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Controlling Microresonator Solitons with the Counter-Propagating Pump
Controlling Microresonator Solitons with the Counter-Propagating Pump
Fan, Zhiwei;Skryabin, Dmitry V.
hv photonics Article Controlling Microresonator Solitons with the Counter-Propagating Pump 1 1,2, Zhiwei Fan and Dmitry V. Skryabin * Department of Physics, University of Bath, Bath BA2 7AY, UK; email@example.com Russian Quantum Center, 143026 Skolkovo, Russia * Correspondence: firstname.lastname@example.org Abstract: Considering a bidirectionally pumped ring microresonator, we provide a concise derivation of the model equations allowing us to eliminate the repetition rate terms and reduce the nonlinear interaction between the counter-propagating waves to the power-dependent shifts of the resonance frequencies. We present the simulation results of the soliton control by swiping the frequency of the counter-propagating wave in the forward and backward directions and with the soliton- blockade effect either present or not. We highlight the non-reciprocity of the forward and backward scans. Furthermore, we report the soliton crystals and breathers existing in the vicinity of the blockade interval. Keywords: frequency comb; microresonator; soliton 1. Introduction Frequency comb generation in the high-Q ring microresonators and the associated dissipative Kerr solitons have reached unprecedented heights of practical relevance . The complexity of the microresonator soliton properties is hard to exaggerate. A challenge has emerged after a series of experiments with the bidirectionally pumped microresonators, Citation: Fan, Z.; Skryabin, D.V. where combs and solitons [2–4] and symmetry breaking [5,6] have been observed in Controlling Microresonator Solitons counter-propagating waves. Studies into the gyroscope applications of these devices are with the Counter-Propagating Pump. also becoming increasingly important [7–11]. Photonics 2021, 8, 239. https:// The problem of the mathematical modeling of the multimode regimes of operation doi.org/10.3390/photonics8070239 of the bidirectional resonators has been addressed and to a large degree resolved only recently by demonstrating the equivalence of the coupled-mode model derived from Received: 29 May 2021 the ab initio Maxwell system to the envelope equations where the nonlinear interaction Accepted: 24 June 2021 (cross-phase modulation, XPM) between the counter-propagating waves is reduced to the Published: 26 June 2021 power-dependent shifts of the resonance frequencies and the fast oscillations with the repetition rate frequencies being eliminated [12–14]. Below, we recapture the main steps of Publisher’s Note: MDPI stays neutral the model derivation and highlight the transition from the four-envelope model accounting with regard to jurisdictional claims in published maps and institutional afﬁl- for the repetition rates to the two-envelope formulation that eliminates them. iations. The deeper insight into the independence of the XPM nonlinearity from the phase relations between the interacting modes [12–14] has recently led us to the prediction of the soliton-blockade effect . The blockade is achieved when the tuning frequency of one of the counter-rotating ﬁelds ﬁrst disrupts and then restores the soliton transmission in the other ﬁeld. The refractive index change of the soliton ﬁeld happens via the XPM Copyright: © 2021 by the authors. effect and is particularly strong in microresonators due to their high ﬁnesses providing Licensee MDPI, Basel, Switzerland. 3–6 orders of magnitude boost to the circulating powers relative to the input one . This article is an open access article distributed under the terms and This creates a variety of opportunities for efﬁcient nonlinear control of optical signals (see, conditions of the Creative Commons e.g., [17,18] and references therein), and, in particular, the soliton blockade effect utilizes Attribution (CC BY) license (https:// the high sensitivity of the nonlinear response of one of the ﬁelds towards the changes of creativecommons.org/licenses/by/ the driving frequency of the counter-propagating one. 4.0/). Photonics 2021, 8, 239. https://doi.org/10.3390/photonics8070239 https://www.mdpi.com/journal/photonics Photonics 2021, 8, 239 2 of 13 In Section 3, we look into how the power and frequency of the soliton and control pumps can be used to expand the soliton-blockade range. We also demonstrate there that the soliton-crystals, well known in the unidirectionally pumped resonators [19,20], emerge from the edges of the blockade interval. We performed both forward and backwards adiabatic scans of the control ﬁeld frequency and demonstrate non-reciprocity of the soliton-blockade effect (see Section 4). In Section 5, we present equations for perturbations around the solitons, which bare the features speciﬁc to the integral nature of the XPM terms and demonstrate the breather states close to the soliton-blockade interval (see, e.g., [21–23] for the breather studies in the unidirectional resonators). 2. Model 2.1. Equations, Field Envelopes and Modes We proceed by postulating that the real electric ﬁeld of a given transverse mode family in a multi-mode ring microresonator can be expressed as  bF(r, z)E(t, q). (1) The whole of the latter expression has units of volts per meter. q 2 (0, 2p] is the angular coordinate along the resonator circumference. t is physical time. b is the normalization constant that re-scalesE to the desirable units. E has units of power in what follows. F(r, z) is the transverse mode proﬁle normalized so that its maximum equals one. The spectrum ofE(t, q) is assumed to be sufﬁciently narrow to allow disregarding the dispersive changes of F(r, z). Separation of the spatial variables in Equation (1) is an approximation that generally works well in a typical microresonator. E can then be sought as a superposition of the two counter-propagating waves, i Mq iw t i Mq+iw t + + E = e Q (q, t) + e Q (q, t) + c.c. . (2) Here, M is the mode number with the resonance frequency w and w is the frequency 0 + of the laser pumping the plus-wave, so that d = w w , (3) 0 0 + deﬁnes the respective detuning. Q are the envelope functions, which can be expressed via their mode expansions as + imq imq Q = Q (t)e , Q = Q (t)e . (4) å m å m m m Q are the time dependent mode amplitudes. m = N/2 + 1, . . . , 0, . . . , N/2 are the 1 2 relative mode numbers and w = w + D m + D m are the corresponding resonance m 0 1 2 frequencies. D /2p is the resonator repetition rate and D /2p is dispersion. 1 2 The structure of the mode expansion in Equation (4) assumes that the effects of D , D , linewidth and nonlinearity will all be embraced by the equations derived for Q after Equation (4) are substituted to the Maxwell equations. Using w in the expo- nents in front of the mode expansions for the clockwise and counter-clockwise ﬁelds in Equation (4) implies that detuning between the frequencies of the lasers pumping the counter-rotating ﬁelds # = w w (5) will resurface in the equations for Q (see Figure 1 for a schematic illustration of the pump arrangements). Photonics 2021, 8, 239 3 of 13 Figure 1. A schematic illustration of the bidirectionally pumped resonator. The presence of the backscattering effects induces coupling between the counter- propagating waves. The authors of [13,14] formally traced what could be expected naturally, + imq imq namely the Q e wave couples to the Q e one. Comparing this coupling structure to m m the deﬁnition of the Q envelope function in Equation (4) suggests the need to deﬁne the envelope functions for the reﬂected waves, (r) (r) + imq imq Q = Q (t)e , Q = Q (t)e . (6) å å + m m m m (r) While the mode amplitudes entering the Q and Q envelopes are the same, the signs of the exponential terms are arranged differently (cf. Equations (4) and (6)). The envelope equations that follow are (r) 1 2 i¶ Q = d Q iD ¶ Q D ¶ Q RQ t + 0 + + 2 + 1 q 2! q 1 2 2 i k(Q H ) g(jQ j + 2jQ j )Q , (7a) + + + + (r) 1 2 i¶ Q = d Q + iD ¶ Q D ¶ Q R Q 0 1 q 2 2! q + 1 i#t 2 2 i k(Q H e ) g(jQ j + 2jQ j )Q , (7b) (r) (r) (r) (r) 1 2 i¶ Q = d Q + iD ¶ Q D ¶ Q RQ t 0 1 q 2 + + + q + (r) (r) (r) (r) 1 2 2 i k(Q H ) g(jQ j + 2jQ j )Q , (7c) + + + (r) (r) (r) (r) 1 2 i¶ Q = d Q iD ¶ Q D ¶ Q R Q t 0 1 q 2 + (r) (r) (r) (r) 1 i#t 2 2 i k(Q H e ) g(jQ j + 2jQ j )Q , (7d) (see  for the ﬁrst-principle derivation using the present scaling and notations). There are several parameters ﬁrst used in Equation (7) and hence requiring deﬁnitions . k is the loaded linewidth parameter. g is the nonlinear coefﬁcients measured, such that the units of g/2p are Hz/W and of gjQ j /2p are Hz. H are the pump parameters, H = hFW /p. Here, h < 1 is the coupling coefﬁcient, F = D /k is the resonator ﬁnesse andW are the on-chip powers of the two lasers. R is the backscattering coefﬁcient, which is taken equal for all the modes (see  for a more general treatment). 2.2. Redeﬁning the Envelope Functions and Eliminating the Repetition-Rate Terms Equations (7) contain the D terms with the opposite signs, therefore the transforma- tion into the rotating reference frame could remove the repetition-rate term, for example, from the plus equations, but then the repetition rate would simply double in the minus equations. At the same time, D is the strongly dominant frequency scale in the problem. Typically, it would be 10 GHz and up to 1 THz, while all the other terms, if taken for the respective range of resonators, would vary from 100 kHz to 100 MHz. This indicates that, if Photonics 2021, 8, 239 4 of 13 the multiples of D are traced in the nonlinear terms, then these terms would be oscillating with the fastest frequency in the model and could be disregarded [12–14]. To reveal the frequency scales associated with D , we replace the Q mode amplitudes in Equations (4) with imD t Q (t) = y (t)e , (8) m m so that the electric ﬁeld in Equation (2) is replaced with im q D t im q+D t i Mq iw t + i Mq+iw t + 1 + 1 E = e y (t)e + e y (t)e + c.c. . (9) å m å m m m Two new sets of the envelope functions (r) imq imq y = y (t)e , y = y (t)e , (10) å m å m m m play a pivotal role in the theory of the microresonators with the counter-propagating waves . Taking the discrete Fourier transforms of y allows reconstructingE (cf. Equations (2), (8) and (9)). Unlike the equations for Q , the ones for y do not contain the D terms and replace the (r) system for four envelopes, Q , Q , with the one for two, (r) 1 2 2 1 i¶ y = (d 2g )y D ¶ y Ry gjy j y i k(y H ), (11a) t + + + + + + + 0 2 2 q 2 (r) (r) (r) (r) (r) (r) 1 2 2 1 i#t i¶ y = (d 2g )y D ¶ y R y gjy j y i k(y H e ), (11b) t 0 + 2 + 2 q 2 Z Z 2p 2p dq dq (r) 2 2 2 g = g jy (t, q)j = g jy (t, q)j = g jy j . (11c) å m 2p 2p 0 0 (r) The above equations could be supplemented with equations for y , y , but this time they are left as an independent pair. With the fabrication techniques rapidly improving, the surface roughness losses, as well as the associated backscattering, can be reduced to such levels that the frequency scale of R drops below the dispersion and nonlinearity induced effects so that we set R = 0 in what follows. Including R 6= 0 represents a separate problem . The notable consequence of the elimination of the D oscillations from the dynamics is that the nonlinear coupling, i.e., the cross-phase modulation (XPM) between the counter- propagating waves, has lost its sensitivity to the phases of the individual modes [12–14] (see Equation (11c)). Now, the XPM is simply expressed by the nonlinear shifts of the detuning, i.e., by the 2g and 2g terms. In what follows, we remove the exponential term from Equation (11b) by applying an (r) (r) i#t obvious substitution, y ! y e , omit the ’(r)’ superscript, and deal with 2 2 1 1 i¶ y = (d 2g )y D ¶ y gjy j y i k(y H ), (12a) t + + + 2 + + + + + 2 2 2 2 1 1 i¶ y = (d 2g )y D ¶ y gjy j y i k(y H ), (12b) t + 2 2 2 where d = d , d = d + #. (13) + 0 0 3. Single-Mode, Single-Soliton and Soliton-Crystal States and Their Role in the Soliton-Blockade Effect The single-mode, m = 0, regime of the resonator operation in the plus and minus ﬁelds is called here the cw-cw state (where cw stands for continuous wave). The respective modal amplitudes can be expressed as i kH (cw) y = , (14) d 2g g i k 2 Photonics 2021, 8, 239 5 of 13 where g solve the coupled real algebraic equations 4g g + [d 2g ] g = gH , (15a) + + + + 4g g + [d 2g ] g = gH , (15b) (cw) g = gjy j . (15c) Apart from the cw-cw states, there are also the soliton-cw, cw-soliton and soliton- soliton states, where the ﬁrst and second words in the solution classiﬁcation characterize the ﬁeld in the y and y components, respectively . Figure 2a shows how the cw-cw solution varies with d for a set of the ﬁxed values of the frequency offset parameter, # = w w . Large j#j separate and make quasi- independent the resonance structures in the two components along the d axis, i.e., if the g vs. d plot has the nonlinearity tilted resonance originating at d = 0, then g has + 0 the similar resonance starting at d #. Varying # from the relatively large negative to the large positive values drags the g resonance across the effectively immobile g one (see Figure 2a). Two resonances overlap and interact strongly for j#j/k . 1. For # = 0 and H = H , the model becomes symmetric and, therefore, its symmetric solution (see the black lines in the # = 0 panels of Figure 2a) goes through the symmetry breaking bifurcation [5,6,15]. The soliton-cw states are solutions of the coupled differential-algebraic system 1 2 2 1 (d 2g )y D ¶ y gjy j y i k(y H ) = 0, (16a) 0 + 2 + + + + + 2 2 2 1 (d + # 2g )y gjy j y i k(y H ) = 0, (16b) 0 + 2p dq 2 2 g = g jy (q)j , g = gjy j . (16c) + + 2p If j#j/k is large, then Equation (16a) becomes quasi-independent from Equation (16b), implying that the system operates in the limit where it is approximately reduced to the Lugiato–Lefever model having the bright soliton solutions around the bistability interval for d > 0 . However, whenj#j becomes the order of the linewidth, the two equations start interacting strongly, and the low power branch of the cw-state in the plus-ﬁeld becomes distorted by the growth of the second resonance peak, see the # < 0 cases in the top row in Figure 2. This interaction also creates the narrow parameter range, where the low power branch of the plus ﬁeld ceases to exist (see the yellow interval in Figure 2b). Around and substan- tially beyond this range, the soliton pulses in the plus-ﬁeld are not able to exist simply because the system does not offer a low-amplitude background state for the bright soliton to nest on (see the grey stripe in Figure 2b and the respective no-soliton range in Figure 2c). This is the essence of the soliton-blockade effect, when the solitons can be switched on and off by tuning the frequency of the counter-propagating ﬁeld, so that # is tuned within the soliton-forbidden interval . Figure 3 shows how the soliton-blockade domains are shaped in the (d , #) and (H , #) parameter spaces, respectively. In particular, the blockade regime is more easily accessible if the minus , i.e., the cw component is pumped harder (see Figure 3b). Figure 4 shows how the XPM coefﬁcients for the plus, g , and minus, g , components of the soliton-cw state (see Equation (16c)) vary with # for the value of d slightly different from the one in Figure 2c. Figure 4a plots the soliton g for the one-, four-, and six-soliton states, i.e., for the soliton crystals. The respective g values corresponding to the cw- component are shown in Figure 4b, and the soliton crystal proﬁles are in Figure 5. While the more detailed understanding of the bifurcations of the soliton crystals [19,20] in the bi-directionally pumped geometry goes beyond our present scope. Photonics 2021, 8, 239 6 of 13 (cw) Figure 2. (a) g = gjy j for the cw-cw states vs. d for a set of #. (b) g for the cw-cw state 0 + vs. # for d /2p = 4.5 MHz. (c) Angular proﬁles of the soliton-cw states vs. #. The grey shaded interval in (b) and the no-solution interval in (c) is the soliton-blockade interval (see text for further details and the same and similar datasets published by us in ). Other parameters are H = 16 W, k/2p = 1.5 MHz, F = 10 , D /2p = 10 kHz, g/2p = 0.4 MHz/W, h = 0.5. 2 Photonics 2021, 8, 239 7 of 13 Figure 3. The soliton-blockade regions in the (d , #) and (H /H , #) planes. (a) H = 16 W; (b) H = 16 W, d /2p = 6 MHz. Figure 4. Soliton-cw states for d /2p = 6 MHz. The other parameters are as in Figure 2c: (a) g (soliton); and (b) g (cw) vs. #. The numbers 1, 4, and 6 indicate the soliton-crystal states with the respective number of the equally spaced solitons (Figure 5). Photonics 2021, 8, 239 8 of 13 Figure 5. Soliton crystal states for #/2p = 1 MHz. (a,b) are the spatial proﬁles and (c,d) are the respective spectra. The other parameters are as in Figure 4. 4. Controlling Solitons by Tuning the Cw-Component: Direct and Reverse Scans We now present the results of the dynamic simulations of Equation (11) by initializing them with the cw-cw states and by scanning the frequency of the control (minus) ﬁeld adiabatically. The adiabatic scans of # were implemented in both forward (# goes from negative to positive) and reverse directions. The outcomes of the two scans are generally different due to the complex and non-symmetric structure of the bistable resonances (see the cw-cw vs. # plot in Figure 2b). First, we choose H /H = 1, and continue to keep d on the positive side, so that + 0 the plus solitons would always exist in the absence of the interaction with the minus ﬁeld. The outcomes of the forward and backward #-scans are shown in the left and right columns of Figure 6, respectively. For the forward scan, the plus solitons keep being generated before #/2p ' 4 MHz, then disappear in the soliton-blockade region and then reappear for #/2p & 7 MHz. Throughout the blockade interval, the intraresonator ﬁeld most typically converges to the turbulent state known for the near-bistability operation in the uni-directional setting. For #/2p & 2 MHz, one can also see that the minus-ﬁeld starts generating solitons. In the forward scan, the soliton-blockade interval for the plus ﬁeld is wider than predicted by the time-independent methods in the previous section. This is because the eight solitons formed in the minus ﬁeld provide the value of g sufﬁcient to keep d 2g outside the plus-soliton existence interval. Simultaneously, the value of d + # 2g is such that the minus ﬁeld can sustain the solitons. The variations of the net power in either Photonics 2021, 8, 239 9 of 13 of the ﬁelds can lead to different scenarios as # is increased. The one that is realized in the dataset shown in Figure 6 is that the downwards ﬂuctuation of g takes the minus ﬁeld outside the soliton regime and brings it to the low power cw-state. Simultaneously, this transition restores the soliton regime in the plus ﬁeld itself. The exit from the blockade regime is also sensitive to the chosen value of H /H and the scan parameters. Generally, nonlinear dynamics under conditions when a parameter is swept across the bifurcation points is an area of research attracting attention in optics and beyond (see, e.g., [24,25] and references therein). Figure 6. The soliton blockade in the plus ﬁeld as is realized using the forward (a1) and reverse (a2) 1/2 adiabatic scans of #, i.e., by tuning w ; d /2p = 4.5 MHz, H = H = 4W . (b1,b2) The magenta 0 + line shows g vs. time. The blue and red lines are g for the cw-cw solution. (c1,c2) The same as + + (a1,a2) but for the minus ﬁeld. (d1,d2) The same as (b1,b2) but for g . The reverse, i.e., #-positive to #-negative, scan shows the much narrower range of the soliton-blockade in the plus ﬁeld, and no solitons in the minus ﬁeld (see the right column in Figure 6). This is because the intra-resonator ﬁeld in the minus component picks and follows the stable low power cw-state until # comes to near zero. Therefore, in the reverse scan, the dynamically seen soliton-blockade range is practically the same as the one predicted by the time-independent analysis in the previous section. We now choose H /H = 1/2, which makes the soliton-blockade impossible (see Figure 3b). The outcomes of the respective forward scan are shown in Figure 7. Following our expectations, we see only the stable solitons in the plus ﬁeld. In the minus ﬁeld, the periodic pattern is induced in the narrow range of ds where the low amplitude cw state is Photonics 2021, 8, 239 10 of 13 unstable. The uninterrupted solitons seeing in the plus ﬁeld during the scan are correlated with the continuous existence of the low amplitude cw-state (see the red line in Figure 7). The narrow and low red bistability peak disturbs the pulse but does not switch the plus ﬁeld into a persistent non-soliton regime (cf. Figure 6). The data in Figures 6 and 7 were generated while adiabatically tuning d /2p between 10 and 30 MHz over the 1 ms time interval (the round-trip time 60 fs, D /2p = 15 GHz). Figure 7. The forward scan performed outside the soliton-blockade conditions, H /H = 1/2, 1/2 H = 2W , d /2p = 6 MHz (cf. Figure 3b): (a,b) the plus and minus ﬁelds during the scan; and (c) g for the cw-cw state over the same interval of #. 5. Soliton Stability and Breathers The problem of determining soliton stability is worthy of some attention, in particular, because the g terms in the governing equations represent a certain challenge in this regard. The soliton stability is determined by computing the eigenvalue of the Jacobian matrix. In order to work out the latter, we seek solutions of Equation (12) as y = A(q) + a(q, t), y = B(q) + b(q, t), where a, b are the small perturbations on top of the soliton or any other spatially inhomogeneous state, y = A(q), y = B(q). + Photonics 2021, 8, 239 11 of 13 Truncating the second and higher orders of a and b, we are left with 2p 1 1 g 2 2 2 2 0 i¶ a = d a D ¶ a 2gj Aj a g A a i k a a jBj dq t + 2 2 2 p Z Z 2p 2p g g 0 0 A Bb dq A B bdq , p p 0 0 2p 1 1 g 2 2 2 2 0 i¶ a = d a + D ¶ a + 2gj Aj a + g A a i k a + a jBj dq t + 2 2 2 p Z Z 2p 2p g g 0 0 + A B bdq + A Bb dq , (17) p p 0 0 2p 1 1 g 2 2 2 2 0 i¶ b = d b D ¶ b 2gjBj b gB b i kb b j Aj dq t 2 2 2 p Z Z 2p 2p g g 0 0 B Aa dq B A adq , p p 0 0 2p 1 1 g 2 2 2 2 0 i¶ b = d b + D ¶ b + 2gjBj b + gB b i kb + b j Aj dq 2 2 p Z Z 2p 2p g g 0 0 + B A adq + B Aa dq . p p 0 0 If the integrals are replaced with the sums, then Equation (17) becomes ¶ ~# = i M~#, where~# = [a, a , b, b ] . The growth rate of the perturbations is introduced by setting lt ~# e . The instability occurs if the real part of l is positive. We computed the soliton spectrum around the blockade interval and found the oscillatory instabilities that stabilize away from the interval (see Figure 8a). When the instabilities are present, they typically lead to the formation of the soliton breather states with different periods (see Figure 8b,c). Figure 8. (a) The maximal growth rate of the soliton instability found on the right from the blockade interval in Figure 2c. (b,c) The soliton breather states found for the green and blue dots in (a), respectively. Photonics 2021, 8, 239 12 of 13 6. Summary We further investigated the soliton-blockade effect in the bidirectional microring resonators. In particular, we presented the simulation results on the soliton control by swiping the frequency of the counter-propagating wave in the forward and backward directions and found that the latter provides a much narrower blockade interval and a wider range of the soliton existence. Furthermore, we demonstrated that the blockade interval can be expanded by increasing the power of the control ﬁeld and that the soliton crystals and breathers exist on the both sides of it. Author Contributions: Data curation, Z.F.; Writing—review and editing, D.V.S. All authors have read and agreed to the published version of the manuscript. Funding: Russian Science Foundation grant number 17-12-01413-P; EU Horizon 2020 Framework Programme (812818, MICROCOMB). Acknowledgments: We acknowledge the illuminating discussions with Magnus Johansson. Conﬂicts of Interest: The authors declare no conﬂict of interest. References 1. Kippenberg, T.J.; Gaeta, A.L.; Lipson, M.; Gorodetsky, M.L. 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Multidisciplinary Digital Publishing Institute
Controlling Microresonator Solitons with the Counter-Propagating Pump
Skryabin, Dmitry V.
, Volume 8 (7) –
Jun 26, 2021
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