2)|, where χ0(2) is the relevant tensor element in the unperturbed material. For RPE LiNbO3 waveguides, |χ (2)(x,y)| is negligible close to the upper surface of the crystal, down to a depth h WG below the surface [9

9. R. V. Roussev, “Optical-frequency mixers in periodically poled lithium niobate: materials, modeling and characterization,” Ph.D. thesis, Stanford University (2006), http://nlo.stanford.edu/system/files/dissertations/rostislav_roussev_thesis_december_2006.pdf.

]. Therefore, χ¯ (2) (x,y) = 0 for y < h WG and χ¯ (2)(x,y) = 1 for yh WG, where y = 0 denotes the upper surface of the crystal.

Third-order nonlinear effects are described by the final two terms in Eq. (5). In writing Eq. (5), we assume that the third-order nonlinear susceptibility can be approximated by χ (3) (ω; ω – Ω, ω′, Ω – ω′) = χE + χR(Ω), where χE is the instantaneous (frequency-independent) electronic susceptibility, and χR(Ω) is the Raman susceptibility (which depends only on the Raman frequency shift Ω). We define χR(Ω) ≡ χR,pkHR(Ω), where χR,pk is the peak Raman susceptibility, and HR(Ω) is the Raman transfer function; HR(Ω) is the Fourier transform of the Raman (temporal) response function h(t). The peak Raman frequency Ωpk is defined as the frequency shift for which |ℑ[χR|] is largest; HR is normalized so that ℑ[HR(±Ωpk)] = ∓1, and the peak Raman susceptibility is defined as χR,pk ≡ |ℑ[χRpk)]|. The complex Raman transfer function is discussed in Appendix A. The values we use for the nonlinear susceptibilities are discussed in Appendix A and in section 4.

Most of the χ (2) interactions modeled by Eq. (5) are highly phase mismatched. In particular, for a given process (SHG, SFG or DFG involving a particular combination of waveguide modes) usually at most one QPM order, denoted m 0, is close to phasematching. To model a different QPM order m, the grid size required increases approximately in proportion to |mm 0|, while the contribution to the pulse dynamics is (approximately) proportional to 1/|mm 0|3. In appendix B, we discuss these terms in more detail and calculate their contribution to the total SPM of the input pulse. In some cases, instead of including higher order terms explicitly in the simulations, their leading-order contributions to the pulse dynamics can be calculated analytically via the cascading approximation, which yields an effective instantaneous χ (3) coefficient for each term. These coefficients can then be added to the true instantaneous χ (3) coefficient χE in Eq. (5), yielding an adjusted and possibly z-dependent effective value of χE (z).

3. Comparison to 1043-nm-pumped experiments

In Figs. 1(a) and 1(b) we show experimental and simulated output spectra corresponding to Fig. 4 of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], respectively. The simulations are in quite good agreement with the experiments. For the 7–11 μm QPM period case, the simulation captures the main features seen experimentally, namely the high power spectral density (PSD) between 1 μm and 1.2 μm, and the “pedestal” generation between 1.4 and 1.75 μm; the reduction in PSD between 1.2 and 1.4 μm in the experiment is also reproduced (although there is a larger reduction in the simulation). For the 7–8 μm QPM period case, the simulated spectrum falls off at approximately the same rate as in the experiment, reaching 10−6 of its peak at 1.45 μm. The simulated spectra were averaged over five simulations with semiclassical quantum noise seeding on the input pulse.

Fig. 1 Output spectrum for the 1043-nm-case of Ref. [1], labeled with the range of periods in the linearly chirped QPM grating. The pulse energy is 3.45 nJ in all cases. (a) Experimentally measured, (b) numerically simulated using Eqs. (5). A QPM grating with a weak linear chirp from 7–8 μm is included for comparison, showing reduced spectral broadening.
Fig. 4 Spectrum for several values of χE, with χR,pk = 0. The values for χE are given in the legend in units of pm2/V2; the other model parameters are the same as those used in Fig. 1(b).

Fig. 2 Pulse evolution in the time-domain for the simulation shown in Fig. 1(b), with QPM period varied from 7–11. (a) The pulse amplitude; the color bar represents |A 0(z,t)|. (b) The phase of the first harmonic part of the pulse (colobar in radians).

This change of sign of the effective χ (3) can be seen in Fig. 2(b), which shows the phase of the pulse after numerically filtering out the second harmonic spectral components. An intensity-dependent phase is accumulated near the start of the QPM grating, but after approximately 7 mm, the rate of phase accumulation changes sign, suggesting that χtotal(3)(z)0 at that point. For comparison, we estimate χtotal(3)(z) analytically with Eq. (16), evaluated for the local grating k-vector Kg(z) and the local frequency of the pulse ωFH(z). This latter quantity is defined by ωFH(z) = [∫ ω|Ã 0(z,ω)|2 ]/[∫ |Ã 0(z,ω)|2 ], where integration is performed only over the first harmonic (FH) region of the pulse. Equation (16) is then evaluated using ωFH(z) for the frequency and Kg(z) for the local grating k-vector. The solid blue lines in Fig. 2 show to the position at which χtotal(3)=0 according to this calculation; this position is close to where the simulated rate of phase accumulation changes sign. Thus, the initial behavior of the pulse in Fig. 2 is described quite well by this simplified picture.

Next, to understand the origin of the spectral components between 1.4 and 1.75 μm, we show in Fig. 3(a) the propagation of the pulse in the frequency domain (plotted on a log scale). The generation of spectral components > 1.4 μm can be seen to occur around 15 mm from the start of the QPM grating. To help explain this and other processes, we show in Fig. 3(b) a simulated cross-FROG spectrogram at the output of the QPM grating, using a 150-fs Gaussian gate pulse. Due to a Raman self frequency shift (SFS) effect [21

21. J. Gordon, “Theory of the soliton self-frequency shift,” Opt. Lett. 11, 662–664 (1986). [CrossRef] [PubMed]

], the FH pulse shifts to lower frequencies and leaves behind a “trail” consisting of spectral components between 275 and 325 THz; simultaneously, a second harmonic (SH) wave is generated around a frequency determined by the spatial dependence of the QPM period, and also by the spatial dependence of the frequency, velocity, and propagation coefficient of the FH pulse. The generation of such waves is expected when there is a significant group index mismatch in addition to a large phase mismatch [19

19. M. Bache, O. Bang, J. Moses, and F. W. Wise, “Nonlocal explanation of stationary and nonstationary regimes in cascaded soliton pulse compression,” Opt. Lett. 32, 2490–2492 (2007). [CrossRef] [PubMed]

]. The group index difference between 521.5 and 1043 nm is ≈0.287 in this case, and the group velocity dispersion (GVD) is positive (and comparable to that of bulk LiNbO3); the phase mismatch is discussed in section 4. In Fig. 3 this SH wave is shown over a limited temporal range (with delay-dependent frequencies of around 460 THz), but it extends over approximately 24 ps (with frequency increasing with delay), corresponding to the relative delay accumulated between the SH and FH frequencies over the length of the waveguide.

Fig. 3 (a) Spectrum versus position in the QPM grating, showing generation of spectral components > 1.4 μm from noise. (b) Simulated cross-FROG spectrogram (150 fs gate), plotted on a dB scale. The reference velocity v ref used in the simulation was the group velocity of the TM00 mode at 990 nm.

Due to the presence of the self frequency shifted FH pulse, its Raman trail, and the generated SH pulse, spectral components between 1.3 and 1.75 μm can be generated by at least two processes. One of these processes is optical parametric amplification (OPA) of quantum noise: spectral components between 1.4 and 1.75 μm are generated via OPA between the SH pulse (which acts as the “pump”) and quantum noise components around the input frequency. The generation of these spectral components from the (semiclassical) noise floor is evident in Fig. 3(a); when quantum noise is turned off in the simulation, these spectral components become much weaker (see Fig. 5). Because of the high SH intensities involved, the OPA process can have high gain, provided that phasematching is satisfied. The gain for this process is quite subtle, however, since the frequency and intensity of the generated SH wave depends on position z via the QPM chirp and Raman SFS of the FH pulse, and hence there is a spatially-dependent “pump” frequency and intensity; additionally, there is a spatially-dependent QPM period, and rapid temporal walk-off between pump, signal and idler spectral components. Thus, this situation differs somewhat from the chirped QPM OPA interactions studied elsewhere [22

22. M. Charbonneau-Lefort, B. Afeyan, and M. M. Fejer, “Optical parametric amplifiers using chirped quasi-phase-matching gratings I: practical design formulas,” J. Opt. Soc. Am. B 25, 463–480 (2008). [CrossRef]

24

24. C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “Ultrabroadband, highly flexible amplifier for ultrashort midinfrared laser pulses based on aperiodically poled Mg:LiNbO3,” Opt. Lett. 35, 2340–2342 (2010). [CrossRef] [PubMed]

], where narrow-bandwidth pumps were assumed. Nonetheless, the general model of Eq. (5) captures this effect. In the spectrogram, the amplified noise can be seen in the crescent-like pattern around the FH which extends out to approximately (6 ps, 330 THz) (for the “signal” components) and (2 ps, 170 THz) (for the “idler” components). In addition to this noise-seeded OPA process, there is a coherent process by which spectral components > 1.3 μm are generated. This process involves the generated SH pulse mixing with the FH pulse and its Raman trail according to phasematched DFG, and can be seen in the weaker “outer” crescent-like patter which extends to approximately (0.8 ps, 150 THz) in Fig. 3(b).

Fig. 5 Spectrum for several values of χR,pk (in pm2/V2) with fixed χE + HR(0)χR,pk = 6.365 × 103 pm2/V2; the other model parameters are the same as those used in Fig. 1(b). For the dashed black line, quantum noise was neglected.

4. Model calibrations for 1043-nm-pumped experiments

In this section, we discuss which effects modeled by Eq. (5) were important in the above simulations, the sensitivity of the simulations to those effects, and how we estimate the values of the χ (3) nonlinear coefficients.

In Fig. 2, we showed that the cascaded χ (2) and χ (3) SPM effects were of opposite sign and comparable magnitude. Here, we will quantify these terms, and hence the initial dynamics, by analyzing the χ (3) and cascaded χ (2) contributions to SPM near the input of the QPM grating. The effective low-frequency-shift third-order nonlinear susceptibility, χtotal(3), is given by Eq. (16). This susceptibility determines the initial rate of SPM for narrow-bandwidth pulses. There are contributions to χtotal(3) from χE, SRS, and from each waveguide mode at each order of the QPM grating via cascaded χ (2) interactions. These cascaded χ (2) contributions are labeled χcascadem,q for QPM order m, SH waveguide mode index q and FH waveguide mode index 0, and are described in appendix B in terms of the phase mismatches Δkm,q 00 [defined in Eq. (14)]. The necessary parameters for calculating χcascadem,q for these experiments are given in Table 2. There are also small but non-negligible nonlinear phase shifts due to cross phase modulation and cross Raman scattering from the SH acting on the FH pulse; these effects are captured by the simulations, but we neglect them for this simplified analysis.

Table 2. Cascading Approximation Parameters at 1043 nm

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As the pulse propagates through the waveguide, the Raman SFS effect causes Δk 0,000 (the phase mismatch for SHG involving the lowest-order waveguide modes and with QPM order 0) to decrease, and hence decreases Δk 1,000 since Δk 0,000 > Kg; however, the QPM chirp increases Δk 1,000 with z since Kg(z) is decreasing. The net result of these two effects for this particular case is that χcascade(3) actually changes sign during propagation, as shown in Fig. 2(b). Since the cancellation between the cascaded χ (2) and χ (3) terms hindered conventional spectral broadening via SPM, the dominant mechanism for generating spectral components > 1.4 μm was optical parametric amplification of quantum noise. This mechanism is consistent with the results of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], where observation of the carrier-envelope-offset frequency f CEO was reported when using a 1580-nm pump, but not for the 1043-nm-pumping case. For many applications, the coherence of the supercontinuum is important; if the cancellation of SPM effects was reduced, the intensity of the SH could be reduced and the rate of FH SPM increased, thereby allowing for coherent supercontinuum generation.

The above analysis of SPM effects depends on the low-frequency χ (3) terms, χE + HR(0)χR,pk, which can be determined quite accurately. However, both χ (3) parameters need to be known or estimated for the simulations; in appendix A, we show that it is difficult to use available literature data to absolutely calibrate both χE and χR,pk simultaneously. To show the importance of these terms beyond the χ (2)χ (3) competition calculated above, we first show that SRS must be included in the model in order to reproduce the experimental results. The importance of SRS is shown in Fig. 4, where we plot the output spectrum for several values of χE while setting χR,pk = 0. For the case with χR,pk = 0 and χE = 0, there is a long-wavelength pedestal which extends to > 2 μm; this pedestal extends further than the spectra shown in Figs. 1(a) and 1(b). However, based on the results of appendix A, χE + HR(0)χR,pk ≈ 6365 pm2/V2. When χE is increased towards this value (while still maintaining the false assumption that χR,pk = 0), the spectral broadening is reduced significantly. For the χE > 3000 pm2/V2 cases in Fig. 4, the bandwidth is much narrower than in the experiments. Furthermore, for each case, there is no spectral “flattening” between 1 and 1.2 μm, and no dip between 1.2 and 1.4 μm; both of these spectral features can be seen in Figs. 1(a) and 1(b).

The results of Fig. 4 show that simulations with SRS neglected differ significantly from the experimental results; the results of section 3 show that when all the terms in Eq. (5) are included, our model is sufficient to reproduce the spectral features observed experimentally, without any adjustments to known model parameters. We can therefore conclude that SRS plays an important role in the dynamics and must be included in the model. Note, in particular, that including SRS leads to the spectral “flattening” shown in Figs. 1(a) and 1(b). The frequency range over which the pulse spectrum is flattened is comparable to the Raman peak with largest frequency shift (≈ 19 THz) (see Fig. 7). Since this flattening also only occurs when SRS is included in the model, we can identify this effect as a Raman SFS.

Fig. 7 Measured imaginary and complex reconstructed stimulated Raman scattering transfer function for e-wave interactions in LiNbO3, based on our XZZ spontaneous Raman scattering measurement. The quality of the fit implies that a sum of Lorentzians is a suitable model for ℑ[HR], so we calculate ℜ[HR] from these fit parameters (dashed red line).

In section 5, we show that χ (3) effects are also important for the 1580-nm-pumped experiments of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], and use the same nonlinear coefficients there as those discussed in this section, with χE scaled by a factor close to theoretical predictions [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

], to accurately model those experiments as well; this agreement suggests that the χ (3) parameters estimated here are realistic. However, due to the χ (2)χ (3) competition, our simulations are sensitive to relatively small deviations in the model parameters and input conditions from our assumptions. In particular, such variations could arise from differences in χ (2) and χ (3) between protonated and congruent LiNbO3, the frequency-dependence of the χ (2) and χ (3) susceptibilities, and the input pulse chirp. Direct measurements of the χ (3) susceptibilities in future work would be of great value for modeling supercontinuum generation in QPM media.

5. Comparison to 1580-nm-pumped experiments

In Fig. 6(a) we show the experimental results for continuum generation from Fig. 3 of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

]; in Fig. 6(b) we show simulations of these experiments using Eqs. (5) for three slightly different values of χE, with the TM00 and TM10 modes and the QPM orders +1, +3, and +5 [index m in Eq. (5)] included in the model, and assuming a 1.2-nJ input pulse. The remaining modes and QPM orders are accounted for via the cascading approximation given by Eq. (16). The simulations capture the spectral broadening observed experimentally, particularly the χE = 4368 pm2/V2 case; the evolution of the spectrum of the pulse for that case is shown in Fig. 6(c). Semiclassical quantum noise on the input pulses was included, as in sections 3 and 4, but did not change the output spectra.

Fig. 6 (a) Experimental data with 1580-nm-pumping from Ref. [1] (b) Simulated output spectrum corresponding to (a) for the TM00, TM10, and TM02 modes. Three slightly different values of χE have been assumed; these values are explained in the text. (c) Evolution of the spectrum of the TM00 mode through the waveguide (dB scale).

The initial dynamics (near the start of the QPM grating) can be described quite accurately via the cascading approximation, discussed in appendix B and applied in section 4. Table 3 gives the relevant parameters for evaluating the cascaded χ (2) terms, χcascadem,q (QPM order m and SH mode q) for SHG of the 1580-nm input pulse. With d 33 = 19.5 pm/V [26

26. I. Shoji, T. Kondo, A. Kitamoto, M. Shirane, and R. Ito, “Absolute scale of second-order nonlinear-optical coefficients,” J. Opt. Soc. Am. B 14, 2268–2294 (1997). [CrossRef]

], Σχcascadem,q=5900pm2/V2. The three largest-magnitude cascaded-χ (2) terms are χcascade1,0=4416pm2/V2, χcascade1,0=1058pm2/V2, and χcascade1,4=473pm2/V2. With the value of χE + HR(0)χR,pk determined in appendix A combined with these cascade contributions, the total effective χ (3) is χtotal(3)=465pm2/V2, of the wrong sign for soliton formation. However, based on a theoretical two-band model, it is predicted that χE should decrease with wavelength [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

]. When such a scaling of χE with frequency is included such that χE (1580 nm) ≈ 4368 pm2/V2, χtotal(3)=626pm2/V2, which then supports soliton formation given the positive GVD of LiNbO3 (and RPE waveguides) at 1580 nm. With this value of χE, the simulations are in good agreement with the experimental results. This scaling corresponds to χE (1580 nm)/χE (1043 nm) = 0.8, which is very close to the predictions from the (oversimplified) two-band model. Note, however, that due to the almost complete cancellation of χ (3) and χ (2) contributions to the total SPM, χ total and hence the output spectrum is sensitive to errors in χE and d 33 of as little as 2%. This sensitivity is illustrated in Fig. 6(b), where the three χE cases shown correspond to values for χE (1580nm)/χE (1043nm) of 0.775, 0.8, and 0.825. For more accurate modeling, it would be necessary to know the model parameters precisely, or to operate in a regime where the χ (2)χ (3) cancellation is not so complete.

Table 3. Cascading Approximation Parameters at 1580 nm

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With the above scaled values of χE, 1580 nm pulse broadens in spectrum and compresses in time due to the negative χtotal(3), positive dispersion, and high intensity. This process is determined by the combined effects of SPM, group velocity dispersion (GVD), group velocity mismatch (GVM), and SRS. Once enough spectral broadening has occurred, various short-wavelength peaks in the spectrum are generated via phasematched processes involving different waveguide modes and QPM orders.

In the experiments, a peak in the PSD occurred around 2.85 μm and was of comparable magnitude and wavelength for different QPM periods, as shown in Fig. 6(a). Similarly, we found that the wavelength of the long-wavelength peak in the simulations is only weakly dependent on QPM period (i.e. a peak occurs at the same wavelength for a relatively wide range of QPM periods, as long as there is sufficient spectral broadening). For conventional DFG of frequency ωDFG = ω 1ω 2, one normally would anticipate Δk = β (ω 1) – β (ω 2) – β (ωDFG) – Kg, and thus a wavelength for the phasematched peak that would depend strongly on the QPM period, in contrast to the observed behavior. However, for a DFG process involving a phase mismatched second harmonic (SH) component as one of the participating waves, the phasematching condition differs from the conventional one.

Consider a FH pulse with center frequency ωFH, effective soliton propagation constant βeff(FH)(ω)β0(ωFH)+(ωωFH)/vFH, and group velocity v FH ≈ (∂β 0(ωFH)/∂ω)−1. The effective propagation coefficient for the phase mismatched SH pulse at frequencies ω in the vicinity of 2ωFH is given approximately by βeff(SH)(ω)2β0(ωFH)+Kg+(ω2ωFH)/vFH. This form arises for a SH pulse π/2 radians out of phase with its driving polarization, aligned temporally with the FH pulse at group velocity vFH. For the DFG process involving such a SH spectral component with frequency ω 2, a FH spectral component with frequency ω 1, and a generated wave at frequency ωDFG = ω 2ω 1, the effective Δk is given by
Δkeff(ωDFG,ω2)=βeff(SH)(ωDFG+ω1)βeff(FH)(ω1)β0(ωDFG)Kg.
(6)

With the approximate forms of the propagation constants βeff(j) described above,
Δkeff(ωDFG)β0(ωFH)+(ωDFGωFH)/vFHβ0(ωDFG),
(7)
independent of Kg, in contrast to the conventional phasematching relation. This type of phase-matching relation was discussed in [27

27. M. Bache, O. Bang, B. B. Zhou, J. Moses, and F. W. Wise, “Optical cherenkov radiation in ultrafast cascaded second-harmonic generation,” Phys. Rev. A 82, 063806 (2010). [CrossRef]

]. There is a dip in the spectra at 2.85 μm: this dip occurs due to OH absorption, which we included in the model as a complex Lorentzian perturbation to the effective index, with a corresponding peak absorption of 3 mm−1.

The wavelength of the above cascaded DFG peak depends sensitively on the waveguide dispersion. With the dispersion relation for the lowest-order mode calculated from our concentration-dependent proton-diffusion model and the dispersion relation for protonated lithium niobate given in Ref. [9

9. R. V. Roussev, “Optical-frequency mixers in periodically poled lithium niobate: materials, modeling and characterization,” Ph.D. thesis, Stanford University (2006), http://nlo.stanford.edu/system/files/dissertations/rostislav_roussev_thesis_december_2006.pdf.

], the wavelength of the cascaded DFG peak is > 3 μm, longer than observed experimentally. However, this waveguide model is calibrated primarily for wavelengths < 2 μm. Furthermore, small changes in fabrication parameters could also lead to shifts in the effective phase mismatch given by Eq. (7). Therefore, some discrepancy with experiments can be expected for processes involving longer wavelengths. The wavelength of the cascaded DFG peak also depends on the pulse frequency, which changes due to the Raman SFS. This SFS depends on the value of χR,pk (which we only estimate via the approach of section 4) and on the input electric field profile (which is not known for these experiments). For Fig. 6(b) we added a small, smooth and monotonic frequency-dependent offset to the effective index in order to shift the cascaded DFG peak to 2.85 μm. We chose a functional form δn(ω)=12[1+tanh((ωLω)/Δω)][(ωωL)/(ωOHωL)]2δn0, where ωL and ωOH corresponding to 2 μm and 2.85 μm, respectively, Δω = 2π × 5 THz, and δn 0 = −8 × 10−4. This functional form was chosen so that δnδn 0 at 2.85 μm, and so that δn ≈ 0 for wavelengths < 2 μm, where our waveguide dispersion model is well-calibrated. This latter constraint helps to ensure that the pulse dynamics are not artificially altered by the effective index offset. Provided the above constraints are met, we have found that the spectrum is relatively insensitive to the functional form of δn(ω). With improved characterization of the RPE waveguide dispersion and the nonlinear parameters of the model, and the input pulse, this offset might not be required.

In comparing this work with prior modeling of the experiments of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], we note that Refs. [8

8. M. Conforti, F. Baronio, and C. De Angelis, “Nonlinear envelope equation for broadband optical pulses in quadratic media,” Phys. Rev. A 81, 053841 (2010). [CrossRef]

,13

13. S. Wabnitz and V. V. Kozlov, “Harmonic and supercontinuum generation in quadratic and cubic nonlinear optical media,” J. Opt. Soc. Am. B27, 1707–1711 (2010).

] also show good agreement between single-envelope simulations and the 1580-nm-pumped experiments we have discussed in this section, despite their neglect of the waveguide mode profiles and Raman nonlinearities. Without both the χE and χR,pk terms, only the cascaded-χ (2) interactions contribute to the total SPM, and so χtotal(3) is much larger. For a plane-wave model, cq = 1 for q = 0 and is zero otherwise (see appendix B), and hence χcascade1,0=5170pm2/V2, from Table 3 and Eq. (16). We found earlier in this section that χtotal(3)626pm2/V2. Therefore, neglecting χ (3) and waveguide modes yields an effective low-frequency χ (3) that is an order of magnitude too large for these particular experiments. If χ (3) and waveguide effects are neglected, the SPM is increased by a similar factor, which will substantially alter the dynamics. For some cases, the measured spectrum can still be recovered by treating the pulse intensity as an unconstrained fit parameter in the simulations, but such an approach would not work in other spectral ranges, such as those considered in section 3.

6. Discussion

We have shown that our nonlinear waveguide model, given in Eq. (5), is in good agreement with the experimental results of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

] with both 1580 nm and 1043 nm pumps, and both uniform and chirped QPM gratings. The χ (2), instantaneous χ (3), and stimulated Raman scattering nonlinearities are all essential for accurately modeling the supercontinuum generation process.

Despite the difficulties in fully calibrating the parameters entering into Eqs. (5), our model is sufficiently accurate to be used to analyze and design QPM gratings and waveguides in order to improve spectral broadening or to perform other ultrafast functionalities such as nonlinear pulse compression. For example, we have shown that for both of the experiments of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], the cascaded-χ (2) and χ (3) susceptibilities were of comparable magnitude but opposite sign, and hence competed with each other. This cancellation of contributions to χtotal(3) significantly reduces the rate of SPM and hence increases the energy requirements for supercontinuum generation. Furthermore, based on our simulations, in the 1043-nm-pumped case the reduced rate of SPM meant that a strong FH pulse had to be used, which in turn led to a strong generated SH pulse; this SH pulse amplified quantum noise, which led to incoherent supercontinuum generation.

In fiber-based supercontinuum generation, the zero group velocity dispersion (GVD) wavelength, λ GVD, is often shifted by waveguide design to be nearby the input wavelength, λ. We have seen from simulations based on Eqs. (5) that shifting λ GVD to be comparable to the input wavelength is one way to significantly enhance the spectral broadening processes for χ (2)-based continua as well, since generated spectral components would then remain overlapped temporally. Furthermore, if the GVD at the input wavelength is negative (λ > λGVD) or negligible (λλGVD), supercontinuum generation could be achieved with a positive χtotal(3); for this case, by choosing Δk 1,000 < 0 the contributions of χ (2) and χ (3) to SPM would add rather than cancel while still being of appropriate sign to support χ (3)-like bright solitons at the input wavelength.

Since RPE waveguides are weakly guiding, significant shifts in λ GVD cannot be obtained. However, λ GVD is conveniently located near 2 μm, making Tm-based laser sources promising candidates for increased spectral broadening in RPE waveguides [28

28. C. R. Phillips, J. Jiang, C. Langrock, M. M. Fejer, and M. E. Fermann, “Self-Referenced Frequency Comb From a Tm-fiber Amplifier via PPLN Waveguide Supercontinuum Generation,” in CLEO:2011 - Laser Applications to Photonic Applications, OSA Technical Digest (CD) (Optical Society of America, 2011), paper PDPA5.

]. An alternative approach is to use tightly confining waveguides, in which a high index contrast could enable shifting λ GVD to the 1.55-μm and 1-μm spectral regions. There is also the possibility of using AlGaAs QPM waveguides, which can be tightly confining and are transparent in the mid-IR. With simultaneous engineering of the waveguide dispersion and the QPM grating, supercontinuum generation may be possible across the mid-IR. With the model of nonlinear interactions in QPM waveguides we have developed here, strategies can be developed for reaching spectral regions not accessible to silica-fiber-based supercontinuum sources, and for performing optimizations made possible by the versatility of QPM gratings [14

14. G. Imeshev, M. M. Fejer, A. Galvanauskas, and D. Harter, “Pulse shaping by difference-frequency mixing with quasi-phase-matching gratings,” J. Opt. Soc. Am. B 18, 534–539 (2001). [CrossRef]

, 23

23. C. R. Phillips and M. M. Fejer, “Efficiency and phase of optical parametric amplification in chirped quasi-phase-matched gratings,” Opt. Lett. 35, 3093–3095 (2010). [CrossRef] [PubMed]

], suggesting a path towards compact and robust traveling-wave frequency comb sources in the IR and mid-IR spectral regions.

A. Material properties

In this appendix, we discuss the nonlinear coefficients contained in Eqs. (5); these must be known accurately in order to quantitatively predict experimental results. For the second-order nonlinear terms in this appendix, we assume d 33 = 25.2 pm/V for 1064-nm-SHG [26

26. I. Shoji, T. Kondo, A. Kitamoto, M. Shirane, and R. Ito, “Absolute scale of second-order nonlinear-optical coefficients,” J. Opt. Soc. Am. B 14, 2268–2294 (1997). [CrossRef]

]; for modeling the experiments of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], we assume d 33 = 25.2 pm/V for 1043-nm-pumping and d 33 = 19.5 pm/V for 1580-nm-pumping [26

26. I. Shoji, T. Kondo, A. Kitamoto, M. Shirane, and R. Ito, “Absolute scale of second-order nonlinear-optical coefficients,” J. Opt. Soc. Am. B 14, 2268–2294 (1997). [CrossRef]

]. We also assume that χ (2) (x,y) = 0 close to the upper surface of the crystal, as described in section 2.

We determined the imaginary part of the normalized Raman transfer function, ℑ[HR], by measuring the spontaneous Raman scattering cross section of congruent LiNbO3 (CLN). The Raman spectrum of LiNbO3 has been measured previously [29

29. A. S. Barker and R. Loudon, “Dielectric properties and optical phonons in LiNbO3,” Phys. Rev. 158, 433 (1967). [CrossRef]

, 30

30. P. J. Delfyett, R. Dorsinville, and R. R. Alfano, “Spectral and temporal measurements of the third-order nonlinear susceptibility of LiNbO3 using picosecond Raman-induce phase-conjugate spectroscopy,” Phys. Rev. B 40, 1885 (1989). [CrossRef]

], but in some cases large relative errors have been reported [29

29. A. S. Barker and R. Loudon, “Dielectric properties and optical phonons in LiNbO3,” Phys. Rev. 158, 433 (1967). [CrossRef]

]; an additional measurement across the Raman spectrum could prove to be useful. For the measurement, we used a WiTec Alpha300 S Raman microscope in the XZZ configuration and at 295 K. The resulting spectrum is shown in Fig. 7 (blue line). To determine ℜ[HR] we fitted the measured ℑ[HR] to a sum of Lorentzians; this reconstructed Raman susceptibility is also shown in Fig. 7. The parameters for the fit are given in Table 1. The terms of the Lorentzian fit have form aj/(fj2f2+2iγjf) for optical frequency f (not angular frequency). The small measured peak at around −4.6 THz (≈ −153 cm−1) was neglected in the fit since it does not correspond to the e-wave polarization component [31

31. N. Surovtsev, V. Malinovskii, A. Pugachev, and A. Shebanin, “The nature of low-frequency raman scattering in congruent melting crystals of lithium niobate,” Phys. Solid State 45, 534–541 (2003). [CrossRef]

]: its presence indicates imperfect polarization discrimination during the measurement.

Table 1. Lorentzian Fit Parameters for HR(Ω)

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In Ref. [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

], n 2 was measured for LiNbO3 with the Z-scan method using 30-ps (FWHM) pulses with a center wavelength of 1.064 μm. Since the reconstructed LiNbO3 Raman spectrum in Fig. 7 does not vary significantly for frequencies < 1 THz, and assuming the 30-ps pulses had bandwidths ≤1 THz, the measured n 2 represented the total low-frequency-shift χ (3) response, given by χ0(3)=χE+χR,pkHR(0)+χcascade(3). This total response can be calculated from n 2 as
χ0(3)=43n2nFH2ɛ0c,
(8)
which in LiNbO3 is given by χ0(3)1575pm2/V2 for n 2 = 0.933 × 10−6 cm2/GW [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

]. The contribution to χ0(3) from the cascaded χ (2) interaction is given, via multiple scale analysis [34

34. C. Conti, S. Trillo, P. Di Trapani, J. Kilius, A. Bramati, S. Minardi, W. Chinaglia, and G. Valiulis, “Effective lensing effects in parametric frequency conversion,” J. Opt. Soc. Am. B 19, 852–859 (2002). [CrossRef]

], by
χcascade(3)=16πdeff23nSHλFH1Δk.
(9)

For the Z-scan configuration used for LiNbO3 in Ref. [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

], d eff = d 33 and Δk = k(2ω) – 2k(ω) = 0.927 μm−1 for 1.064-μm pumping [35

35. D. H. Jundt, “Temperature-dependent Sellmeier equation for the index of refraction, ne, in congruent lithium niobate,” Opt. Lett. 22, 1553–1555 (1997). [CrossRef]

], and so χcascade(3)4830pm2/V2. Importantly, χcascade(3) is of opposite sign to the electronic and Raman contributions. After subtracting the contribution of the cascading term (which can be characterized more easily since d 33 is relatively well-known), we find that χE + χR,pkHR(0) = 6365 pm2/V2, much larger than χ0(3).

Additional measurements are needed to determine both χE and χR,pk. One way to determine χR,pk separately from χE is by measuring the SRS gain at Ωpk. However, tabulated values for the Raman gain coefficient are inconsistent with the n 2 measurements described above. Consider the Raman gain coefficient, (gS/IL), evaluated at Stokes frequency ωS = ωL – Ωpk for some pump laser frequency ωL. It can be shown from Eq. (5) that
χR,pk=nSnLɛ0cλS3πgSIL,
(10)
where the reduction in gain due to Stokes-anti-Stokes coupling effects are neglected when defining (gS/IL), and nS and nL are the refractive indices at the Stokes and pump wavelengths, respectively. In LiNbO3 with a 1.064-μm pump laser, the gain coefficient corresponding to a contribution of 100% from SRS to the value of χ0(3)χcascade(3)=6365pm2/V2 calculated above is given by (gS/IL)max ≈ 2.51 cm/GW, using HR(0) = 0.17 from Fig. 7. However, the gain coefficient tabulated in Ref. [36

36. R. Boyd, “Stimulated Raman scattering and stimulated Rayleigh-Wing scattering,” in “Nonlinear Optics”, R. Boyd (Academic, 2008). [CrossRef]

] is 9.4 cm/GW at 694 nm [36

36. R. Boyd, “Stimulated Raman scattering and stimulated Rayleigh-Wing scattering,” in “Nonlinear Optics”, R. Boyd (Academic, 2008). [CrossRef]

], which corresponds to 6.1 cm/GW at 1.064 μm (by scaling with optical frequency); this value is more than twice as large as the upper bound for (gS/IL) provided by the n 2 measurement (assuming χE > 0).

The tabulated values for (gS/IL) are also inconsistent with recent experiments with intense IR pulses. The Raman gain rate gR, in cm−1, can be approximated as [36

36. R. Boyd, “Stimulated Raman scattering and stimulated Rayleigh-Wing scattering,” in “Nonlinear Optics”, R. Boyd (Academic, 2008). [CrossRef]

]
gR2[igSΔkR/2(ΔkR/2)2],
(11)
where gS is the coupling coefficient between the intensity at the Stokes shifted wave at ωS and the pump laser at ωL = ωS + ωR,pk, and the phase mismatch ΔkR = 2k(ωL) – k(ωS) – k(ωAS) for anti-Stokes frequency ωAS. In turn, gS is determined via the Raman gain coefficient (gS/IL). In Refs. [24

24. C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “Ultrabroadband, highly flexible amplifier for ultrashort midinfrared laser pulses based on aperiodically poled Mg:LiNbO3,” Opt. Lett. 35, 2340–2342 (2010). [CrossRef] [PubMed]

, 37

37. C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “High-power mid-infrared optical parametric chirped-pulse amplifier based on aperiodically poled Mg:LiNbO3,” presented at the Conference on Lasers and Electro-optics (2011).

], a 1-cm-long MgO:LiNbO3 crystal was pumped with 1.064-μm pulses with intensities of > 7 GW/cm2 and durations of 12 ps (FWHM); the corresponding pulse bandwidth was significantly narrower than the linewidth of the main peaks of HR(Ω) shown in Fig. 7. Assuming (gS/IL) = 6.1 cm/GW, Eq. (11) predicts an SRS gain in this case of 136 dB (185 dB if ΔkR → ∝). Despite this high predicted gain, no Stokes wave was observed experimentally. If we instead assume the n 2-based upper bound on (gS/IL) of 2.51 cm/GW, the gain at the Stokes frequency would be approximately 69 dB.

Based on the above considerations, it is difficult to use available literature data to absolutely calibrate the third-order nonlinear coefficients in LiNbO3 (which could also be different from those of RPE LiNbO3). The peak Raman susceptibility χR,pk can be bounded above by the non-linear refractive index and by the absence of SRS in the high intensity experiments discussed, and χR,pk HR(0) + χE can be estimated from the measured value of n 2. For this paper, we further constrain the susceptibilities to yield output spectra in quantitative agreement with the super-continuum generation experiments of Ref. [1

1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

], which we model in Sections 5 and 3. The χ (3) values that we use are given by χR,pk = 5.3 × 103 pm2/V2, and χE (1043 nm) = 5.46 × 103 pm2/V2; these parameters are discussed further in section 4. For the 1580-nm-pumping case discussed in section 5, χE is scaled according to theoretical predictions [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

], such that χE (1580 nm)/χE (1043 nm) ≈ 0.8.

Lastly, note that in Ref. [25

25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

] there was a significant discrepancy between measurements of n 2 and a theoretical calculation based on a simplified two-band model: at 532 nm, the measured n 2 was approximately 9.1× the value at 1064 nm, while the two-band theory predicted a scaling factor of only 2.5. A possible resolution to this discrepancy is that there is a large negative contribution to n 2 from χcascade(3) in the 1064-nm case but not in the 532-nm case. Since 532 nm lies above half the LiNbO3 bandgap, it has a very large phase mismatch with its second harmonic at 266 nm [38

38. E. D. Palik and G. Ghosh, Handbook of Optical Constants of Solids (Academic, 1985).

], so the contribution to n 2 from χcascade(3) is much smaller. Given the scaling of χE with frequency, the additional 532-nm data point for n 2 is sufficient, in principle, to determine both χE and χR,pk at 1.064 μm. With the measured value of n 2 = 8.25 × 10−6 cm2/GW, χ0(3)(532nm)=14.6×103pm2/V2. If we assume that χcascade(3)(532nm)=0 and that χR,pk = 5.3 × 103 pm2/V2 is non-dispersive, then χE (532 nm) = 13.4 × 103 pm2/V2, and χE (532 nm)/χE (1064 nm) ≈ 2.5, in good agreement with the (oversimplified) two-band model.

B. Cascading approximation for QPM waveguides

In the limit of a large phase mismatch, χ (2) interactions can be approximated by χ (3)-like self-and cross-phase-modulation (SPM and XPM) terms; this approach is termed the cascading approximation, and has been discussed extensively [33

33. G. I. Stegeman, D. J. Hagan, and L. Torner, “χ(2) cascading phenomena and their applications to all-optical signal processing, mode-locking, pulse compression and solitons,” Opt. Quantum Electron. 28, 1691–1740 (1996). [CrossRef]

]. In Appendix A, we used this approximation to constrain the χ (3) susceptibilities given a known nonlinear refractive index. In this appendix, we determine the cascading approximation for the case of QPM waveguide interactions. This calculation gives a total effective χ (3), denoted χtotal(3), which determines the rate of SPM for narrow-bandwidth pulses. In the experiments modeled in sections 3 and 5 the conditions for the validity of the approximation are not always satisfied for all of the waveguide modes and QPM orders, but it nonetheless provides valuable insight into the pulse propagation dynamics (especially near the start of the QPM grating),

In Ref. [34

34. C. Conti, S. Trillo, P. Di Trapani, J. Kilius, A. Bramati, S. Minardi, W. Chinaglia, and G. Valiulis, “Effective lensing effects in parametric frequency conversion,” J. Opt. Soc. Am. B 19, 852–859 (2002). [CrossRef]

], the cascading approximation was determined from coupled wave equations for SHG via a multiple-scale analysis. A similar procedure can be used to determine the cascading approximation for QPM waveguides based on Eq. (5). To proceed with the analysis, we first split each mode envelope An into first harmonic (FH) and second harmonic (SH) pulse components An(FH) and An(SH), with carrier frequencies ωFH and ωSH = 2ωFH, respectively. In principle, Eq. (5) can give rise to pulses around carrier frequencies FH for all positive integers m. However, for pulses with a bandwidth ΔΩ less than an octave, i.e. for ΔΩ ≪ ωFH, often only the components around ωFH and ωSH are relevant, to lowest order in the perturbation. Furthermore, higher order modes around ωFH can often be neglected, for example in the case when the waveguide only supports a single mode at that frequency. With these assumptions, Eq. (5) yields the following simplified time-domain coupled wave equations for SHG,
A0,FHz+D^0,FHA0,FH=i(ω2ug0β0c2)ωFH×[q,mdmθq00expiΔkm,q00(z')dz'A0,FH*Aq,SH+3(χE+HR(0)χR,pk)8θ0000|A0,FH|2A0,FH]Aq,SHz+D^q,SHAq,SH=i(ω2ugqβqc2)ωSH×[mdmθq00expiΔkm,q00(z')dz'A0,FH22+6(χE+HR(0)χR,pk)8θqq00|A0,FH|2Aq,SH]
(12)
where the FH and SH envelopes are given in terms of the following approximate form for the spectrum of the total electric field
E˜(x,y,z,ω)12B˜0(x,y,ωFH)A0,FH(z,ωωFH)expi(β0(ωFH)ωFH/vref)+12qB˜q(x,y,ωSH)Aq,SH(z,ωωSH)expi(βq(ωSH)ωSH/vref)
(13)
and θq 00 = Θq 00(ωSH,ωFH) and θqq 00 = Θqq 00(ωSH, ωSH, ωFH). In Eqs. (12) and (13), the spatial mode profiles and coupling coefficients have been evaluated at the optical carrier frequencies. We have also assumed that the intensity of the SH pulse is much lower than that of the FH pulse, and have therefore neglected the χ (3) terms involving |An,SH|2. For the purposes of this simplified analysis, we have assumed that the pulse bandwidth is narrow enough that HR(Ω) can be approximated as HR(0); this approximation does not apply for supercontinuum generation [in the simulations, we use HR(Ω)], but is useful for estimating the rate of SPM for the FH pulse at the start of the QPM grating. The phase mismatch terms are given by
Δkm,q00(z)=βq(2ωFH)2β0(ωFH)mKg(z)
(14)
for QPM order m and waveguide mode q of the SH pulse. If the characteristic length defined by L m1,m2 ≡ |Δk m1,q00 – Δk m2,q00|−1 is much shorter than any other characteristic lengths of the problem for all m 1m 2, the multiple-scale analysis can be applied. In Ref. [34

34. C. Conti, S. Trillo, P. Di Trapani, J. Kilius, A. Bramati, S. Minardi, W. Chinaglia, and G. Valiulis, “Effective lensing effects in parametric frequency conversion,” J. Opt. Soc. Am. B 19, 852–859 (2002). [CrossRef]

], where the linear operators represented diffracting beams, the Rayleigh range would be a relevant characteristic length. For pulses, the group velocity mismatch length between the FH and SH pulses is one of several important characteristic length scales. Given a sufficiently small value of L m1,m2 for all m 1m 2, and assuming that there is no SH pulse input at the start of the interaction, multiple scale analysis of Eq. (12) yields
A0,FHz+D^0,FHA0,FH=iωFHg0,FHn0,FHc(m,qdm2θq002ωFHgq,SHnq,SHc1Δkm,q00(z))|A0,FH|2A0,FHiωFHg0,FHn0,FHc(3θ0000(χE+HR(0)χR,pk)8)|A0,FH|2A0,FH,
(15)
where nq,j = βq(ωj)c/ωj and gn,j = gn(ωj) for wave j (j = FH or j = SH) and mode normalization coefficient gn(ω) given by Eq. (4). To analyze the different terms, it is convenient to introduce a simpler normalization of the mode profiles than the one used in sections 3 and 5 to analyze broadband pulses. If we choose gn(ω) = an(ω) instead of gn(ω) = an(ω)1/3, the mode profiles Bn(x, y,ω) are dimensionless. With this definition of gn, θ 0000/g 0(ωFH) = 1, and the total effective χ (3) is given by
χtotal(3)(z)=χE+χR,pkHR(0)m,qcq16πdm23nq,SHλFH1Δkm,q00(z)χE+χR,pkHR(0)+m,qχcascade(m,q)(z)
(16)
where the cascading reduction factors cq are given by
cq=(χ¯BFH2Bq,SHdxdy)2(|BFH4|dxdy)(|Bq,SH|2dxdy),
(17)
independent of the choice of normalization of the spatial mode profiles Bj; χ¯(x,y) is the transverse spatial profile of the second order nonlinear susceptibility, which appears in Eq. (3). The χ (2) contributions to χtotal(3) have forms analogous to Eq. (9). In sections 3 and 5 the values for these terms are discussed, and we show that almost complete cancellation of χtotal(3) can occur. The parameters cq and Δk 0, q 00 are given in Tables 2 and 3 for the 1043-nm and 1580-nm pumped calculations discussed in sections 4 and 5, respectively.

For bulk interactions, χcascade(3) can be found by taking cq = δq 0 and dm = δm 0 d 33 where δij is the Kronecker delta. To calculate the cascaded phase shifts, we assumed a pulse centered around a particular carrier frequency. However, some care should be taken with this procedure since during the supercontinuum generation process the center frequency of the pulse can shift. This frequency shift can reduce the accuracy of the cascading approximation for terms that are nearly phasematched. For the simulations performed here, we add Σχcascadem,q to the instantaneous third-order nonlinear coefficient χE, with summation performed over all terms except those which are either explicitly included in Eq. (5) or for which m = 1.

Acknowledgments

This research was supported by the U.S. Air Force Office of Scientific Research (AFOSR) under grants FA9550-09-1-0233 and FA9550-05-1-0180.

References and links

1.

C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett. 32, 2478–2480 (2007). [CrossRef] [PubMed]

2.

T. Fuji, J. Rauschenberger, A. Apolonski, V. S. Yakovlev, G. Tempea, T. Udem, C. Gohle, T. W. Hänsch, W. Lehnert, M. Scherer, and F. Krausz, “Monolithic carrier-envelope phase-stabilization scheme,” Opt. Lett. 30, 332–334 (2005). [CrossRef] [PubMed]

3.

J. M. Dudley, G. Genty, and S. Coen, “Supercontinuum generation in photonic crystal fiber,” Rev. Mod. Phys. 78, 1135–1184 (2006). [CrossRef]

4.

J. Price, T. Monro, H. Ebendorff-Heidepriem, F. Poletti, P. Horak, V. Finazzi, J. Leong, P. Petropoulos, J. Flanagan, G. Brambilla, X. Feng, and D. Richardson, “Mid-IR supercontinuum generation from nonsilica microstructured optical fibers,” IEEE J. Sel. Top. Quantum Electron. 13, 738–749 (2007). [CrossRef]

5.

C. R. Phillips and M. M. Fejer, “Stability of the singly resonant optical parametric oscillator,” J. Opt. Soc. Am. B 27, 2687–2699 (2010). [CrossRef]

6.

C. Langrock, S. Kumar, J. McGeehan, A. Willner, and M. M. Fejer, “All-optical signal processing using χ(2) nonlinearities in guided-wave devices,” J. Lightwave Technol. 24, 2579–2592 (2006). [CrossRef]

7.

X. Yu, L. Scaccabarozzi, J. S. Harris, P. S. Kuo, and M. M. Fejer, “Efficient continuous wave second harmonic generation pumped at 1.55 μm in quasi-phase-matched AlGaAs waveguides,” Opt. Express 13, 10742–10748 (2005). [CrossRef] [PubMed]

8.

M. Conforti, F. Baronio, and C. De Angelis, “Nonlinear envelope equation for broadband optical pulses in quadratic media,” Phys. Rev. A 81, 053841 (2010). [CrossRef]

9.

R. V. Roussev, “Optical-frequency mixers in periodically poled lithium niobate: materials, modeling and characterization,” Ph.D. thesis, Stanford University (2006), http://nlo.stanford.edu/system/files/dissertations/rostislav_roussev_thesis_december_2006.pdf.

10.

M. Kolesik and J. V. Moloney, “Nonlinear optical pulse propagation simulation: from Maxwell’s to unidirectional equations,” Phys. Rev. E 70, 036604 (2004). [CrossRef]

11.

G. Genty, P. Kinsler, B. Kibler, and J. M. Dudley, “Nonlinear envelope equation modelling of sub-cycle dynamics and harmonic generation in nonlinear waveguides,” Opt. Express 15, 5382–5387 (2007). [CrossRef] [PubMed]

12.

M. Conforti, F. Baronio, and C. De Angelis, “Ultrabroadband optical phenomena in quadratic nonlinear media,” IEEE Photon. J. 2, 600–610 (2010). [CrossRef]

13.

S. Wabnitz and V. V. Kozlov, “Harmonic and supercontinuum generation in quadratic and cubic nonlinear optical media,” J. Opt. Soc. Am. B27, 1707–1711 (2010).

14.

G. Imeshev, M. M. Fejer, A. Galvanauskas, and D. Harter, “Pulse shaping by difference-frequency mixing with quasi-phase-matching gratings,” J. Opt. Soc. Am. B 18, 534–539 (2001). [CrossRef]

15.

M. M. Fejer, G. A. Magel, D. H. Jundt, and R. L. Byer, “Quasi-phase-matched second harmonic generation: tuning and tolerances,” IEEE J. Quantum Electron. 28, 2631–2654 (1992). [CrossRef]

16.

X. Liu, L. Qian, and F. W. Wise, “High-energy pulse compression by use of negative phase shifts produced by the cascade χ(2) : χ(2) nonlinearity,” Opt. Lett. 24, 1777–1779 (1999). [CrossRef]

17.

J. Moses and F. W. Wise, “Soliton compression in quadratic media: high-energy few-cycle pulses with a frequency-doubling crystal,” Opt. Lett. 31, 1881–1883 (2006). [CrossRef] [PubMed]

18.

S. Ashihara, J. Nishina, T. Shimura, and K. Kuroda, “Soliton compression of femtosecond pulses in quadratic media,” J. Opt. Soc. Am. B19, 2505–2510 (2002).

19.

M. Bache, O. Bang, J. Moses, and F. W. Wise, “Nonlocal explanation of stationary and nonstationary regimes in cascaded soliton pulse compression,” Opt. Lett. 32, 2490–2492 (2007). [CrossRef] [PubMed]

20.

M. Bache and F. W. Wise, “Type-I cascaded quadratic soliton compression in lithium niobate: compressing femtosecond pulses from high-power fiber lasers,” Phys. Rev. A 81, 053815 (2010). [CrossRef]

21.

J. Gordon, “Theory of the soliton self-frequency shift,” Opt. Lett. 11, 662–664 (1986). [CrossRef] [PubMed]

22.

M. Charbonneau-Lefort, B. Afeyan, and M. M. Fejer, “Optical parametric amplifiers using chirped quasi-phase-matching gratings I: practical design formulas,” J. Opt. Soc. Am. B 25, 463–480 (2008). [CrossRef]

23.

C. R. Phillips and M. M. Fejer, “Efficiency and phase of optical parametric amplification in chirped quasi-phase-matched gratings,” Opt. Lett. 35, 3093–3095 (2010). [CrossRef] [PubMed]

24.

C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “Ultrabroadband, highly flexible amplifier for ultrashort midinfrared laser pulses based on aperiodically poled Mg:LiNbO3,” Opt. Lett. 35, 2340–2342 (2010). [CrossRef] [PubMed]

25.

R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron. 32, 1324–1333 (1996). [CrossRef]

26.

I. Shoji, T. Kondo, A. Kitamoto, M. Shirane, and R. Ito, “Absolute scale of second-order nonlinear-optical coefficients,” J. Opt. Soc. Am. B 14, 2268–2294 (1997). [CrossRef]

27.

M. Bache, O. Bang, B. B. Zhou, J. Moses, and F. W. Wise, “Optical cherenkov radiation in ultrafast cascaded second-harmonic generation,” Phys. Rev. A 82, 063806 (2010). [CrossRef]

28.

C. R. Phillips, J. Jiang, C. Langrock, M. M. Fejer, and M. E. Fermann, “Self-Referenced Frequency Comb From a Tm-fiber Amplifier via PPLN Waveguide Supercontinuum Generation,” in CLEO:2011 - Laser Applications to Photonic Applications, OSA Technical Digest (CD) (Optical Society of America, 2011), paper PDPA5.

29.

A. S. Barker and R. Loudon, “Dielectric properties and optical phonons in LiNbO3,” Phys. Rev. 158, 433 (1967). [CrossRef]

30.

P. J. Delfyett, R. Dorsinville, and R. R. Alfano, “Spectral and temporal measurements of the third-order nonlinear susceptibility of LiNbO3 using picosecond Raman-induce phase-conjugate spectroscopy,” Phys. Rev. B 40, 1885 (1989). [CrossRef]

31.

N. Surovtsev, V. Malinovskii, A. Pugachev, and A. Shebanin, “The nature of low-frequency raman scattering in congruent melting crystals of lithium niobate,” Phys. Solid State 45, 534–541 (2003). [CrossRef]

32.

R. Schiek, R. Stegeman, and G. I. Stegeman, “Measurement of third-order nonlinear susceptibility tensor elements in lithium niobate,” in “Frontiers in Optics,” (Optical Society of America, 2005), p. JWA74.

33.

G. I. Stegeman, D. J. Hagan, and L. Torner, “χ(2) cascading phenomena and their applications to all-optical signal processing, mode-locking, pulse compression and solitons,” Opt. Quantum Electron. 28, 1691–1740 (1996). [CrossRef]

34.

C. Conti, S. Trillo, P. Di Trapani, J. Kilius, A. Bramati, S. Minardi, W. Chinaglia, and G. Valiulis, “Effective lensing effects in parametric frequency conversion,” J. Opt. Soc. Am. B 19, 852–859 (2002). [CrossRef]

35.

D. H. Jundt, “Temperature-dependent Sellmeier equation for the index of refraction, ne, in congruent lithium niobate,” Opt. Lett. 22, 1553–1555 (1997). [CrossRef]

36.

R. Boyd, “Stimulated Raman scattering and stimulated Rayleigh-Wing scattering,” in “Nonlinear Optics”, R. Boyd (Academic, 2008). [CrossRef]

37.

C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “High-power mid-infrared optical parametric chirped-pulse amplifier based on aperiodically poled Mg:LiNbO3,” presented at the Conference on Lasers and Electro-optics (2011).

38.

E. D. Palik and G. Ghosh, Handbook of Optical Constants of Solids (Academic, 1985).

OCIS Codes
(190.4410) Nonlinear optics : Nonlinear optics, parametric processes
(320.6629) Ultrafast optics : Supercontinuum generation

ToC Category:
Ultrafast Optics

History
Original Manuscript: May 23, 2011
Revised Manuscript: July 19, 2011
Manuscript Accepted: July 19, 2011
Published: September 12, 2011

Citation
C. R. Phillips, Carsten Langrock, J. S. Pelc, M. M. Fejer, I. Hartl, and Martin E. Fermann, "Supercontinuum generation in quasi-phasematched waveguides," Opt. Express 19, 18754-18773 (2011)
http://www.opticsinfobase.org/oe/abstract.cfm?URI=oe-19-20-18754


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References

  1. C. Langrock, M. M. Fejer, I. Hartl, and M. E. Fermann, “Generation of octave-spanning spectra inside reverse-proton-exchanged periodically poled lithium niobate waveguides,” Opt. Lett.32, 2478–2480 (2007). [CrossRef] [PubMed]
  2. T. Fuji, J. Rauschenberger, A. Apolonski, V. S. Yakovlev, G. Tempea, T. Udem, C. Gohle, T. W. Hänsch, W. Lehnert, M. Scherer, and F. Krausz, “Monolithic carrier-envelope phase-stabilization scheme,” Opt. Lett.30, 332–334 (2005). [CrossRef] [PubMed]
  3. J. M. Dudley, G. Genty, and S. Coen, “Supercontinuum generation in photonic crystal fiber,” Rev. Mod. Phys.78, 1135–1184 (2006). [CrossRef]
  4. J. Price, T. Monro, H. Ebendorff-Heidepriem, F. Poletti, P. Horak, V. Finazzi, J. Leong, P. Petropoulos, J. Flanagan, G. Brambilla, X. Feng, and D. Richardson, “Mid-IR supercontinuum generation from nonsilica microstructured optical fibers,” IEEE J. Sel. Top. Quantum Electron.13, 738–749 (2007). [CrossRef]
  5. C. R. Phillips and M. M. Fejer, “Stability of the singly resonant optical parametric oscillator,” J. Opt. Soc. Am. B27, 2687–2699 (2010). [CrossRef]
  6. C. Langrock, S. Kumar, J. McGeehan, A. Willner, and M. M. Fejer, “All-optical signal processing using χ(2) nonlinearities in guided-wave devices,” J. Lightwave Technol.24, 2579–2592 (2006). [CrossRef]
  7. X. Yu, L. Scaccabarozzi, J. S. Harris, P. S. Kuo, and M. M. Fejer, “Efficient continuous wave second harmonic generation pumped at 1.55 μm in quasi-phase-matched AlGaAs waveguides,” Opt. Express13, 10742–10748 (2005). [CrossRef] [PubMed]
  8. M. Conforti, F. Baronio, and C. De Angelis, “Nonlinear envelope equation for broadband optical pulses in quadratic media,” Phys. Rev. A81, 053841 (2010). [CrossRef]
  9. R. V. Roussev, “Optical-frequency mixers in periodically poled lithium niobate: materials, modeling and characterization,” Ph.D. thesis, Stanford University (2006), http://nlo.stanford.edu/system/files/dissertations/rostislav_roussev_thesis_december_2006.pdf .
  10. M. Kolesik and J. V. Moloney, “Nonlinear optical pulse propagation simulation: from Maxwell’s to unidirectional equations,” Phys. Rev. E70, 036604 (2004). [CrossRef]
  11. G. Genty, P. Kinsler, B. Kibler, and J. M. Dudley, “Nonlinear envelope equation modelling of sub-cycle dynamics and harmonic generation in nonlinear waveguides,” Opt. Express15, 5382–5387 (2007). [CrossRef] [PubMed]
  12. M. Conforti, F. Baronio, and C. De Angelis, “Ultrabroadband optical phenomena in quadratic nonlinear media,” IEEE Photon. J.2, 600–610 (2010). [CrossRef]
  13. S. Wabnitz and V. V. Kozlov, “Harmonic and supercontinuum generation in quadratic and cubic nonlinear optical media,” J. Opt. Soc. Am.B27, 1707–1711 (2010).
  14. G. Imeshev, M. M. Fejer, A. Galvanauskas, and D. Harter, “Pulse shaping by difference-frequency mixing with quasi-phase-matching gratings,” J. Opt. Soc. Am. B18, 534–539 (2001). [CrossRef]
  15. M. M. Fejer, G. A. Magel, D. H. Jundt, and R. L. Byer, “Quasi-phase-matched second harmonic generation: tuning and tolerances,” IEEE J. Quantum Electron.28, 2631–2654 (1992). [CrossRef]
  16. X. Liu, L. Qian, and F. W. Wise, “High-energy pulse compression by use of negative phase shifts produced by the cascade χ(2) : χ(2) nonlinearity,” Opt. Lett.24, 1777–1779 (1999). [CrossRef]
  17. J. Moses and F. W. Wise, “Soliton compression in quadratic media: high-energy few-cycle pulses with a frequency-doubling crystal,” Opt. Lett.31, 1881–1883 (2006). [CrossRef] [PubMed]
  18. S. Ashihara, J. Nishina, T. Shimura, and K. Kuroda, “Soliton compression of femtosecond pulses in quadratic media,” J. Opt. Soc. Am.B19, 2505–2510 (2002).
  19. M. Bache, O. Bang, J. Moses, and F. W. Wise, “Nonlocal explanation of stationary and nonstationary regimes in cascaded soliton pulse compression,” Opt. Lett.32, 2490–2492 (2007). [CrossRef] [PubMed]
  20. M. Bache and F. W. Wise, “Type-I cascaded quadratic soliton compression in lithium niobate: compressing femtosecond pulses from high-power fiber lasers,” Phys. Rev. A81, 053815 (2010). [CrossRef]
  21. J. Gordon, “Theory of the soliton self-frequency shift,” Opt. Lett.11, 662–664 (1986). [CrossRef] [PubMed]
  22. M. Charbonneau-Lefort, B. Afeyan, and M. M. Fejer, “Optical parametric amplifiers using chirped quasi-phase-matching gratings I: practical design formulas,” J. Opt. Soc. Am. B25, 463–480 (2008). [CrossRef]
  23. C. R. Phillips and M. M. Fejer, “Efficiency and phase of optical parametric amplification in chirped quasi-phase-matched gratings,” Opt. Lett.35, 3093–3095 (2010). [CrossRef] [PubMed]
  24. C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “Ultrabroadband, highly flexible amplifier for ultrashort midinfrared laser pulses based on aperiodically poled Mg:LiNbO3,” Opt. Lett.35, 2340–2342 (2010). [CrossRef] [PubMed]
  25. R. DeSalvo, A. Said, D. Hagan, E. Van Stryland, and M. Sheik-Bahae, “Infrared to ultraviolet measurements of two-photon absorption and n2 in wide bandgap solids,” IEEE J. Quantum Electron.32, 1324–1333 (1996). [CrossRef]
  26. I. Shoji, T. Kondo, A. Kitamoto, M. Shirane, and R. Ito, “Absolute scale of second-order nonlinear-optical coefficients,” J. Opt. Soc. Am. B14, 2268–2294 (1997). [CrossRef]
  27. M. Bache, O. Bang, B. B. Zhou, J. Moses, and F. W. Wise, “Optical cherenkov radiation in ultrafast cascaded second-harmonic generation,” Phys. Rev. A82, 063806 (2010). [CrossRef]
  28. C. R. Phillips, J. Jiang, C. Langrock, M. M. Fejer, and M. E. Fermann, “Self-Referenced Frequency Comb From a Tm-fiber Amplifier via PPLN Waveguide Supercontinuum Generation,” in CLEO:2011 - Laser Applications to Photonic Applications, OSA Technical Digest (CD) (Optical Society of America, 2011), paper PDPA5.
  29. A. S. Barker and R. Loudon, “Dielectric properties and optical phonons in LiNbO3,” Phys. Rev.158, 433 (1967). [CrossRef]
  30. P. J. Delfyett, R. Dorsinville, and R. R. Alfano, “Spectral and temporal measurements of the third-order nonlinear susceptibility of LiNbO3 using picosecond Raman-induce phase-conjugate spectroscopy,” Phys. Rev. B40, 1885 (1989). [CrossRef]
  31. N. Surovtsev, V. Malinovskii, A. Pugachev, and A. Shebanin, “The nature of low-frequency raman scattering in congruent melting crystals of lithium niobate,” Phys. Solid State45, 534–541 (2003). [CrossRef]
  32. R. Schiek, R. Stegeman, and G. I. Stegeman, “Measurement of third-order nonlinear susceptibility tensor elements in lithium niobate,” in “Frontiers in Optics,” (Optical Society of America, 2005), p. JWA74.
  33. G. I. Stegeman, D. J. Hagan, and L. Torner, “χ(2) cascading phenomena and their applications to all-optical signal processing, mode-locking, pulse compression and solitons,” Opt. Quantum Electron.28, 1691–1740 (1996). [CrossRef]
  34. C. Conti, S. Trillo, P. Di Trapani, J. Kilius, A. Bramati, S. Minardi, W. Chinaglia, and G. Valiulis, “Effective lensing effects in parametric frequency conversion,” J. Opt. Soc. Am. B19, 852–859 (2002). [CrossRef]
  35. D. H. Jundt, “Temperature-dependent Sellmeier equation for the index of refraction, ne, in congruent lithium niobate,” Opt. Lett.22, 1553–1555 (1997). [CrossRef]
  36. R. Boyd, “Stimulated Raman scattering and stimulated Rayleigh-Wing scattering,” in “Nonlinear Optics”, R. Boyd (Academic, 2008). [CrossRef]
  37. C. Heese, C. R. Phillips, L. Gallmann, M. M. Fejer, and U. Keller, “High-power mid-infrared optical parametric chirped-pulse amplifier based on aperiodically poled Mg:LiNbO3,” presented at the Conference on Lasers and Electro-optics (2011).
  38. E. D. Palik and G. Ghosh, Handbook of Optical Constants of Solids (Academic, 1985).

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