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Optics Express

Optics Express

  • Editor: Andrew M. Weiner
  • Vol. 21, Iss. 2 — Jan. 28, 2013
  • pp: 1374–1394
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Schmidt decompositions of parametric processes I: Basic theory and simple examples

C. J. McKinstrie and M. Karlsson  »View Author Affiliations


Optics Express, Vol. 21, Issue 2, pp. 1374-1394 (2013)
http://dx.doi.org/10.1364/OE.21.001374


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Abstract

Parametric devices based on four-wave mixing in fibers perform many signal-processing functions required by optical communication systems. In these devices, strong pumps drive weak signal and idler sidebands, which can have one or two polarization components, and one or many frequency components. The evolution of these components (modes) is governed by a system of coupled-mode equations. Schmidt decompositions of the associated transfer matrices determine the natural input and output mode vectors of such systems, and facilitate the optimization of device performance. In this paper, the basic properties of Schmidt decompositions are derived from first principles and are illustrated by two simple examples (one- and two-mode parametric amplification). In a forthcoming paper, several nontrivial examples relevant to current research (including four-mode parametric amplification) will be discussed.

© 2013 OSA

1. Introduction

Parametric devices based on four-wave mixing (FWM) in fibers can amplify, frequency convert, phase conjugate, regenerate and sample optical signals in classical communication systems [1

1. M. E. Marhic, Fiber Optical Parametric Amplifiers, Oscillators and Related Devices (Cambridge, 2007). [CrossRef]

6

6. S. Radic, “Parametric signal processing,” IEEE J. Sel. Top. Quantum Electron. 18, 670–680 (2012). [CrossRef]

]. They can also generate photon pairs for quantum information experiments [7

7. R. Loudon, The Quantum Theory of Light, 3rd Ed. (Oxford, 2000).

10

10. O. Cohen, J. S. Lundeen, B. J. Smith, G. Puentes, P. J. Mosley, and I. A. Walmsley, “Tailored photon-pair generation in optical fibers,” Phys. Rev. Lett. 102, 123603 (2009). [CrossRef] [PubMed]

]. Three different types of FWM are illustrated in Fig. 1. Modulation interaction (MI) is the degenerate process in which two photons from the same pump are destroyed, and signal and idler (sideband) photons are created (2πpπs + πi, where πj represents a photon with frequency ωj). Inverse MI is the degenerate process in which two photons from different pumps are destroyed and two signal photons are created (πp + πq → 2πs). Phase conjugation (PC) is the nondegenerate process in which two different pump photons are destroyed and two different sideband photons are created (πp + πqπs + πi). The polarization properties of these processes are reviewed in [11

11. K. Inoue, “Polarization effect on four-wave mixing efficiency in a single-mode fiber,” IEEE J. Quantum Electron. 28, 883–894 (1992). [CrossRef]

14

14. C. J. McKinstrie, H. Kogelnik, and L. Schenato, “Four-wave mixing in a rapidly-spun fiber,” Opt. Express 14, 8516–8534 (2006). [CrossRef] [PubMed]

].

Fig. 1 Frequency diagrams for (a) modulation interaction, (b) inverse modulation interaction, and (c) outer-band and (d) inner-band phase conjugation. Long arrows denote pumps (p and q), whereas short arrows denote sidebands (s and i). Downward arrows denote modes that lose photons, whereas upward arrows denote modes that gain photons.

Parametric interactions of weak sidebands, driven by strong pumps, are governed by coupled-mode equations (CMEs) of the form
dzX=AX+BX*,
(1)
where z is distance, dz = d/dz, X = [xj] is the vector of sideband amplitudes (modes), A = [αjk] and B = [βjk] are coefficient matrices, and * denotes a complex conjugate. The entries of the amplitude vector could be the amplitudes of distinct monochromatic sidebands (continuous waves), or different frequency components of multichromatic sidebands (pulses), with one or two polarization components. For uniform fibers (media) the coupling coefficients are constants, whereas for nonuniform media they vary with distance. Because Eq. (1) is linear in the amplitude vector and its conjugate, the (explicit or implicit) solution of Eq. (1) can be written in the input–output (IO) form
X(z)=M(z)X(0)+N(z)X*(0),
(2)
where M = [μjk] and N = [νjk] are transfer (Green) matrices.

Recall that every complex matrix M has the singular value (Schmidt) decomposition M = VDU, where U and V are unitary and D is diagonal [15

15. G. W. Stewart, “On the early history of the singular value decomposition,” SIAM Rev. 35, 551–566 (1993). [CrossRef]

, 16

16. G. J. Gbur, Mathematical Methods for Optical Physics and Engineering (Cambridge, 2011), Sec. 5.4.

]. The columns of U are the eigenvectors of MM, the columns of V are the eigenvectors of MM, and the entries of D are the (common) non-negative eigenvalues of MM and MM. In the context of parametric interactions, the laws of Hamiltonian mechanics impose constraints on the transfer matrices M and N, which ensure that they have the simultaneous (related) decompositions M = VDμU and N = VDνUt, where Dμ = diag(μj), Dν = diag(νj) and their entries (Schmidt coefficients) satisfy the auxiliary equations μj2νj2=1[17

17. A. K. Ekert and P. L. Knight, “Relationship between semiclassical and quantum-mechanical input-output theories of optical response,” Phys. Rev. A 43, 3934–3938 (1991). [CrossRef] [PubMed]

, 18

18. S. L. Braunstein, “Squeezing as an irreducible resource,” Phys. Rev. A 71, 055801 (2005). [CrossRef]

]. Hence, Eq. (2) can be rewritten in the form
X(z)=VDμUX(0)+VDνUtX*(0),
(3)
where U, V, Dμ and Dν depend implicitly on z. It follows from Eq. (3) that the columns of U define input (Schmidt) mode vectors, the columns of V define output vectors, and the mode amplitudes j(0) and j(z), which are the components of X relative to the input and output Schmidt bases, respectively, satisfy the one-mode squeezing equations
x¯j(z)=μj(z)x¯j(0)+νj(z)x¯j*(0).
(4)
Equations (3) and (4) are remarkable. They imply that every parametric process, no matter how complicated, can be decomposed into independent squeezing processes, the properties of which are known (in-phase signal quadratures are stretched, whereas out-of-phase quadratures are squeezed) [7

7. R. Loudon, The Quantum Theory of Light, 3rd Ed. (Oxford, 2000).

, 19

19. H. P. Yuen, “Two-photon states of the radiation field,” Phys. Rev. A 13, 2226–2243 (1976). [CrossRef]

, 20

20. C. M. Caves, “Quantum limits on noise in linear amplifiers,” Phys. Rev. D 26, 1817–1839 (1982). [CrossRef]

]. This decomposition is the answer to the second of the aforementioned questions. The first question is harder to answer, but in a forthcoming paper it will be shown that the (known) form of the decomposition sometimes provides guidance that facilitates the solution of the CMEs.

The simultaneous Schmidt decomposition is more than an elegant mathematical result. Determining the Schmidt modes and coefficients of a parametric process, and the device based upon it, facilitates the optimization of device performance. For example, in photon pair generation by pulsed pumps, the Schmidt modes are the (temporal) wave-packets of the signal and idler photons, and the squares of the Schmidt coefficients are the probabilities with which photon pairs are produced [21

21. C. K. Law, I. A. Walmsley, and J. H. Eberly, “Continuous frequency entanglement: Effective finite Hilbert space and entropy control,” Phys. Rev. Lett. 84, 5304–5307 (2000). [CrossRef] [PubMed]

, 22

22. W. P. Grice, A. B. U’Ren, and I. A. Walmsley, “Eliminating frequency and space-time correlations in multiphoton states,” Phys. Rev. A 64, 063815 (2001). [CrossRef]

]. If one designs the system so that only one coefficient is nonzero, the output state is pure, as required for a variety of quantum information experiments. In photon frequency conversion by pulsed pumps, the Schmidt modes are the natural input and output wave-packets of the signal and idler photons, and the squares of the Schmidt coefficients are the conversion probabilities [23

23. M. G. Raymer, S. J. van Enk, C. J. McKinstrie, and H. J. McGuinness, “Interference of two photons of different color,” Opt. Commun. 283, 747–752 (2010). [CrossRef]

, 24

24. C. J. McKinstrie, L. Mejling, M. G. Raymer, and K. Rottwitt, “Quantum-state-preserving optical pulse reshaping and multiplexing by four-wave mixing in fibers,” Phys. Rev. A 85, 053829 (2012). [CrossRef]

]. One can optimize two-photon interference (or single-photon conversion) experiments by designing the system so that at least one squared coefficient is 0.5 (or 1.0). It is well known that one-mode squeezing and stretching processes dilate the coherent and incoherent parts of input signals by the same amounts. The former processes are of interest in quantum optics, because the out-of-phase quadratures have smaller fluctuations than vacuum quadratures [19

19. H. P. Yuen, “Two-photon states of the radiation field,” Phys. Rev. A 13, 2226–2243 (1976). [CrossRef]

, 20

20. C. M. Caves, “Quantum limits on noise in linear amplifiers,” Phys. Rev. D 26, 1817–1839 (1982). [CrossRef]

], whereas the latter are of interest in optical communications, because they can amplify in-phase signals without degrading their signal-to-noise ratios [25

25. C. J. McKinstrie, M. Yu, M. G. Raymer, and S. Radic, “Quantum noise properties of parametric processes,” Opt. Express 13, 4986–5012 (2005). [CrossRef] [PubMed]

, 26

26. Z. Tong, C. Lundström, P. A. Andrekson, M. Karlsson, and A. Bogris, “Ultralow noise, broadband phase-sensitive optical amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron. 18, 1016–1032 (2012). [CrossRef]

]. To minimize the effects of noise in a communication system, one should encode information in the Schmidt modes of the system (superpositions of frequency or polarization components), not the physical modes (individual components). Encoding information in this way also maximizes the information capacity of the system [27

27. E. Telatar, “Capacity of multi-antenna Gaussian channels,” Euro. Trans. Telecom. 10, 585–595 (1999). [CrossRef]

, 28

28. C. J. McKinstrie and N. Alic, “Information efficiencies of parametric devices,” J. Sel. Top. Quantum Electron. 18, 794–811 (2012). [CrossRef]

].

Aspects of Schmidt decompositions were also discussed in [29

29. C. J. McKinstrie, “Unitary and singular value decompositions of parametric processes in fibers,” Opt. Commun. 282, 583–593 (2009). [CrossRef]

, 30

30. D. A. Edwards, J. D. Fehribach, R. O. Moore, and C. J. McKinstrie, “An application of matrix theory to the evolution of coupled modes,” to appear in SIAM Rev.

]. In these articles, the mathematical properties of Schmidt decompositions, and the adjoint decompositions to which they are related, were described in detail, and some simple examples (including one- and two-mode amplification) were mentioned briefly. In this paper, the physical consequences of Schmidt decompositions are emphasized, the aforementioned simple examples are discussed in detail and key results are explained in the context of Hamiltonian dynamics. In a forthcoming paper, new solutions of the CMEs are obtained and used to discuss several nontrivial examples of current interest (including four-mode parametric amplification).

2. Simple examples of Schmidt decompositions

2.1. One-mode amplification

Consider a one-mode parametric process (inverse MI), which is governed by the equation
dzx=iδx+iγx*,
(5)
where x is the mode amplitude, δ is the (real) mismatch coefficient and γ is the (complex) coupling coefficient [31

31. C. J. McKinstrie and S. Radic, “Phase-sensitive amplification in a fiber,” Opt. Express 12, 4973–4979 (2004). [CrossRef] [PubMed]

, 32

32. K. Croussore and G. Li, “Phase and amplitude regeneration of differential phase-shift keyed signals using phase-sensitive amplification,” IEEE J. Sel. Top. Quantum Electron. 14, 648–658 (2008). [CrossRef]

]. Equation (5) depends linearly on x and x*, so its solution can be written in the input–output (IO) form
x(z)=μ(z)x(0)+ν(z)x*(0).
(6)
For the common case in which δ and γ are constants, the transfer (Green) functions
μ(z)=cos(kz)+iδsin(kz)/k,ν(z)=iγsin(kz)/k,
(7)
and the characteristic wavenumber k = (δ2 − |γ|2)1/2. If coupling is stronger than mismatch (|γ| > δ), the system is unstable. The transfer functions (7) satisfy the auxiliary equation
|μ(z)|2|ν(z)|2=1.
(8)
They also have the interesting properties
μ(z)=μ*(z),ν(z)=ν(z).
(9)
Furthermore, if γ is real, then ν*(z) = −ν(z). These properties are not accidental.

Because the transfer functions usually have different phases, some analysis is required to determine the consequences of Eq. (6). Let μ = |μ|eμ and ν = |ν|eν, and define the sum and difference phases ϕs = (ϕμ + ϕν)/2 and ϕd = (ϕνϕμ)/2, respectively. Then Eq. (6) can be rewritten in the form
x(z)=v|μ|u*x(0)+v|ν|ux*(0),
(10)
where u = ed and v = es are phase factors (input and output phase references). If the signal phase ϕx = ϕu = ϕd, the terms on the right side of Eq. (10) add constructively: The signal is said to be in-phase and is amplified (stretched) by the factor |μ| + |ν|. Conversely, if ϕx = ϕd + π/2, the terms on the right side of Eq. (10) add destructively: The signal is said to be out-of-phase and is attenuated (squeezed) by the factor |μ|+ |ν| = 1/(|μ|−|ν|). If one were to measure the phase of the input signal relative to the aforementioned reference phase, one would say that the real quadrature is amplified and the imaginary quadrature is attenuated. Notice that Eq. (10) has the canonical form of Eq. (3).

It is instructive to formalize the method of solution. Equation (5) and its conjugate can be written in the augmented matrix form
dzY=iLyY,
(11)
where the 2 × 1 mode vector Y = [x, x*]t and the 2 × 2 coefficient matrix
Ly=[δγγ*δ].
(12)
Notice that Ly is specified by three real parameters (δ, γr and γi). The solution of Eq. (11) can be written in the IO form
Y(z)=Ty(z)Y(0),
(13)
where the transfer (Green) matrix
Ty(z)=exp(iLyz).
(14)
Two important results follow from Eqs. (12) and (14). First, tr(Ly) = 0, so det(Ty) = 1, and second, Ty(z)=Ty1(z). Because Eqs. (11) and (13) describe two copies of the same process (the original and its conjugate), the transfer matrix can be written in the form
Ty(z)=[μ(z)ν(z)ν*(z)μ*(z)].
(15)
Notice that det(Ty) = |μ|2 − |ν|2 = 1, so Ty is defined by three real parameters (|ν|,ϕμ and ϕν), the same number that specified Ly. Notice also that
Ty(z)=[μ(z)ν(z)ν*(z)μ*(z)]=Ty1(z)=[μ*(z)ν(z)ν*(z)μ(z)].
(16)
Hence, μ(−z) = μ*(z) and ν(−z) = −ν(z), as stated in Eqs. (9).

It was stated in Sec. 1 that the Schmidt vectors of Ty are the eigenvectors of TyTy and TyTy, and the Schmidt coefficients are the square roots of the (common) eigenvalues of these matrices. One can determine these eigensystems explicitly, or simply verify that
Ty(z)=121/2[eiϕseiϕseiϕseiϕs][|μ|+|ν|00|μ||ν|]121/2[eiϕdeiϕdeiϕdeiϕd],
(17)
where ϕs and ϕd were defined before Eq. (10). All three matrices in Eq. (17) depend on z. The evolution equation (11) governs x and x* simultaneously, so it is natural that the associated transfer matrix describes stretching and squeezing simultaneously. Specifically, Eq. (17) shows that the stretching condition is 2ϕx = 2ϕd, whereas the squeezing condition is 2ϕx = 2ϕd + π, and the associated Schmidt coefficients are reciprocals. The input Schmidt vectors, which are the natural inputs for one-mode amplification, correspond to in-phase and out-of-phase signals. These results are consistent with the discussion that follows Eq. (10).

Equations (15) and (16) show that there is a simple relation between the transfer matrix and its inverse. The replacements μμ* and ν → −ν (which do not affect the moduli of the transfer functions) are equivalent to ϕμ → −ϕμ and ϕνϕν + π and, ultimately, to esied and edies. By making these replacements in decomposition (17), one obtains the inverse decomposition
Ty1(z)=121/2[ieiϕdieiϕdieiϕdieiϕd][|μ|+|ν|00|μ||ν|]121/2[ieiϕsieiϕsieiϕsieiϕs].
(18)
For reference, neither the forward decomposition (17), nor the backward decomposition (18), is unique.

2.2. Two-mode amplification

Now consider a two-mode parametric process (MI or PC), which is governed by the CMEs
dzx1=iδ1x1+iγx2*,dzx2=iδ2x2+iγx1*,
(19)
where xj is a mode amplitude and δj is a (real) mismatch coefficient [1

1. M. E. Marhic, Fiber Optical Parametric Amplifiers, Oscillators and Related Devices (Cambridge, 2007). [CrossRef]

, 2

2. J. Hansryd, P. A. Andrekson, M. Westlund, J. Li, and P. O. Hedekvist, “Fiber-based optical parametric amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron. 8, 506–520 (2002). [CrossRef]

]. We will refer to mode 1 as the signal and mode 2 as the idler. The solutions of Eqs. (19) can be written in the IO forms
x1(z)=μ11(z)x1(0)+ν12(z)x2*(0),x2(z)=μ22(z)x2(0)+ν21(z)x1*(0).
(20)
If δj and γ are constants, the two-mode transfer functions μ11(z) = e(z)μ(z), ν12(z) = e(z)ν(z), μ22(z) = e*(z)μ(z) and ν21(z) = e*(z)ν(z), where the one-mode transfer functions μ(z) and ν(z) are defined by Eqs. (7), with the mismatch coefficient δ replaced by (δ1 + δ2)/2, and the phase factor e(z) = exp[i(δ1δ2)z/2] = exp(δ). The two-mode transfer functions satisfy the auxiliary equations
|μ11(z)|2|ν12(z)|2=1,|μ11(z)|2|ν21(z)|2=1,
(21)
and have the interesting properties
μ11(z)=μ11*(z),ν12(z)=ν21(z).
(22)
Furthermore, if γ is real, then ν12*(z)=ν21(z). One obtains additional constraints and properties by interchanging the subscripts 1 and 2. For the special case in which δ1 = δ2, solutions (20) reduce to solution (6).

In the first of Eqs. (20), the transfer functions have the common factor e(z), which affects the output signal phase, but does not affect the interference conditions. Hence, if ϕ1 + ϕ2 = ϕνϕμ = 2ϕd, the terms in the first of Eqs. (20) add constructively: The sidebands are said to be in-phase and (if their input amplitudes are equal) are stretched by the factor |μ| + |ν|. Conversely, if ϕ1 + ϕ2 = ϕνϕμ + π, the terms in the first of Eqs. (20) add destructively: The sidebands are said to be out-of-phase and (if their input amplitudes are equal) are squeezed by the factor |μ| + |ν| = 1/(|μ| − |ν|). The same interference conditions apply to the second of Eqs. (20).

Equations (19) can be written in the standard matrix form
dzX=iLxX,
(23)
where the 2 × 1 mode vector X=[x1,x2*]t and the 2 × 2 coefficient matrix
Lx=[δ1γγ*δ2].
(24)
Notice that Lx is specified by four real parameters (δ1, δ2, γr and γi). Notice also that Eq. (23) is closed (involves only x1 and x2*), so no augmentation is necessary. (The equation for [x1*,x2]t is simply the conjugate of the stated equation and contains no extra information.) The solution of Eq. (23) can be written in the IO form
X(z)=Tx(z)X(0),
(25)
where the transfer matrix
Tx(z)=exp(iLxz).
(26)
The properties of Lx and Tx [Eqs. (24) and (26)] differ only slightly from those of Ly and Ty [Eqs. (12) and (14)]. Because tr(Lx) = δ1δ2 ≠ 0, det(Tx) = exp[i(δ1δ2)z] ≠ 1. Nonetheless, Tx(−z) = T−1(z).

It follows from Eqs. (20) that the transfer matrix
Tx(z)=[μ11(z)ν12(z)ν21*(z)μ22*(z)]=e(z)[μ(z)ν(z)ν*(z)μ*(z)].
(27)
Notice that Tx is determined by four real parameters (|ν|, ϕμ, ϕν and ϕδ), the same number that specified Lx. Notice also that
Tx(z)=e(z)[μ(z)ν(z)ν*(z)μ*(z)]=Tx1(z)=e*(z)[μ*(z)ν(z)ν*(z)μ(z)].
(28)
Hence, μ11(z)=μ11*(z) and ν12(−z) = −ν21(z), as stated in Eqs. (22).

The only difference between Eqs. (15) and (27) is the phase factor e(z) = eδ, so it follows from Eq. (17) that the forward transfer matrix has the Schmidt decomposition
Tx(z)=eiϕδ21/2[eiϕseiϕseiϕseiϕs][|μ|+|ν|00|μ||ν|]121/2[eiϕdeiϕdeiϕdeiϕd],
(29)
where ϕs and ϕd were defined after Eq. (9), and ϕδ was defined after Eq. (20). Suppose that the input sideband phases are measured relative to the reference phase ϕd. Then decomposition (29) implies that the combination x1+x2* is stretched, whereas the combination x1x2* is squeezed. For stretching, the optimal phase condition is ϕ1 + ϕ2 = 0, whereas for squeezing, the optimal phase condition is ϕ1 + ϕ2 = π. These results are consistent with the results stated after Eq. (22). It follows from Eq. (18) that the backward transfer matrix has the Schmidt decomposition
Tx1(z)=121/2[ieiϕdieiϕdieiϕdieiϕd][|μ|+|ν|00|μ||ν|]eiϕδ21/2[ieiϕsieiϕsieiϕsieiϕs].
(30)
One can derive Eq. (30) from Eq. (29) by making the replacements esied, edies and eδeδ.

It only remains to rewrite Eq. (25) in the canonical form of Eq. (3). By combining Eqs. (25) and (27), one finds that the transfer matrices
M=[μ1100μ22]=[eμ00e*μ],
(31)
N=[0ν12ν210]=[0eνe*ν0].
(32)
These matrices have the Schmidt-like decompositions
M=121/2[eiϕδieiϕδeiϕδieiϕδ][μ00μ]121/2[11ii],
(33)
N=121/2[eiϕδieiϕδeiϕδieiϕδ][ν00ν]121/2[11ii].
(34)
It is the difference between the last matrices in Eqs. (33) and (34) that allows M to be diagonal and N to be off-diagonal. Decompositions (33) and (34) are not quite in canonical form, because the elements of the diagonal matrices (μ and ν) are complex. However, by generalizing the derivation of Eq. (10), one obtains the Schmidt decompositions
M=eiϕs21/2[eiϕδieiϕδeiϕδieiϕδ][|μ|00|μ|]eiϕd21/2[11ii],
(35)
N=eiϕs21/2[eiϕδieiϕδeiϕδieiϕδ][|ν|00|ν|]eiϕd21/2[11ii].
(36)
For reference, decompositions (29), (30), (35) and (36) are not unique.

At this point, it is instructive to introduce the superposition modes
x±=(x1±x2)/21/2.
(37)
By combining Eqs. (19), one obtains the superposition-mode equations
dzx±=iδsx±+iδdx±iγx±*,
(38)
where the mismatch coefficients δs = (δ1 + δ2)/2 and δd = (δ1δ2)/2. For the special case in which δd = 0, the sum (+) and difference (−) modes evolve independently: Each mode undergoes a one-mode parametric process that is governed by Eq. (4) or (6). Because the coupling coefficients in Eqs. (38) differ by a factor of −1, the input phases required for stretching differ by π/2, as do the input phases required for squeezing. This method of analysis fails for the general case in which δd ≠ 0. However, the concept of a superposition mode remains useful.

Decompositions (35) and (36), which are valid for arbitrary values of δd, show that it is appropriate to measure the input sideband phases relative to the (common) reference phase ϕd, as did decomposition (29). With this convention, the real quadrature of the sum mode and the imaginary quadrature of the difference mode are stretched. It follows from Eq. (4), in which the transfer functions are non-negative by construction, that the imaginary sum quadrature and the real difference quadrature are squeezed.

In summary, the Schmidt decompositions (17), (29), (35) and (36) reveal automatically the input-phase conditions required for stretching and squeezing [26

26. Z. Tong, C. Lundström, P. A. Andrekson, M. Karlsson, and A. Bogris, “Ultralow noise, broadband phase-sensitive optical amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron. 18, 1016–1032 (2012). [CrossRef]

]. The decompositions of the augmented and standard transfer matrices feature stretched and squeezed modes, whereas the decompositions of the canonical matrices feature only stretched modes (from which the squeezed modes can be deduced). These decompositions also lead to the concept of superposition modes: The Schmidt modes of a system are the natural superposition modes of that system.

3. Basic theory of Schmidt decompositions

In the preceding section, the Schmidt decompositions of two simple parametric processes were determined explicitly, and were shown to have interesting and useful properties. In this section, the basic properties of such decompositions are established for arbitrary parametric processes. The evolution of a conservative system of m coupled modes is governed by the Hamiltonian
H=XJX+XKX*+XtK*X,
(39)
where X is an m × 1 mode-amplitude vector, and J and K are m × m coefficient matrices. In order for the first term on the right side of Eq. (39) to be real, J must be Hermitian. The sum of the second and third terms is real by construction, and one can always write these terms in such a way that K is symmetric. For reference, J is specified by (up to) m2 real parameters, whereas K is specified by m(m + 1) real parameters. By applying the (complex) Hamilton equation
dzX=iH/X
(40)
to Hamiltonian (39), one obtains the CME
dzX=iJX+iKX*.
(41)
(The complex Hamiltonian formalism is reviewed in the Appendix.) Equation (41) depends linearly on X and X*, so its solution can be written in the IO form
X(z)=M(z)X(0)+N(z)X*(0),
(42)
where M and N are m × m transfer matrices. [Equation (41) is just Eq. (1), with A = iJ and B = iK, and Eq. (42) is just Eq. (2), repeated for convenience.] For the special case in which m = 1, J = δ and K = γ, and Eq. (41) reduces to Eq. (5).

Alternatively, one can write Eq. (41) and its conjugate as the augmented matrix equation
dzY=iLyY,
(43)
where the 2m × 1 mode vector and 2m × 2m coefficient matrix are
Y=[XX*],Ly=[JKK*J*],
(44)
respectively. The solution of Eq. (43) can be written in the IO form
Y(z)=Ty(z)Y(0),
(45)
where Ty is the 2m × 2m transmission matrix. If Ly is a constant matrix, then
Ty(z)=exp(iLyz).
(46)
Because Eq. (43) describes two copies of the same process (the original and its conjugate), the m × m blocks of Ty are the transfer matrices M and N, which appeared in Eq. (42), and their conjugates [see Eq. (15)]. Clearly, Eqs. (43)(46) are generalizations of Eqs. (11)(14).

In general, each component of X is coupled to every component of X and X*. However, there are many important systems in which a subset of the components of X (denoted by X1 and called the signal vector) is coupled to itself and a different subset of X* (denoted by X2* and called the idler vector). Such systems, which include the MI- and PC-based systems described in the introduction, are governed by the Hamiltonian
H=X1J1X1+X2J2X2+X1KX2*+X1tK*X2,
(47)
where J1, J2 and K are coefficient matrices. J1 and J2 are Hermitian, whereas K is arbitrary. For definiteness, suppose that X1 and X2 are n × 1 vectors, where 2nm, so J1, J2 and K are n × n matrices. Then J1 and J2 are each specified by (up to) n2 real parameters, whereas K is specified by 2n2 real parameters. By applying the Hamilton equations
dzXj=iH/Xj
(48)
to Hamiltonian (47), one obtains the CMEs
dzX1=iJ1X1+iKX2*,dzX2=iJ2X2+iKtX1*.
(49)
The solutions of Eqs. (49) can be written in the IO forms
X1(z)=M11(z)X1(0)+N12(z)X2*(0),X2(z)=M22(z)X2(0)+N21(z)X1*(0),
(50)
where M11, N12, M22 and N21 are n × n transfer matrices. For the special case in which n = 1, J1 = δ1, J2 = δ2 and K = γ, and Eqs. (49) reduce to Eqs. (19).

Alternatively, Eqs. (49) can be rewritten as the standard matrix equation
dzX=iLxX,
(51)
where the 2n × 1 mode vector and 2n × 2n coefficient matrix are
X=[X1X2*],Lx=[J1KKJ2*],
(52)
respectively. The solution of Eq. (51) can be written in the IO form
X(z)=Tx(z)X(0),
(53)
where Tx is the 2n × 2n transmission matrix. If Lx is a constant matrix, then
Tx(z)=exp(iLxz).
(54)
The n × n blocks of Tx are M11, N12, N21* and M22* [see Eq. (27)]. Clearly, Eqs. (51)(54) are generalizations of Eqs. (23)(26).

Although the special system described by Eq. (51) is of lower order that the general system described by Eq. (43), its mathematical structure is more complicated, because the diagonal blocks of Lx are not necessarily equal and the off-diagonal blocks are not necessarily symmetric, as are the corresponding blocks of Ly [Eqs. (44) and (52)]. Hence, we will derive the properties of the special system and deduce the corresponding properties of the general system.

In two-mode amplification, the Manley-Rowe-Weiss (MRW) variable c = |x1|2 − |x2|2 is conserved [33

33. J. M. Manley and H. E. Rowe, “Some general properties of nonlinear elements—Part I. General energy relations,” Proc. IRE 44, 904–913 (1956). [CrossRef]

, 34

34. M. T. Weiss, “Quantum derivation of energy relations analogous to those for nonlinear reactances,” Proc. IRE 45, 1012–1013 (1957).

]. The classical (quantal) interpretation of this result is that the difference between the action (photon) fluxes of the signal and idler is conserved (sideband photons are produced in pairs). In the present context, the MRW variable C=X1X1X2tX2*=XSX. By combining this definition with Eq. (51), one finds that
dzC=iX(SLxLxS)X.
(57)
Thus, Eq. (56) is both a necessary and sufficient condition for the conservation of action (photon) flux.

The Schmidt decomposition theorem [16

16. G. J. Gbur, Mathematical Methods for Optical Physics and Engineering (Cambridge, 2011), Sec. 5.4.

] states that the transfer matrix Tx = VDU, where U and V are unitary, and D is positive and diagonal. (Positivity is required because Tx is invertible.) The columns of U (input Schmidt vectors) are the eigenvectors of TxTx, the columns of V (output Schmidt vectors) are the eigenvectors of TxTx and the entries of D (Schmidt coefficients) are the square roots of the (common) eigenvalues of TxTx and TxTx. The stated decomposition is not unique: One can multiply the columns of U and V by the same set of phase factors, and one can permute (reorder) the columns of U and V in the same way, without invalidating the decomposition. Notice that the input vectors of Tx are the eigenvectors of TxTx and the output vectors of Tx are the eigenvectors of TxTx. Hence, the input (output) vectors of Tx are the output (input) vectors of Tx and the Schmidt coefficients of Tx equal those of Tx.

Before determining the Schmidt decomposition in its entirety, we pause to prove an important result about the Schmidt coefficients. For any matrix A,
det(λIA)=det[S(λIA)S]=det(λISAS).
(61)
Hence, pre- and post-multiplying a matrix by S does not change its eigenvalues. (Furthermore, if E is an eigenvector of A, then SE is an eigenvector of SAS.) For any invertible matrix A and any other matrix B,
det(λIAB)=det[A1(λIAB)A]=det(λIBA).
(62)
Hence, interchanging two matrices does not change the eigenvalues of their product. Equation (62) implies that TxTx and TxTx have the same (positive) eigenvalues, and the identity Tx1=STxS implies that
(TxTx)1=Tx1(Tx)1=(STxS)(STxS)=S(TxTx)S,
(63)
(TxTx)1=(Tx)1Tx1=(STxS)(STxS)=S(TxTx)S.
(64)
Hence, if λ is an eigenvalue of TxTx or TxTx, so also is λ−1. (Furthermore, if E is the eigenvector associated with λ, then SE is the eigenvector associated with λ−1.) Because the Schmidt coefficients of Tx are the square roots of these eigenvalues, they always occur in reciprocal pairs, as they did in Eqs. (17) and (29).

We now return to the Schmidt decomposition. Equations (59) and (60) impose several constraints on the block matrices M11, N12, N21 and M22, which allow the decomposition to be determined. The former equation implies that
Tx(z)=[M11(z)N12(z)N21*(z)M22*(z)]=STx(z)S=[M11(z)N21t(z)N12(z)M22t(z)].
(65)
By equating the blocks in Eq. (65), one finds that
M11(z)=M11(z),N12(z)=N21t(z),M22(z)=M22(z),N21(z)=N12t(z).
(66)
Notice that Eqs. (66) reduce to Eqs. (9) and (22) in the appropriate limits. The latter equation implies that
Tx1(z)Tx(z)=[M11N21tN12M22t][M11N12N21*M22*]=[M11M11N21tN21*M11N12N21tM22*M22tN21*N12M11M22tM22*N12N12]=[I00I],
(67)
Tx(z)Tx1(z)=[M11N12N21*M22*][M11N21tN12M22t]=[M11M11N12N12N12M22tM11N21tN21*M11M22*N12M22*M22tN21*N21t]=[I00I].
(68)
The block matrices have the Schmidt decompositions M11=V11D11U11, N12=V12D12U12, N21=V21D21U21 and M22=V22D22U22. By substituting these decompositions in the diagonal terms in Eqs. (67) and (68), one finds that
U11D112U11U21*D212U21t=I,
(69)
U22*D222U11tU12D122U12=I,
(70)
V11D112V11V12D122V12=I,
(71)
V22*D222V22tV21*D212V21t=I.
(72)
Equation (69) can be rewritten as U11(D112I)U11=U21*D212U21t Because the unitary decomposition of a Hermitian matrix is unique (apart from phase shifts and ordering), U21*=U11=U1 and D112D212=I. By continuing in this way, one finds that U12=U22*=U2* and D222D122=I, V12 = V11 = V1 and D112D122=I, and V21*=V22*=V2* and D222D212=I. Hence, D11 = D22 = Dμ and D12 = D21 = Dν, where Dμ2Dν2=I. The equations associated with the off-diagonal terms in Eqs. (67) and (68) are satisfied identically. By assembling the preceding results, one finds that [35

35. A. I. Lvovsky, W. Wasilewski, and K. Banaszek, “Decomposing a pulsed optical parametric amplifier into independent squeezers,” J. Mod. Opt. 54, 721–733 (2007). [CrossRef]

, 36

36. C. J. McKinstrie, M. G. Raymer, and H. J. McGuinness, “Spatial-temporal evolution of asymmetrically-pumped phase conjugation I: General formalism,” Alcatel-Lucent ITD-09-48636Q, available upon request.

]
Tx(z)=[V1DμU1V1DνU2tV2*DνU1V2*DμU2t].
(73)
Notice that Eq. (73) does not define the overall phases of the blocks uniquely. If one were to replace Uj and Vj by Ujej and Vjej, respectively, the diagonal blocks of the transfer matrix would be unaltered, whereas the off-diagonal blocks would be multiplied by ei(ϕ1+ϕ2) and ei(ϕ1+ϕ2). One can exploit this non-uniqueness to write some decompositions in particularly simple ways.

For each of the n sets of Schmidt modes in decomposition (73), there are two Schmidt coefficients (|μ| and |ν|) and four phase combinations (ϕv1ϕu1, ϕv1 + ϕu2, − ϕv2ϕu1 and −ϕv2 + ϕu2). However, only one of the coefficients is independent (|ν|) and only three of the combinations are independent. If one defines the reference phase (ϕv1ϕv2ϕu1 + ϕu2)/2, then the first and fourth combinations have the relative phases ±(ϕv1 + ϕv2ϕu1ϕu2)/2, whereas the second and third combinations have the relative phases ±(ϕv1 + ϕv2 + ϕu1 + ϕu2)/2. Thus, if the Schmidt modes are known, only 4n real parameters are required to specify the transfer matrix. [The other 4n(n − 1) parameters in the coefficient matrix specify the Schmidt modes.] For the special case in which J1 = J2 and K = Kt, the signal and idler equations are identical, so U1 = U2, V1 = V2 and the reference phase is 0. Thus, if the Schmidt modes are known, only 3n real parameters are required to specify the transfer matrix. [The other 2n(n − 1) parameters in the coefficient matrix specify the Schmidt modes.] These results are consistent with Eqs. (15) and (27), which apply to cases in which n = 1.

According to the Schmidt decomposition theorem, the Schmidt coefficients and input Schmidt modes are determined by the eigenvalues and eigenvectors of the Hermitian matrix
TxTx=[U1(Dν2+Dν2)U1U1(2DμDν)U2tU2*(2DμDν)U1U2*(Dμ2+Dν2)U2t].
(74)
Short calculations show that this matrix has the eigensystem (written in block form)
D±=(Dμ±Dν)2,E±=[U1t,±U2]t/21/2.
(75)
Similarly, the coefficients and output modes are determined by the eigenvalues and eigenvectors of the Hermitian matrix
TxTx=[V1(Dν2+Dν2)V1V1(2DμDν)V2tV2*(2DμDν)V1V2*(Dμ2+Dν2)V2t].
(76)
Short calculations show that this matrix has the eigensystem
D±=(Dμ±Dν)2,E±=[V1t,±V2]t/21/2.
(77)
Notice that in both eigensystems E = SE±, as was predicted after Eq. (64). Hence, if E± are the eigenvectors of TxTx or TxTx, then SE± = E are the eigenvectors of STxTxS or STxTxS. It follows from Eqs. (75) and (77) that the transfer matrix (73) has the Schmidt decomposition
Tx(z)=121/2[V1V1V2*V2*][Dμ+Dν00DμDν]121/2[U1U2tU1U2],
(78)
where Uj, Vj, Dμ and Dν depend implicitly on z. Let X1 = U1X̄1 and X2 = U2X̄2, where 1 = [x̄1j]t and X̄2 = [x̄2j]t. Then decomposition (78) implies that the (input) combinations x¯1jx¯2j* are stretched, whereas the combinations x¯1jx¯2j* are squeezed. Notice that Eq. (78) reduces to Eqs. (17) and (29) in the appropriate limits.

If the forward transfer matrix can be written in the form (73), the backward transfer matrix can be written in the form
Tx1(z)=[U1DμV1U1DνV2tU2*DνV1U2*DμV2t].
(79)
One could obtain Eq. (79) from Eq. (73) by interchanging the input and output modes, and changing the sign of Dν. However, Dν is non-negative by construction, so this empirical rule is not canonical.

It is instructive to consider the decomposition of the backward matrix, which can be calculated in (at least) three ways. First, Tx = VDU, so the laws of matrix algebra require that Tx1=UD1V. Hence,
Tx1(z)=121/2[U1U1U2*U2*][DμDν00Dμ+Dν]121/2[V1V2tV1V2t].
(80)
In decomposition (80), which is consistent with Eq. (79), the input and output modes are interchanged, as are the associated stretching and squeezing factors.

Second, the identity Tx1=STxS requires that
Tx1(z)=121/2[U1U1U2*U2*][Dμ+Dν00DμDν]121/2[V1V2tV1V2t].
(81)
In decomposition (81), which is also consistent with Eq. (79), the input and output modes are interchanged, as are the associated stretched and squeezed modes, but the stretching and squeezing factors are unaltered.

Decompositions (80) and (81) are equivalent only because Schmidt decompositions are not unique! To explore this issue, it is useful to define the permutation matrix
P=[0II0].
(82)
When P acts to the right on a matrix, it interchanges the row blocks of that matrix, and when P acts to the left, it interchanges the column blocks. It is easy to verify that
D1=PDP.
(83)
Decomposition (80) works by permuting the Schmidt coefficients so that stretched modes are squeezed and vice versa. It is also easy to verify that
SU=UP,VS=PV.
(84)
Decomposition (81) works by permuting the Schmidt vectors so that squeezed modes are stretched and vice versa. These actions are equivalent ways to obtain the same result: In the (common) inversion formula Tx1=UPDPV, the P matrices can act to the middle, or to the outsides.

Third, the identity Tx(z)=Tx1(z) implies that there are simple relations between the forward and backward Schmidt coefficients and vectors. Equations (73) and (79) do not determine these relations uniquely. However, by applying the transformations
Dμ(z)Dμ(z),Dν(z)Dν(z),Uj(z)iVj(z),Vj(z)iUj(z)
(85)
to Eq. (73), one does obtain Eq. (79). Transformations (85) allow Dμ and Dν to remain positive. By applying them twice, one finds that Uj(z) → iVj(−z) → i2Uj(z) and Vj(z) → iUj(−z) → i2Vj(z). These results are acceptable in the context of a Schmidt decomposition, because the signs (phases) of the Schmidt modes are not unique. By applying transformations (85) to the forward decomposition (78), one obtains a backward decomposition that is similar to decomposition (81): The first unitary matrix is multiplied by i and the second is multiplied by −i, so the decompositions are equivalent. Notice that transformations (85) are consistent with Eqs. (17) and (18), and Eqs. (29) and (30). In the former case u1 = u2 = ed and v1 = v2 = es, whereas in the latter case u1 = u2 = ed, v1 = es+δ and v2 = esδ.

It only remains to rewrite Eq. (73) in the canonical form of Eq. (3). The transfer matrices are
M=[M1100M22],N=[0N12N210],
(86)
where M11 and N12 are the upper blocks of Tx and N21 and M22 are the conjugates of the lower blocks. It is easy to verify that
M=121/2[V1iV1V2iV2][Dμ00Dμ]121/2[U1U2iU1iU2],
(87)
N=121/2[V1iV1V2iV2][Dν00Dν]121/2[U1tU2tiU1tiU2t].
(88)
Relative to the basis vectors contained in U1 and U2, the real quadratures of the sum modes and the imaginary quadratures of the difference modes are stretched. It follows from Eq. (4), in which the transfer functions are non-negative by construction, that the imaginary sum quadratures and the real difference quadratures are squeezed. These results are valid for the general case in which J1J2 and KKt. Notice that Eqs. (87) and (88) reduce to Eqs. (35) and (36) in the appropriate limit.

4. Unifying principles

The specific results derived in the preceding section are closely connected to the general properties of Hamiltonian dynamical systems [37

37. H. Goldstein, Classical Mechanics, 2nd Ed. (Addison-Wesley, 1980).

,38

38. V. I. Arnold, Mathematical Methods of Classical Mechanics, 2nd Ed. (Springer, 2000).

]. To make these connections, one must rewrite the CMEs (49) in the simplest form possible. By using the fact that J2 is Hermitian, one can rewrite Hamiltonian (47) in the alternative form
H=X1J1X1+X2tJ2*X2*+X1KX2*+X2tKX1,
(89)
and by applying the alternative Hamilton equations
dzX1=iH/X1,dzX2*=iH/X2t
(90)
to Hamiltonian (89), one can reproduce the aforementioned CMEs. Equations (89) and (90) are equivalent to the Hamiltonian
H=XGX
(91)
and the single Hamilton equation
dzX=iSH/X,
(92)
where the mode vector and coefficient matrix are
X=[X1X2*],G=[J1KKJ2*],
(93)
respectively, and the spin matrix S was defined in Eq. (55). Notice that G is Hermitian. (Consequently, if L = SG, then SL = LS.) Equations (91) and (92) are said to be in canonical form.

Suppose that X′ = TX, where T is an arbitrary transformation (change-of-variables) matrix. Then, in component form,
dzxi=jTijdzxj=ijkTijSjkH/xk*=ijkiTijSjkTlk*H/xl*.
(94)
A transformation is said to be canonical (symplectic) if the equation for xi has the same Hamiltonian form as the equation for xi. Equation (94) shows that T is symplectic if and only if
jklTijSjkTlk*=Sil.
(95)
Condition (95) can be rewritten in the matrix form
TST=S.
(96)
If condition (96) is satisfied, then T−1 = STS and (T)−1 = STS. For reference, the set of (nonsingular) matrices that satisfy condition (96) form a group with respect to multiplication.

Now suppose that T(z) is the transfer matrix for the system, which satisfies the evolution equation
dzT=iSGT,
(97)
together with the input condition T (0) = I. Then
dz(TST)=i(TS2GTTGS2T)=0,
(98)
because G is Hermitian. Hence,
TST=S.
(99)
Equation (99) implies that
(TX)S(TX)=X(TST)X=XSX.
(100)
Hence, the MRW variable XSX is conserved. Equation (99) also implies that T−1 = STS and (T)−1 = STS.

The transfer matrix must also satisfy the symplectic condition (96), because the Hamilton equation (92) retains its form as X evolves. One can prove this statement directly. Alternatively, by multiplying the identity TST = S by S(T)−1 on the left and T−1S on the right, one can show that S = (TS)S(ST) = TST. Hence, the transfer matrix satisfies the symplectic condition, which is equivalent to the MRW condition (99). Notice that the proofs of the preceding results were based on the assumption that T is a linear transformation, but not on the assumption that G is a constant: The results remain valid when G is a function of z.

The key results of the preceding section are consequences of the identity T−1 = STS. For example, if (TT)E = λE, where λ ≠ 0, then
(TT)SE=S(STS)(STS)E=S(TT)1E=λ1SE.
(101)
Thus, not only do the Schmidt coefficients occur in reciprocal pairs, but there is also a simple relation between the associated Schmidt vectors. The Schmidt decomposition (73) owes its form to the constraints imposed on the blocks of the transfer matrix [Eqs. (67) and (68)], which are just (STS)T = I and T(STS) = I. Decompositions (78), (79), (81), (87) and (88) all follow directly from decomposition (73). Furthermore, if T = VDU, then
T1=S(UDV)S=(SU)D(SV).
(102)
It is always true that the input (output) vectors of T are the output (input) vectors of T. Equations (101) and (102) show that the stretched modes for the forward transformation are the squeezed modes for the backward transformation. These relationships guarantee that the combined transformation is the identity transformation.

The preceding results show that the Schmidt decomposition of the transfer matrix owes its (extremely useful) form to the symplectic properties of the associated evolution equation, which is Hamiltonian. There are many physical processes for which knowledge of the underlying mathematical (algebraic) structure facilitates the derivation of important physical results. Consequently, it is worthwhile to review some definitions and make some specific connections.

Nonsingular matrices form a variety of groups with respect to multiplication. GL(m,C) is the general linear group, whose members are m × m complex matrices. SL(m,C) is the special linear group of degree m, whose members are unimodular (have determinant 1).

U(m) is the unitary group, whose members are m × m unitary matrices (which are complex by definition). The actions of these matrices preserve the quadratic form XX, where X is an arbitrary m × 1 vector. SU(m) is the special unitary group of degree m, whose members have determinant 1. It is sometimes called the unimodular unitary group. These groups occur in models of conservative phenomena, in which XX is the total power (or energy). For example, U(2) and SU(2) underly polarization rotation, beam splitting, directional coupling and (stable) frequency conversion.

U(n,n) is the pseudo-unitary group, whose members are 2n × 2n complex matrices. The actions of these matrices preserve the quadratic form XSX, where X is an arbitrary 2n × 1 vector. (This group has a subgroup of diagonal matrices that are unitary, but most of its members are nonunitary.) In the context of parametric amplification, XSX is the MRW variable and Tx is a member of U(n,n). SU(n,n) is the special pseudo-unitary group of degree 2n, whose members have determinant 1. (This group also has a unitary subgroup.) In the aforementioned context, Ty is a member of SU(n,n). Ty is also a member of the symplectic group Sp(2n), whose members satisfy condition (96) and have determinant 1. (In the Appendix, it is shown that this definition of the symplectic group is equivalent to the standard definition, which involves a different auxiliary matrix.) The mathematical properties of continuous groups are described in [39

39. D. H. Sattinger and O. L. Weaver, Lie groups and Algebras with Applications to Physics, Geometry and Mechanics (Springer, 1986). [CrossRef]

, 40

40. M. Hamermesh, Group Theory and its Application to Physical Problems (Dover, 1989).

] and several simple examples (including those mentioned above) are described in [41

41. H. Takenaka, “A unified formalism for polarization optics by using group theory,” Nou. Rev. Opt. 4, 37–41 (1973). [CrossRef]

44

44. C. C. Gerry, “Remarks on the use of group theory in quantum optics,” Opt. Express 8, 76–85 (2001). [CrossRef] [PubMed]

].

5. Summary

Parametric devices based on four-wave mixing in fibers perform many signal-processing functions required by optical communication systems. In these devices, strong pumps drive weak signal and idler sidebands, which can have one or two polarization components, and one or many frequency components. The evolutions of these components (modes) are governed by systems of coupled-mode equations [Eq. (1)], the solutions of which are specified by transfer matrices [Eq. (2)]. Schmidt decompositions of these transfer matrices [Eq. (3)] determine the natural input and output mode vectors of such systems, and the effects of the systems on the associated mode amplitudes [Eqs. (4)].

In Sec. 2, two simple examples were considered: one- and two-mode parametric amplification. The transfer matrices for these processes were determined explicitly, as were their Schmidt decompositions. In the first process, different quadratures of the same mode are stretched and squeezed (dilated), whereas in the second process, different combinations (superpositions) of the signal and idler modes are dilated. These superpositions are the Schmidt modes. For every mode (quadrature) that is stretched, there is another mode (quadrature) that is squeezed by the same amount. Furthermore, simple relations exist between the transfer matrices for the forward and backward processes [Eqs. (27) and (28)], and their Schmidt decompositions [Eqs. (29) and (30)]. These properties turn out to be generic.

In Sec. 4, the specific properties established for parametric processes were shown to be general properties of Hamiltonian dynamical systems. The symplectic identity [Eq. (96)] was shown to be equivalent to the Manley-Rowe-Weiss identity [Eq. (99)]. Either identity (together with the laws of matrix algebra) is sufficient to establish the main properties of Schmidt decompositions: For every Schmidt mode that is stretched in a forward (or backward) transformation, there is another, closely related, mode that is squeezed by the same amount. Furthermore, the forward and backward transformations have the same Schmidt coefficients, and the input squeezed (stretched) modes for the backward transformation are the output stretched (squeezed) modes for the forward transformation. These relationships guarantee that the combined transformation is the identity transformation.

Not only do Schmidt decompositions provide physical insight into complicated parametric processes, they also provide the mathematical means to optimize the performance of devices based on these processes. In a forthcoming paper, several nontrivial examples relevant to current research (including four-mode parametric amplification) will be discussed in detail.

Appendix: Real and complex Hamiltonian systems

Consider a real dynamical system with Hamiltonian H(q, p), where q and p are conjugate variables. The Hamilton equations (for z-evolution) are
dzp=H/q,dzq=H/p.
(103)
By defining the vector variable X = [x1, x2]t = [p,q]t, one can rewrite Eqs. (103) in the matrix form
dzX=JH/X,
(104)
where the auxiliary matrix
J=[0110]
(105)
and the derivative of H is taken componentwise. Notice that J2 = −I.

Now consider the quadratic Hamiltonian
H=α¯p2/2+β¯pq+γ¯q2/2,
(106)
where ᾱ, β̄ and γ̄ are real constants (parameters). By applying the Hamilton equations (103) to Hamiltonian (106), one obtains the linear evolution equations
dzp=β¯p+γ¯q,dzq=α¯pβ¯q.
(107)
Alternatively, one can rewrite Hamiltonian (106) in the compact form
H=XtGX/2,
(108)
where the coefficient matrix
G=[α¯β¯β¯γ¯].
(109)
Notice that G is symmetric. By applying the Hamilton equation (104) to Hamiltonian (108), one obtains the matrix evolution equation
dzX=JGX,
(110)
which is equivalent to the component equations (107).

Suppose that X′ = TX, where T is an arbitrary transformation (change-of-variables) matrix. Then, in component form,
dzxi=jTijdzxj=jkTijJjkH/xk=jklTijJjkTlkH/xl.
(111)
Hence, the transformation is symplectic if and only if
jklTijJjkTlk=Jil.
(112)
This condition can be rewritten in the matrix form
TJTt=J.
(113)
If condition (113) is satisfied, then T−1 = −JTtJ and (Tt)−1 = −JTJ.

Now suppose that T(z) is the transfer matrix for the system, which satisfies the evolution equation
dzT=JGT,
(114)
together with the input condition T(0) = I. Then
dz(TtJT)=TtJ2GT+TtGtJtJT=0,
(115)
because Jt = −J, J2 = −I and G = Gt. Hence,
TtJT=J.
(116)
Conditions (113) and (116) are equivalent, and require that det(T) = ±1. However, G is symmetric, so tr(JG) = 0 and det(T) = 1. Hence, T is a member of Sp(2,R), the three-parameter group whose members are symplectic 2 × 2 matrices with determinant 1. [This group is isomororphic to SL(2,R) and SU(1,1).]

To establish the connection between real and complex dynamical systems, one defines the complex variables
a=(q+ip)/21/2,a*=(qip)/21/2,
(117)
in which case
p=i(a*a)/21/2,q=(a*+a)/21/2.
(118)
The Hamiltonian H(a,a*) = H[q(a,a*), p(a,a*)], where a and a* are treated as independent variables. By combining Eqs. (103), (117) and (118), one obtains the complex Hamilton equations
dza=iH/a*,dza*=iH/a.
(119)
By defining the vector variable X = [a,a*]t, one can rewrite Eqs. (119) in the matrix form
dzX=iSH/X*,
(120)
where the auxiliary matrix S was defined in Eq. (55) and the derivative of H is taken componentwise. Equation (120) is equivalent to Eq. (104).

Now consider the quadratic Hamiltonian
H=δ|a|2+γ(a*)2/2+γ*a2/2,
(121)
which also involves three (real) parameters (δ, γr and γi). By applying the Hamilton equations (119) to Hamiltonian (121), one obtains the linear evolution equations
dza=iδa+iγa*,dza*=iδa*iγ*a.
(122)
The first of Eqs. (122) is just Eq. (5), which describes one-mode squeezing. One can reconcile the real and complex descriptions of this process by setting
α¯=δγr,β¯=γi,γ¯=δ+γr.
(123)
Alternatively, one can rewrite Hamiltonian (121) in the compact form
H=XGX/2,
(124)
where the coefficient matrix
G=[δγγ*δ].
(125)
Notice that G is Hermitian. By applying the Hamilton equation (120) to Hamiltonian (124) componentwise, and reassembling the results, one obtains the matrix evolution equation
dzX=iSGX.
(126)
However, it is easier to rewrite Hamiltonian (124) without the factor of 2, and differentiate it with respect to X, treated as a single (vector) variable. This approach was taken in the main text.

Because Eqs. (110) and (126) are equivalent descriptions of the same system, their consequences, prime among which are the symplectic identities, must also be equivalent. To prove this result explicitly, define Xr = [p, q]t and Xc = [a, a*]t. Then Xc = UXr, where the unitary matrix
U=121/2[i1i1].
(127)
It follows from these definitions that if Xr (z) = Tr(z)Xr (0), then Xc(z) = UTr(z)UXc(0), so Tc(z) = UTr(z)U. By multiplying Eq. (113) by U on the left and U on the right, one finds that
U(TrJTrt)U=(UTrU)(UJU)(UTrtU)=UJU.
(128)
But UJU = iS, so Eq. (128) is just Eq. (96). Thus, the real and complex symplectic identities are equivalent, so the groups formed by Tr and Tc [Sp(2,R) and SU(1,1)] are isomorphic. This equivalence extends to systems of 2n variables.

References and links

1.

M. E. Marhic, Fiber Optical Parametric Amplifiers, Oscillators and Related Devices (Cambridge, 2007). [CrossRef]

2.

J. Hansryd, P. A. Andrekson, M. Westlund, J. Li, and P. O. Hedekvist, “Fiber-based optical parametric amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron. 8, 506–520 (2002). [CrossRef]

3.

J. H. Lee, “All-optical signal processing devices based on holey fiber,” IEICE Trans. Electron. E88-C, 327–334 (2005). [CrossRef]

4.

S. Radic and C. J. McKinstrie, “Optical amplification and signal processing in highly nonlinear optical fiber,” IEICE Trans. Electron. E88-C, 859–869 (2005). [CrossRef]

5.

P. A. Andrekson and M. Westlund, “Nonlinear optical fiber based high resolution all-optical waveform sampling,” Laser Photon. Rev. 1, 231–248 (2007). [CrossRef]

6.

S. Radic, “Parametric signal processing,” IEEE J. Sel. Top. Quantum Electron. 18, 670–680 (2012). [CrossRef]

7.

R. Loudon, The Quantum Theory of Light, 3rd Ed. (Oxford, 2000).

8.

M. Fiorentino, P. L. Voss, J. E. Sharping, and P. Kumar, “All-fiber photon-pair source for quantum communications,” IEEE Photon. Technol. Lett. 14, 983–985 (2002). [CrossRef]

9.

M. Halder, J. Fulconis, B. Cemlyn, A. Clark, C. Xiong, W. J. Wadsworth, and J. G. Rarity, “Nonclassical 2-photon interference with separate intrinsically narrowband fibre sources,” Opt. Express 17, 4670–4676 (2009). [CrossRef] [PubMed]

10.

O. Cohen, J. S. Lundeen, B. J. Smith, G. Puentes, P. J. Mosley, and I. A. Walmsley, “Tailored photon-pair generation in optical fibers,” Phys. Rev. Lett. 102, 123603 (2009). [CrossRef] [PubMed]

11.

K. Inoue, “Polarization effect on four-wave mixing efficiency in a single-mode fiber,” IEEE J. Quantum Electron. 28, 883–894 (1992). [CrossRef]

12.

C. J. McKinstrie, H. Kogelnik, R. M. Jopson, S. Radic, and A. V. Kanaev, “Four-wave mixing in fibers with random birefringence,” Opt. Express 12, 2033–2055 (2004). [CrossRef] [PubMed]

13.

M. E. Marhic, K. K. Y. Wong, and L. G. Kazovsky, “Fiber optic parametric amplifiers with lineary or circularly polarized waves,” J. Opt. Soc. Am. B 20, 2425–2433 (2003). [CrossRef]

14.

C. J. McKinstrie, H. Kogelnik, and L. Schenato, “Four-wave mixing in a rapidly-spun fiber,” Opt. Express 14, 8516–8534 (2006). [CrossRef] [PubMed]

15.

G. W. Stewart, “On the early history of the singular value decomposition,” SIAM Rev. 35, 551–566 (1993). [CrossRef]

16.

G. J. Gbur, Mathematical Methods for Optical Physics and Engineering (Cambridge, 2011), Sec. 5.4.

17.

A. K. Ekert and P. L. Knight, “Relationship between semiclassical and quantum-mechanical input-output theories of optical response,” Phys. Rev. A 43, 3934–3938 (1991). [CrossRef] [PubMed]

18.

S. L. Braunstein, “Squeezing as an irreducible resource,” Phys. Rev. A 71, 055801 (2005). [CrossRef]

19.

H. P. Yuen, “Two-photon states of the radiation field,” Phys. Rev. A 13, 2226–2243 (1976). [CrossRef]

20.

C. M. Caves, “Quantum limits on noise in linear amplifiers,” Phys. Rev. D 26, 1817–1839 (1982). [CrossRef]

21.

C. K. Law, I. A. Walmsley, and J. H. Eberly, “Continuous frequency entanglement: Effective finite Hilbert space and entropy control,” Phys. Rev. Lett. 84, 5304–5307 (2000). [CrossRef] [PubMed]

22.

W. P. Grice, A. B. U’Ren, and I. A. Walmsley, “Eliminating frequency and space-time correlations in multiphoton states,” Phys. Rev. A 64, 063815 (2001). [CrossRef]

23.

M. G. Raymer, S. J. van Enk, C. J. McKinstrie, and H. J. McGuinness, “Interference of two photons of different color,” Opt. Commun. 283, 747–752 (2010). [CrossRef]

24.

C. J. McKinstrie, L. Mejling, M. G. Raymer, and K. Rottwitt, “Quantum-state-preserving optical pulse reshaping and multiplexing by four-wave mixing in fibers,” Phys. Rev. A 85, 053829 (2012). [CrossRef]

25.

C. J. McKinstrie, M. Yu, M. G. Raymer, and S. Radic, “Quantum noise properties of parametric processes,” Opt. Express 13, 4986–5012 (2005). [CrossRef] [PubMed]

26.

Z. Tong, C. Lundström, P. A. Andrekson, M. Karlsson, and A. Bogris, “Ultralow noise, broadband phase-sensitive optical amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron. 18, 1016–1032 (2012). [CrossRef]

27.

E. Telatar, “Capacity of multi-antenna Gaussian channels,” Euro. Trans. Telecom. 10, 585–595 (1999). [CrossRef]

28.

C. J. McKinstrie and N. Alic, “Information efficiencies of parametric devices,” J. Sel. Top. Quantum Electron. 18, 794–811 (2012). [CrossRef]

29.

C. J. McKinstrie, “Unitary and singular value decompositions of parametric processes in fibers,” Opt. Commun. 282, 583–593 (2009). [CrossRef]

30.

D. A. Edwards, J. D. Fehribach, R. O. Moore, and C. J. McKinstrie, “An application of matrix theory to the evolution of coupled modes,” to appear in SIAM Rev.

31.

C. J. McKinstrie and S. Radic, “Phase-sensitive amplification in a fiber,” Opt. Express 12, 4973–4979 (2004). [CrossRef] [PubMed]

32.

K. Croussore and G. Li, “Phase and amplitude regeneration of differential phase-shift keyed signals using phase-sensitive amplification,” IEEE J. Sel. Top. Quantum Electron. 14, 648–658 (2008). [CrossRef]

33.

J. M. Manley and H. E. Rowe, “Some general properties of nonlinear elements—Part I. General energy relations,” Proc. IRE 44, 904–913 (1956). [CrossRef]

34.

M. T. Weiss, “Quantum derivation of energy relations analogous to those for nonlinear reactances,” Proc. IRE 45, 1012–1013 (1957).

35.

A. I. Lvovsky, W. Wasilewski, and K. Banaszek, “Decomposing a pulsed optical parametric amplifier into independent squeezers,” J. Mod. Opt. 54, 721–733 (2007). [CrossRef]

36.

C. J. McKinstrie, M. G. Raymer, and H. J. McGuinness, “Spatial-temporal evolution of asymmetrically-pumped phase conjugation I: General formalism,” Alcatel-Lucent ITD-09-48636Q, available upon request.

37.

H. Goldstein, Classical Mechanics, 2nd Ed. (Addison-Wesley, 1980).

38.

V. I. Arnold, Mathematical Methods of Classical Mechanics, 2nd Ed. (Springer, 2000).

39.

D. H. Sattinger and O. L. Weaver, Lie groups and Algebras with Applications to Physics, Geometry and Mechanics (Springer, 1986). [CrossRef]

40.

M. Hamermesh, Group Theory and its Application to Physical Problems (Dover, 1989).

41.

H. Takenaka, “A unified formalism for polarization optics by using group theory,” Nou. Rev. Opt. 4, 37–41 (1973). [CrossRef]

42.

Y. S. Kim and M. E. Noz, “Illustrative examples of the symplectic group,” Am. J. Phys. 51, 368–375 (1983). [CrossRef]

43.

A. Mufti, H. A. Schmitt, and M. Sargent, “Finite-dimensional matrix representations as calculational tools in quantum optics,” Am. J. Phys. 61, 729–733 (1993). [CrossRef]

44.

C. C. Gerry, “Remarks on the use of group theory in quantum optics,” Opt. Express 8, 76–85 (2001). [CrossRef] [PubMed]

OCIS Codes
(060.2320) Fiber optics and optical communications : Fiber optics amplifiers and oscillators
(190.4380) Nonlinear optics : Nonlinear optics, four-wave mixing
(270.6570) Quantum optics : Squeezed states

ToC Category:
Nonlinear Optics

History
Original Manuscript: September 11, 2012
Revised Manuscript: December 5, 2012
Manuscript Accepted: December 30, 2012
Published: January 14, 2013

Citation
C. J. McKinstrie and M. Karlsson, "Schmidt decompositions of parametric processes I: Basic theory and simple examples," Opt. Express 21, 1374-1394 (2013)
http://www.opticsinfobase.org/oe/abstract.cfm?URI=oe-21-2-1374


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References

  1. M. E. Marhic, Fiber Optical Parametric Amplifiers, Oscillators and Related Devices (Cambridge, 2007). [CrossRef]
  2. J. Hansryd, P. A. Andrekson, M. Westlund, J. Li, and P. O. Hedekvist, “Fiber-based optical parametric amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron.8, 506–520 (2002). [CrossRef]
  3. J. H. Lee, “All-optical signal processing devices based on holey fiber,” IEICE Trans. Electron.E88-C, 327–334 (2005). [CrossRef]
  4. S. Radic and C. J. McKinstrie, “Optical amplification and signal processing in highly nonlinear optical fiber,” IEICE Trans. Electron.E88-C, 859–869 (2005). [CrossRef]
  5. P. A. Andrekson and M. Westlund, “Nonlinear optical fiber based high resolution all-optical waveform sampling,” Laser Photon. Rev.1, 231–248 (2007). [CrossRef]
  6. S. Radic, “Parametric signal processing,” IEEE J. Sel. Top. Quantum Electron.18, 670–680 (2012). [CrossRef]
  7. R. Loudon, The Quantum Theory of Light, 3rd Ed. (Oxford, 2000).
  8. M. Fiorentino, P. L. Voss, J. E. Sharping, and P. Kumar, “All-fiber photon-pair source for quantum communications,” IEEE Photon. Technol. Lett.14, 983–985 (2002). [CrossRef]
  9. M. Halder, J. Fulconis, B. Cemlyn, A. Clark, C. Xiong, W. J. Wadsworth, and J. G. Rarity, “Nonclassical 2-photon interference with separate intrinsically narrowband fibre sources,” Opt. Express17, 4670–4676 (2009). [CrossRef] [PubMed]
  10. O. Cohen, J. S. Lundeen, B. J. Smith, G. Puentes, P. J. Mosley, and I. A. Walmsley, “Tailored photon-pair generation in optical fibers,” Phys. Rev. Lett.102, 123603 (2009). [CrossRef] [PubMed]
  11. K. Inoue, “Polarization effect on four-wave mixing efficiency in a single-mode fiber,” IEEE J. Quantum Electron.28, 883–894 (1992). [CrossRef]
  12. C. J. McKinstrie, H. Kogelnik, R. M. Jopson, S. Radic, and A. V. Kanaev, “Four-wave mixing in fibers with random birefringence,” Opt. Express12, 2033–2055 (2004). [CrossRef] [PubMed]
  13. M. E. Marhic, K. K. Y. Wong, and L. G. Kazovsky, “Fiber optic parametric amplifiers with lineary or circularly polarized waves,” J. Opt. Soc. Am. B20, 2425–2433 (2003). [CrossRef]
  14. C. J. McKinstrie, H. Kogelnik, and L. Schenato, “Four-wave mixing in a rapidly-spun fiber,” Opt. Express14, 8516–8534 (2006). [CrossRef] [PubMed]
  15. G. W. Stewart, “On the early history of the singular value decomposition,” SIAM Rev.35, 551–566 (1993). [CrossRef]
  16. G. J. Gbur, Mathematical Methods for Optical Physics and Engineering (Cambridge, 2011), Sec. 5.4.
  17. A. K. Ekert and P. L. Knight, “Relationship between semiclassical and quantum-mechanical input-output theories of optical response,” Phys. Rev. A43, 3934–3938 (1991). [CrossRef] [PubMed]
  18. S. L. Braunstein, “Squeezing as an irreducible resource,” Phys. Rev. A71, 055801 (2005). [CrossRef]
  19. H. P. Yuen, “Two-photon states of the radiation field,” Phys. Rev. A13, 2226–2243 (1976). [CrossRef]
  20. C. M. Caves, “Quantum limits on noise in linear amplifiers,” Phys. Rev. D26, 1817–1839 (1982). [CrossRef]
  21. C. K. Law, I. A. Walmsley, and J. H. Eberly, “Continuous frequency entanglement: Effective finite Hilbert space and entropy control,” Phys. Rev. Lett.84, 5304–5307 (2000). [CrossRef] [PubMed]
  22. W. P. Grice, A. B. U’Ren, and I. A. Walmsley, “Eliminating frequency and space-time correlations in multiphoton states,” Phys. Rev. A64, 063815 (2001). [CrossRef]
  23. M. G. Raymer, S. J. van Enk, C. J. McKinstrie, and H. J. McGuinness, “Interference of two photons of different color,” Opt. Commun.283, 747–752 (2010). [CrossRef]
  24. C. J. McKinstrie, L. Mejling, M. G. Raymer, and K. Rottwitt, “Quantum-state-preserving optical pulse reshaping and multiplexing by four-wave mixing in fibers,” Phys. Rev. A85, 053829 (2012). [CrossRef]
  25. C. J. McKinstrie, M. Yu, M. G. Raymer, and S. Radic, “Quantum noise properties of parametric processes,” Opt. Express13, 4986–5012 (2005). [CrossRef] [PubMed]
  26. Z. Tong, C. Lundström, P. A. Andrekson, M. Karlsson, and A. Bogris, “Ultralow noise, broadband phase-sensitive optical amplifiers and their applications,” IEEE J. Sel. Top. Quantum Electron.18, 1016–1032 (2012). [CrossRef]
  27. E. Telatar, “Capacity of multi-antenna Gaussian channels,” Euro. Trans. Telecom.10, 585–595 (1999). [CrossRef]
  28. C. J. McKinstrie and N. Alic, “Information efficiencies of parametric devices,” J. Sel. Top. Quantum Electron.18, 794–811 (2012). [CrossRef]
  29. C. J. McKinstrie, “Unitary and singular value decompositions of parametric processes in fibers,” Opt. Commun.282, 583–593 (2009). [CrossRef]
  30. D. A. Edwards, J. D. Fehribach, R. O. Moore, and C. J. McKinstrie, “An application of matrix theory to the evolution of coupled modes,” to appear in SIAM Rev.
  31. C. J. McKinstrie and S. Radic, “Phase-sensitive amplification in a fiber,” Opt. Express12, 4973–4979 (2004). [CrossRef] [PubMed]
  32. K. Croussore and G. Li, “Phase and amplitude regeneration of differential phase-shift keyed signals using phase-sensitive amplification,” IEEE J. Sel. Top. Quantum Electron.14, 648–658 (2008). [CrossRef]
  33. J. M. Manley and H. E. Rowe, “Some general properties of nonlinear elements—Part I. General energy relations,” Proc. IRE44, 904–913 (1956). [CrossRef]
  34. M. T. Weiss, “Quantum derivation of energy relations analogous to those for nonlinear reactances,” Proc. IRE45, 1012–1013 (1957).
  35. A. I. Lvovsky, W. Wasilewski, and K. Banaszek, “Decomposing a pulsed optical parametric amplifier into independent squeezers,” J. Mod. Opt.54, 721–733 (2007). [CrossRef]
  36. C. J. McKinstrie, M. G. Raymer, and H. J. McGuinness, “Spatial-temporal evolution of asymmetrically-pumped phase conjugation I: General formalism,” Alcatel-Lucent ITD-09-48636Q, available upon request.
  37. H. Goldstein, Classical Mechanics, 2nd Ed. (Addison-Wesley, 1980).
  38. V. I. Arnold, Mathematical Methods of Classical Mechanics, 2nd Ed. (Springer, 2000).
  39. D. H. Sattinger and O. L. Weaver, Lie groups and Algebras with Applications to Physics, Geometry and Mechanics (Springer, 1986). [CrossRef]
  40. M. Hamermesh, Group Theory and its Application to Physical Problems (Dover, 1989).
  41. H. Takenaka, “A unified formalism for polarization optics by using group theory,” Nou. Rev. Opt.4, 37–41 (1973). [CrossRef]
  42. Y. S. Kim and M. E. Noz, “Illustrative examples of the symplectic group,” Am. J. Phys.51, 368–375 (1983). [CrossRef]
  43. A. Mufti, H. A. Schmitt, and M. Sargent, “Finite-dimensional matrix representations as calculational tools in quantum optics,” Am. J. Phys.61, 729–733 (1993). [CrossRef]
  44. C. C. Gerry, “Remarks on the use of group theory in quantum optics,” Opt. Express8, 76–85 (2001). [CrossRef] [PubMed]

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