Patent Application
20230267361
This disclosure generally relates to a method and system for generating photonic Fock states, and particularly for generating single photons with arbitrarily weak photonic nonlinearities.
Single-photon Fock states are critical resource in a myriad of quantum information protocols and technologies. Generation of single-photon Fock states often relies on having a large single photon nonlinearity in optical cavity. In some applications, it may be desirable to generate single photon Fock states even when the single photon nonlinearity is weak in comparison to, for example, single photon loss rate of the optical cavity.
For a more complete understanding of the invention, reference is made to the following description and accompanying drawings, in which:
and |2
;
By way of introduction, single-photon Fock states may constitute critical ingredients in a myriad of applications including but not limited to those relying on quantum information protocols and technologies. Generation of single-photon Fock states, for example, may rely on various types of photonic nonlinear interactions via other states in particular systems. In some implementations, single-photon Fock states may be generated in a cavity mode in a photonic cavity relying on a nonlinear interaction between photons in the cavity modes via other states present within the photonic cavity. The generation of single-photon Fock states essentially may be based on one of various types of single-photon blockade effect resulting from the nonlinear photonic interactions. Essentially, as a result of such blockade effect, the presence of a single photon prevents the generation of a photonic quantum state having multiphoton components. More broadly, as a result of a blockade effect, the presence of a particular photonic Fock state prevents generation of a photonic state with higher photon numbers.
An example of a nonlinear photonic cavity system for generating single-photon Fock states is shown in in the cavity energy level diagram 5 for this particular cavity mode. The photon number states or Fock states are shown as |1
, |2
, and |3
for single-photon, two-photon, and three-photon states, respectively. Due the photonic nonlinear interaction in this system, the resonant frequency between the single-photon state |1
and the two-photon state |2
may be shifted from that between the vacuum state |0
and the single-photon state |1
, as shown by 6 of
and the three-photon state |3
may be also shifted from those between the vacuum state |0
and the single-photon state |1
and between the single-photon state |1
and the two-photon state |2
, as shown by 7 in
As a result of the energy level shifts due to the photonic nonlinearity, the electromagnetic driving fields 3 having a drive frequency resonant between the vacuum state |0 and the single-photon state |1
for exciting the single-photon state |1
would be significantly off resonance with the transition from the single-photon state |1
to the two-photon state |2
and off resonance with the transition from the two-photon state |2
to the three-photon state |3
as shown by the arrows 8, under the condition that the photonic nonlinearity U is much larger than the cavity loss rate, i.e., U»K, such that the broadening of the energy-shifted two-photon state |2
and three-photon state |3
, as characterized by the photonic cavity loss rate κ, is significantly smaller than the energy level shifts characterized by U. Under such a condition, excitation of multi-photon states including the two-photon state |2
, the three-photon state |3
and other higher number photon states would be avoided and single-photon Fock state may be generated. This particular photon-blockade effect is essentially based on introducing sufficiently large nonlinearities in a spectrum of the system that causes energy shift of photon Fock states in the presence of other photonic states. Such type of photon blockade mechanism thus requires that the single-photon nonlinearities to be much larger than the loss rate and may be achieved in highly nonlinear optical cavities incorporating single atoms, quantum dots, superconducting resonators, or the like. This type of photon blockade may be out of reach in more conventional systems that exhibit only weak nonlinearities (e.g. optical micro or nanoresonators fabricated using materials with intrinsic x(3) nonlinearities).
In some implementations of this disclosure, an alternative photon blockade effect may be achieved by using photonic nonlinearities to modify matrix elements of an effective cavity driving process as opposed to introducing nonlinearities into an optical response spectrum (such as the mechanism of
Ĥ
block={tilde over (Λ)}3{circumflex over (a)}†({circumflex over (a)}†{circumflex over (a)}−r)+h.c. (1)
where the parameter r is tuned to 1 for single photon. Here a is the cavity annihilation operator, {tilde over (Λ)}3 is the amplitude of an effective nonlinear driving process.
By construction and as illustrated in the energy level diagram 9 in and one photon state |1
, but does not allow driving from |1
to the |2
photon state, as shown by the zero matrix element 10. Crucially, as this blockade is a matrix element effect, it is effective even if cavity loss is much larger than the nonlinearity U, because this mechanism does not require large energy or spectral shift of the two-photon state |1
, as shown by 11. As such, only single photon state is accessible as a result of the nonzero matrix element between the vacuum state |0
and the single-photon state |1
(shown by 12). The two photon state |2
cannot be accessed due the zero matrix element 10. Higher number photon states such as the three-photon state |3
cannot be accessed even though the matrix element between the state |2
and |3
may be non-zero (as shown by |3
because the lack of excitation of the two-photon state 12).
While the basic mechanism in the Hamiltonian construction of Equation (1) appears straightforward, it describes an unusual nonlinear driving element. Despite its exotic form, the disclosure below shows that it can be achieved using standard ingredients including but not limited to a standard Kerr-type nonlinearity (strength U), along with standard single-photon (alternatively referred to as one-photon) and two-photon (i.e. parametric) drives. Such single-photon blockade mechanism may be effective even if the Kerr nonlinearity strength U is much weaker than the cavity loss rate κ. In addition, as described in further detail below, this scheme can also be realized using three-wave mixing type (i.e. x(2)) nonlinearities rather than standard Kerr-type nonlinearities. This underlying mechanism may also be extended and used to generate more complex blockaded states including but not limited to multi-mode non-Gaussian entangled states.
The mechanism disclosed herein exhibits metastability and two distinct slow relaxation timescales. The intermediate-time dynamics can be very different from the ultimate steady state. In particular, this shorter time regime enables the near-perfect generation of Fock states (including states with highly negative Wigner functions), in a way that is robust against imperfections. The mechanism functions under a generic driven Kerr cavity system. Unlike currently explored other photon-blockade schemes, such as the so called “unconventional photon blockade” (UPB), where states with arbitrarily small g(2)(0) correlation functions can be generated using extremely weak nonlinearities, the scheme disclosed herein generated near perfect photon Fock states rather than Gaussian states that have positive-definite Wigner functions, that do not exhibit a true cut-off in their photon number distribution, and that only exhibit suppressed intensity fluctuations in the limit where the average photon number is vanishingly small.
In some implementations, this basic mechanism may be harnessed using a time-dependent protocol that generates propagating Fock states in a variety of realistic weakly-nonlinear optical systems in the lab frame. For example, truly photon-blockaded states using a cavity mode with an extremely weak Kerr nonlinearity U«κ and standard one- and two-photon drives may be generated in a displaced frame. In the displaced frame, and for ideal matching of drive amplitudes, these states may have zero population of states with two or more photons, and moreover, can have a population of the |1 Fock state that approaches one. The example time-dependent protocol thus involves transforming the system into a displaced frame, generating the Fock state in the displaced frame, and then transforming the system back into the lab frame. It can also be shown that this process is robust again modest errors in matching the two drive amplitudes appropriately.
The basic example protocol is sketched in
The end result of the example three-step protocol above is a blockaded, approximate single-photon state in resonant mode of the cavity. To turn this into a propagating single photon state, the cavity may be configured to couple to a waveguide structure or transmission line. With such coupling, the intracavity single-photon blockaded state in the resonant mode may then preferentially leak out into the waveguide structure of the transmission line after some time as an approximate Fock state in a propagating mode with an exponential profile. While such coupling may increase cavity loss rate κ, that would not be overly detrimental to the example protocol because this protocol can be effective even if the cavity nonlinearity U is much smaller than the total loss rate κ of the cavity mode.
The initial and final displacements in the example protocol above are critical for achieving the final lab-frame photon-blockaded state. As described in further detail below, these example displacements should correspond to amplitudes αb, −αb respectively, where the single-photon drive amplitudes are determined by Eq. (8). A failure to perform this ideally represents another possible imperfection that would degrade the scheme above. Even if the one-photon drive used to perform these displacements can be calibrated perfectly, the weak cavity nonlinearity U can cause errors during steps 1 and 3 of the protocol. The dominant error may include an unwanted parametric drive generated via U. In some implementations, such an error could be canceled by also applying compensating two photon drive Λ2≠0 during steps 1 and 3; this is depicted in 22 and 23 of
The one-photon and two-photon drives described above should be detuned from the resonance cavity frequency in a commensurate manner. Specifically, if the one photon drive is detuned from the one-photon resonance frequency an offset amount δ, then the two-photon drive may be detuned from the two photon resonance frequency an amount of 2δ.
The two photon drive above may be implemented using a multi-color or multi-tone (e.g., two-color or two-tone) scheme. For example, the two-photon drive may be achieved by using multiple (e.g., two) linear drives in combination. The sum of the frequencies of the multiple linear drives (e.g., two linear drives) may match the frequency needed for a target two-photon drive described above. The multiple linear drives thus combine to achieve the effect of a target two-photon drive. The frequency of each of the multiple linear drives may be selected such that they do not cause appreciable driving of the single photon resonance.
While the various implementations described here are designed to be capable of generating Fock states even when the nonlinearity in the cavity relevant to the resonance mode is weaker than the cavity loss rate, these implementations equally apply to situations where the cavity nonlinearity is relatively strong (e.g., larger than the cavity loss rate).
In
The electromagnetic wave sources 27 of
The system 24 may optionally include the D.C. electric or magnetic sources 28 and/or 29. The D.C. sources 28 and/or 29 may be used to generate and apply D.C. electric and/or magnetic fields to the cavity 25 to modify characteristics of the the cavity to effectuate a control of the resonance mode (e.g., to modify the index of fraction or other optical properties of the cavity medium).
The nonlinear cavity 25 of
The transport device 30 of
The example target Hamiltonian shown in Equation (1) appears to include somewhat exotic nonlinear drives. This section provides detailed discussion of example approaches for achieving such nonlinear drive for generation of single-photon Fock state that may be implemented in the three-step protocol described above and in the system shown in
For example, the nonlinear cavity may include a single mode of a bosonic resonator (with resonant frequency ωc, a lowering operator â) having a weak self-Kerr nonlinearity U, which is subject to both one- and two-photon drives with amplitudes Λ1 and Λ2 respectively and drive frequencies ω1 and ω2 respectively. The starting lab-frame Hamiltonian is thus (=1):
Ĥ=ω
c
â†â+U(â+â†)4+(θ1e−iω
It is assumed that the only nonlinearity in this lab-frame Hamiltonian is Kerr interaction U, which is allowed to be extremely weak, e.g., U«κ. The two-photon drive, for example, may be a standard parametric drive and can be realized without requiring a strong single-photon nonlinearity. For example, a standard approach would be to use a weak nonlinear coupling to an auxiliary pump mode that is strongly driven, or just simply apply two additional (linear) drive tones to the main cavity mode. The approach taken herein may be agnostic to the specific mechanism used to realize Λ2.
In some implementations, the one-photon and two-photon drive frequencies ω1 and ω2 may be chose to satisfy 2ω1=ω2=2(ωc−Δ) (detuned from the resonance frequency ωc by Δ). It is also assumed that ωc is the largest frequency in the system, allowing for a rotating wave approximation (RWA) on both the nonlinearity and the drive terms. Starting from the lab-frame Hamiltonian above, moving to the rotating frame set by ω1, and making a standard RWA, the lab-frame Hamiltonian in the rotating frame would be:
Ĥ
RWA=Uâ†â†ââ+Δâ\â+(Λ1â\+Λ2â\â\+h.c.). (3)
The single photon loss at the rate κ may be included using a standard Lindblad master equation description. The reduced density matrix of the cavity mode may be designated by {circumflex over (ρ)}, and it may follow:
where [â]Ô=(âÔâ554 −{â†â, Ô}/2) represents the standard Lindblad dissipative superoperator.
From a quantum optics perspective, the driven cavity mode has an extremely weak Kerr nonlinearity, and are driven by simple quadratic driving terms (which on their own would only generate simple Gaussian states). In some implementations, a general approach may be taken to use linear driving (i.e., a displacement in phase space) to effectively enhance the effects of nonlinearity U. Such linear displacements may be used in other situations to enhance the properties of weakly nonlinear systems by yielding tunable linear dynamics (e.g. in parametric amplifiers realized by driving weakly nonlinear cavities, or tunable sideband interactions in quantum optomechanics). Such linear dynamics does not allow for the generation of non-classical, non-Gaussian states. Here, however, a displacement may be used to generate an effective nonlinear cavity drive with a strength»U. The interaction realized here represents a single-photon cavity drive whose magnitude is controlled by the photon number of the cavity itself, as opposed to that of a highly nonlinear auxiliary system utilized in implementations in other circumstances. The disclosure below shows that with appropriate parameter tuning, a simple displacement of the weakly nonlinear Hamiltonian in Equation (3) may yield the kind of nonlinear driving interaction that are particularly suitable for the generation of single-photon Fock state in the displaced frame. Specifically, by moving to a displaced frame of the cavity, â→âα where α is an arbitrary displacement parameter, a displacement-enhanced nonlinearity may be generated. Upon moving to a displaced frame of the cavity, the Kerr nonlinearity may generate, among corrections to the other terms in HRWA, the desired nonlinear drive {tilde over (Λ)}3â†â†â+h.c. with drive amplitude: {tilde over (Λ)}3=2Uα (see Equation (8)).
A target blockade Hamiltonian in the displaced frame, for example may be:
H
target=({tilde over (Λ)}3â†(â†â−r)+h.c.)+Uâ†â†ââ. (5)
where parameter r may be set to positive integer. This Hamiltonian describes a nonlinear driving process that may pump up an initial vacuum state to the n=r Fock state, but no higher. For r=1, the driving process can only pump up to a single-photon Fock state.
To achieve this, a displacement transformation to a new frame where the original photonic vacuum is shifted to a coherent state |−α. The corresponding unitary is
α=exp(aâ†−α*{circumflex over (α)}), which may transform the lowering operator â→â+α. In this new displacement frame, the master equation for the cavity system has the same form as Equation (4), but with a modified displaced Hamiltonian Ĥα.
Ĥ
α
=Uâ†â†ââ+
All of the terms in the original lab-frame Hamiltonian appear in Ĥ60 , but with altered coefficients. The displaced-frame Hamiltonian parameters may be:
{tilde over (Δ)}=Δ4U═α═2, (7a)
{tilde over (Λ)}1=Λ1+αΔ+2α*Λ2+2U|α|2α−1/2iκα, (7b)
{tilde over (Λ)}2=Λ2Uα2, (7c)
{tilde over (Λ)}3=2Uα. (7d)
A desired strength for the nonlinear drive amplitude {tilde over (Λ)}3 in Ĥα, may be determined first. Then the displacement parameter a may be picked to achieve this nonlinear drive amplitude. For example, the displacement parameter α may be picked via Equation (7d) as:
In some implementations, it may be designed that the drive amplitude {tilde over (Λ)}3{tilde under (>)}κ. As such, a large displacement may be desirable if the nonlinearity U is weak according to Equation (8). Further, in some implementations, the original drive parameters {tilde over (Λ)}1, {tilde over (Λ)}2, and Δ may be selected to make the remaining terms in the full displaced Hamiltonian Ĥα in Equation (6) match Ĥtarget. This leads to the choices:
With this example choice of drive parameters and displacement parameter α, the displaced-frame Hamiltonian Ĥα, would have the desired form of the target blockade-producing Hamiltonian in Equation (5). If r in Equation (9a) is picked to be an integer, it follows that blockaded dynamics in the displaced frame may be achieved. For example, the parameters can be chosen to achieve r=1 and if the system starts in the vacuum state in the displaced frame (i.e. a coherent state in the lab frame), then the full system dynamics may be confined to the Fock states n=0, n=1 in the displaced frame, regardless of how small the original value of U is.
The disclosure above thus shows how the basic principle of Equation (1) can be realized using an arbitrarily-weak Kerr nonlinearity and standard one- and two-photon driving processes. The magnitude of the nonlinear driving in the displaced frame may be the product of the original Kerr nonlinearity U (which could be extremely small) and the displacement a (which at this stage, may be assumed to be large). The process described so far only yields blockaded states and Fock states in the displaced frame. As described in the three-step protocol above , similar principle may be harnessed to generate true lab-frame Fock states by a displacement drive to transform the system from the displaced frame back to the lab frame.
The dynamics of the driven cavity system in the displaced frame is described below. Such dynamics may be described by the displaced-frame master equation
where Ĥtarget is given by Equation (5). The example dynamics is given below by taking the parameter r as close to, but not identical, to its ideal value for an n=1 Fock state blockade, i.e. r=1 +δr. In practice, δr corresponds to a failure to exactly match the one and two photon drive amplitudes in the ideal required manner, as dictated by Equations (9a) and (9b). This formulaic approach may facilitate an understanding of the temporal dynamics on time scales t{tilde under (<)}1/κ, and identifying optimal parameters for generating Fock states.
A perfect parameter tuning of δr=0 indicates ideal blockade dynamics where the drive cannot connect the n=1 and n=2 Fock states in the displaced frame. Within the blockade manifold spanned by {|0, |1
)}, the cavity behaves like a two-level-system which is resonantly driven with Rabi frequency ∝{tilde over (Λ)}3, i.e. Ĥtarget→{tilde over (Λ)}3|1
0|+h.c. As there is no probability of having two or more photons, for this perfect tuning of r, the equal-time g(2) correlation function (defined as g(2)(0)≡
â†â†ââ)/
â†â)2) is always exactly 0. To generate a single-photon state, for example, an effective it-pulse may be performed. This may be implemented by turning on the one and two photon drives (with amplitudes given by Equations (9a) and (9b)) for a time tπ=π/(2|{tilde over (Λ)}3|). This allows a generation of a Fock state in the limit where tπ«1/κ, or |{tilde over (Λ)}3|/κ»1. This condition can be met even if U«κ.
The dominant error mechanism for the scheme described above may be identified as an inability to perfectly match the drive amplitudes Λ1 and Λ2 as required to achieve r=1. For small mismatch Or, there may be only a weak matrix element connecting |1 to |2
. This may mean that there may still be approximate blockade over a long timescale, enabling the production of non-classical blockaded states. The desired perfect single photon blockade may require matching of the linear and cubic driving terms in the displaced-frame Hamiltonian Ĥα (in Equation (6)), i.e., {tilde over (Λ)}1=−{tilde over (Λ)}3 (i.e. r=1). Deviations from this amplitude-matching condition may then degrade the Fock state generation scheme above. The dimensionless relative amplitude error in the single-photon drive amplitude, δλ1, may be defined via
{tilde over (Λ)}1=−{tilde over (Λ)}3(1+δλ1). (11)
While in general both the magnitude and phase of δλ1 are important, for the small deviations, only the magnitude may significantly matter. The parameter δλ1 may be taken as real and positive for all of the numerical simulations below.
To get some analytic insight into the impact of this imperfection, the most interesting regime of small imperfection |δλ1|«1 and large effective driving, |{tilde over (Λ)}3|>κ, may be considered. For short times, dissipation can be neglected, and further, the dynamics may be restricted to the states |0, |1
, and |2
(as the leakage to higher levels is weak). In this regime, the instantaneous g(2)(0; t) may be time-independent and given by
g
(2)(0; t)=|δλ1|2. (12)
This suggests that highly blockaded states are possible without requiring an incredibly precise balancing of drive amplitudes.
â†â
˜0.5 and g(2)(0)<0.1 can be produced.
While for applications, the relatively robust blockade generation described above at short times is more than sufficient, the nature of the long-time steady state may further provide insight into the underlying dynamics. For δλ1=0, the blockade may be perfect for all times, and the steady state may have no population of higher Fock states. With imperfections, the situation may be different. As shown above, the short-time blockade may be relatively robust against amplitude mismatch errors. This however may not true for the infinite-time state. As discussed in more detail below for infinite-time state, for δλ1=0, the system may have a long-lived, metastable high-photon number state that is only able to decay via quantum tunneling.
This may manifest itself as an extremely slow relaxation rate (i.e., dissipative gap):
(see Equation (22) below). This exponentially small dissipative gap directly leads to the fragility of the steady-state photon blockade to mismatches of drive amplitude. A simple perturbative argument suggests that the steady-state blockade may be lost when |δλ1|≃γslow/κ, i.e., even when |δλ1|«1. This fragility may make the steady-state effect difficult to achieve.
This small dissipative gap, while affecting infinite-time blockade, however, does not render the finite-time blockade presented above fragile. Instead, for an imperfect system that starts from vacuum, there is a distinct metastable regime of relevance whose dynamics is controlled by a new timescale unrelated to 1/γslow. The relevant rate Γesc now corresponds to a slow escape from the blockaded subspace. For imperfect amplitude matching (δλ1≠0), there is a weak coupling between blockaded and un-blockaded subspaces. Once in the un-blockaded subspace, the system can eventually populate the weakly metastable, high-amplitude state. While this escape destroys the blockade and results in a very large average photon number in the steady state (infinite-time limit), this corruption occurs over a very slow timescale 1/Fe, and would only minimally affect the finite-time blockade dynamics.
The escape rate Γesc can be estimated using a Fermi's Golden Rule (FGR) argument where δλ1 (the imperfection in the single photon drive amplitude) is treated as a perturbation. This is consistent with the numerically observed behaviour that the average intracavity photon number approaches its steady-state value exponentially.
Defining δ{tilde over (Λ)}1={tilde over (Λ)}3×δλ1, an approximate FGR calculation yields (FGR process is provided in more detail below in a separate section):
with c is a dimensionless number. While in general it will depend on other parameters in the unperturbed Hamiltonian, for κ»{tilde over (Λ)}3, it may be found that it is constant: c=1. In contrast, for the regime of interest κ˜{tilde over (Λ)}3, a simple analytic estimate is not possible. It may, however be found from numerics in this regime (i.e. by fitting the long-time relaxation of the average photon number shown in 36 of but spread out in Fock space and thus harder to jump into. The latter effect leads generically to c<1.
The slow escape rate Γesc defines a time window over which the blockaded subspace is isolated from the rest of Hilbert space. In order to prepare Fock states, one just needs this time to be long compared to inverse drive amplitudes. In practice, this leads to the weak constraint on drive-amplitude matching δλ1<1. This is to be contrasted against the exponentially more demanding condition needed for blockade physics in the steady state, δλ1<γslow/κ. The vast difference in these conditions means that our blockade mechanism is with reach of various experimental platforms, whereas in contrast the steady-state version of the effect may be impractical.
The Fock state generation using the dynamics of Equation (4) for times much shorter than the full relaxation time of the system is described with respect to infinite-time steady state in further detail below.
As discussed above, when r is exactly tuned to an integer, the steady-state exhibits blockade: the steady-state photon number distribution truncates at n=r. This blockade phenomena may be lost even for extremely small deviations of r away from an integer. This manifests itself as an anti-resonance phenomena when the average photon number in the steady state, {circumflex over (n)}
ss is plotted versus r. There appear to be a sharp dip in this quantity when r is an integer, with the width of these features Δr«1. This behavior is illustrated in 39
{circumflex over (n)}(r)
ss centered at r=1 is plotted as a function of {tilde over (Λ)}3/U. The plot shows an exponential dependence on this parameter. Away from the blockade point r=1, the steady-state photon number is approximately constant and has a large value »1.
Both the large average steady-state photon number away from integer r, and the extremely small anti-resonance widths can be understood starting with a semiclassical analysis, which reveals a large-amplitude metastable state. The semiclassical equation of motion for the amplitude α=â
that follows from Eq. (4) is:
For r=0 (nonlinear drive only) and κ=0, the steady-state solutions to this equation are α0=0 (with multiplicity 2) and
Because a regime where U«{tilde over (Λ)}3 is assumed, this amplitude is typically very large. Including non-zero κ and r, it can be found that the first-order correction to this amplitude is small. To first order, it may be obtained that
It may also be confirmed that this is an accurate description of the large-amplitude state by numerically finding the fixed points to Equation (15) without assuming small r, κ.
It may be further shown that this semiclassical solution may be stable by performing a standard linear stability analysis of the semiclassical equations. The eigenvalues of the linearized equations of motion for α and α* about αha are
which have negative real parts, indicating linear stability at the semiclassical level. Turning to the quantum problem, this system always has a unique steady state, which for integer r, may be a blockaded state. Hence, for integer r, the above semiclassically-stable state is only unstable due to quantum effects (i.e. precisely the blockade physics we have described, which is intimately tied to the discreteness of photon number).
Returning to the quantum problem, it may be found that upon numerically diagonalizing Ĥblock in Eq. (1), there may be an eigenstate |Φ with photon number
{circumflex over (n)}
≈|αha|2 where αha is given by Eq. (18). Focusing on the single photon blockade, r=1, the Liouvillian Eq. (4) may be diagonalized, which reveals that there is generically a single nonzero eigenvalue γslow which is significantly smaller than κ; all other decay rates are of order κ or larger. It can be shown that this eigenvalue corresponds to the decay of the Hamiltonian eigenstate |Φ
and that the value is exponentially small in {tilde over (Λ)}3/U.
Working under the assumption that |Φ is the state whose decay is given by the Liouvillian eigenvalue γslow, a first order degenerate Liouvillian perturbation theory may be used to estimate γslow. The exact eigenstates within the single photon blockade manifold {|0
, |1
)} are given by
These eigenstates span the {|0, |1
} manifold so that
0|Φ
=
1|Φ
=0.
Using the numerically computed |Φ, it may be found that it is reasonably well approximated by the coherent state |αha
with overlap |
αha|Φ
2>0.96 for U«{tilde over (Λ)}3. An orthogonality with the blockade eigenstates Eq. (19) may be enforced which provides the approximate eigenstate
where is the normalization constant. Under the assumption that |ϕ
is an approximate eigenstate of Ĥblock, the relevant three-eigenstate degenerate manifold of the unperturbed Liouvillian
0=−i[Ĥblock, ⋅] is {|υ+
υ+|, |υ−
z,←υ−|ϕ
ϕ|}. (The third exact eigenstate is |Φ
Φ| of course.) The perturbation is single photon loss
1=κ[â] (21)
where [{circumflex over (X)}] is the standard Lindblad dissipator. The three state subspace with respect to
1 may be diagonalized to compute the eigenvalues. The irrelevant eigenvalues are γ0=0, whose eigenvector is the κ«{tilde over (Λ)}3 limit of the single photon blockade steady state, and γ1=κ/2, whose eigenvector describes population imbalance relative to the steady state. The final eigenvalue is the only one whose eigenvector involves |ϕ
ϕ| and for U«{tilde over (Λ)}3 may be given by
γslow≈κ|αha|2(1+2|αha|2)e−α
This shows that the dissipative gap of the blockade Liouvillian spectrum may be exponentially small in U/{tilde over (Λ)}3«1 due to a quasistable eigenstate of the coherent Hamiltonian. The description above thus provides a quantitative and intuitive understanding of the sensitivity of the steady state to small deviations of r away from integer values, explaining the sharp anti-resonance phenomena.
Further details are provided below on how to use Fermi's Golden Rule (FGR) to estimate the slow rate Γesc (Equation ((14)) which governs escape from the blockaded subspace in the presence of imperfect drive amplitudes. Consider first the simple case where κ«{tilde over (Λ)}3. The system Hamiltonian can be written as Ĥ=Ĥ0+(δ{tilde over (Λ)}â†+h.c.), where Ĥ0 is the ideal Hamiltonian with perfect drive amplitude matching (i.e. Ĥ0=Ĥtarget with r=1, c.f. Eq. (5)). Treating the last term as a perturbation, and letting |ϕj(Ej) denote eigenstates (eigenvalues) of Ĥ0, application of FGR yields
Here |ϕ± are the two blockade-subspace eigenstates of Ĥ0, and ΔE=Ej−E±. Note that only final states in the unblockaded subspace are included. When the energies and decay rates are expressed in units of κ, this expression can be put into the form Γesc=c|{tilde over (Λ)}1|2/κ. The last factor in Eq. (23) corresponds to the lifetime-broadened density of states of each unblockaded eigenstate; for weak κ, the decay rate γj=κ
ϕj|â†â|ϕj
. This general form matches that of Equation (14), with a prefactor c that in general depends on the unblockaded eigenstates of Ĥ0 and hence U/{tilde over (Λ)}3. A good agreement between Equation (23) (computed from exact diagonalization) and the rate extracted from numerical simulations of the system dynamics for weak K may be found. As an example, {tilde over (Λ)}3=100, may be considered. For U/{tilde over (Λ)}3=0.2, the estimate is c=0.0051 and the extracted value from the numerics is c=0.0047, and for U/{tilde over (Λ)}3=0.3 the estimate is c=0.0036 and the extracted value is c=0.0045. These are typical of this parameter regime. The small value of c here directly reflects the delocalization of the unblockaded eigenstates.
For more general regimes, it is trickier to directly apply FGR, as one can no longer treat the effects of κ by simply lifetime broadening each unperturbed eigenstate. For κ», on can use the fact that the large dissipation will disrupt the formation of coherent eigenstates outside the blockaded subspace. In this case, Γesc, may be estimated by considering a transition from either |ϕ± to the Fock state |2
, whose decay rate is simply 2κ. This leads to an approximate decay rate corresponding to Equation (14) with parameter-independent constant c=1. For the most relevant regime κ˜{tilde over (Λ)}3, it is difficult to rigorously calculate the decay rate as neither κ nor the unblockaded coherent dynamics can be treated perturbatively. numerically a good agreement can be found to the general form in Equation (14) with c˜0.25. Heuristically, this is consistent with the results presented above; the slightly smaller value of c corresponds to the partial delocalization of unblockaded eigenstates.
In the limit K»{tilde over (Λ)}3, the strong κ will cause states with populations of high number Fock states to rapidly decay. This implies that any eigenstates of the Hamiltonian with significant population of the |2 Fock state will be heavily localized in Fock space near |2
. Therefore an assumption may be made that Fock state |2
is a valid final eigenstate to which the cavity jumps. Using as the initial state either of the blockade eigenstates |ϕ±
=(|0
±|1
)/√{square root over (2)}, and estimating the final density of states as the amplitude decay rate ρ(Ef)=1/κ of |2
(the energy difference is «κ), the escape rate may be arrived at:
Thus c=1 in this limit. it is expected that this to be an upper bound on c because the energy eigenstates of the blockade Hamiltonian Equation (5) which have significant |2 populations are highly delocalized in Fock space. This implies that the matrix elements will be smaller and that the decay rates of the final states will be larger than what we have estimated here. Together these imply a smaller Γesc.
In the limit κ«{tilde over (Λ)}3, the Hamiltonian eigenstates may not be perturbed much by the loss, thus the exact energy eigenstates in the Fermi's Golden Rule matrix elements may be used. While there are no analytic expressions for these but they may be obtained numerically by exact diagonalization. Generically a few eigenstates with large |2 population may be identified and they may be all very delocalized in Fock space, but relatively close in energy to the unblockaded eigenstates |ϕ35
. Treating κ as a perturbation, the first order decay rate of each state |ϕj
may be computed as:
Using these decay rates and the unperturbed energies of the eigenstates, the escape rate may be further estimated as already shown above in Equation (23)
The short-time blockade dynamics considered above requires {tilde over (Λ)}3>κ. Via Equation (8), it can be seen that this is possible even if U«κ, as long as a large displacement αb is used. While at a fundamental level such large displacements pose no problems, at a practical level they can create issues. This can be seen explicitly in the process of transforming the displaced-frame Fock states produced by Equation (5) to true lab-frame Fock states.
Given this possible concern, it may also be relevant to describe the dynamics of system where |{tilde over (Λ)}3|«κ, a regime that could be reached with small U and modest displacements α. Consider first the case where the drive amplitudes are perfectly matched, implying r=1 in Equation (5). In this case, the system approaches the infinite-time, perfectly-blockaded steady state on a timescale ˜1/κ. This state has zero probability for having more than one photon, and the single photon occupancy may be expressed as
Hence, having a weak {tilde over (Λ)}3/κ does not break the blockade, but just reduces the population of the one photon state. In this weak drive regime, the blockade is much more robust to amplitude mismatch errors.
Based on the example three-step protocol and other photon blockade dynamics described above, a numerically simulation of the performance the example three-step protocol may be generated. In particular, the Step 2 may be modeled by evolving the cavity system as per the full master equation in Equation (4)). The possibly imperfect displacement operations in steps 1 and 3 may be further modeled as a combination of a perfect displacement and the injection of thermal noise (corresponding to and adding
In addition to displacement errors, drive amplitude mismatches discussed above may be further considered. A full Fock state generation protocol may be performed with small δλ1≠0. The figure of merit for the Fock state generation protocol may be the instantaneous second order coherence g(2)(0) at the end of the protocol as a function of U/κ.
Shown in 1═/{circumflex over (ρ)}|1
=0.5. The numerical results show that the blockade protocol is effective even for U/κ˜0.03, and moreover, is robust against both small displacement errors and small amplitude match errors. There is no fundamental limit against applying our protocol for even smaller values of U. The dashed line sin
The basic scheme in Eq. (1) may be further extended to a mechanism that allows the generation of entangled M-mode non-Gaussian photonic states using only weak nonlinearities U«κ. The disclosure below describe the extension to a M=2 mode system with lowering operators â1, â2. The additional implementations below realize the nonlinear driving Hamiltonian in Equation (1) in a displaced frame, but now the single-mode lowering operator a applies to a collective mode, e.g.
{circumflex over (b)}=(â1+â2)/√{square root over (2)} (27)
The desired nonlinear driving Hamiltonian may be
Ĥ
block2={tilde over (Λ)}{circumflex over (b)}†({circumflex over (b)}†{circumflex over (b)}−r)+h.c. (28)
The description below focus on realizing a single-excitation state, and hence take r=1. Following the various principles above, the ideal dynamics under this Hamiltonian can prepare a single excitation in the collective b mode. If the photon number of each a1, a2 mode is used to separately encode, for example, m a quantum bit (qubit), then the single excitation state produced may be a maximally entangled Bell state (|10+|01
)/√{square root over (2)}.
The Hamiltonian in Equation ((28) amounts to simply replacing â in Equation (1) of the main text with the collective mode {circumflex over (b)}. It thus immediately follows that if the four-wave mixing Hamiltonian Equation (3) is taken and â is replaced with {circumflex over (b)}, then the resulting Hamiltonian is equivalent to Equation (28) up to unitary displacement operators of modes â1 and â2.
Given the above, a simple substitution in Equation (3) yields the required form of the starting two-mode Hamiltonian
Ĥ
RW A,2
=U{circumflex over (b)}†{circumflex over (b)}†{circumflex over (b)}{circumflex over (b)}+Δ{circumflex over (b)}†{circumflex over (b)}+(Λ1{circumflex over (b)}†+Λ2{circumflex over (b)}†{circumflex over (b)}†+h.c.). (29)
where again {circumflex over (b)} is given by Equation (27). The linear and quadratic terms that are generated describe linear drives on both modes, detuning and beam-splitter couplings, and parametric drives (both degenerate and non-degenerate). The nonlinear four-wave mixing terms take the form
Note that now both self and cross Kerr interactions and four-wave mixing processes that exchange interactions between modes 1 and 2 may be required.
Following the principles described in various sections above, each mode âj may be displaced by the amplitude αb given in Equation (8). By further tuning the parameters the drive parameters Λ1, Λ2 and detuning parameter Δ as per Eqs. (9), the above two-mode Hamiltonian is unitarily equivalent to the desired two-mode blockade Hamiltonian in Eq. (28). The basic ideas may thus be generalized directly to multimode systems. Other related approaches are also possible. While the introduction of modes does involve more complexities, the basic feature of the original scheme remains: generation of non-classical blockaded (now entangled) states is possible even if the four-wave mixing nonlinearities U are much weaker that photonic loss.
The disclosed Fock state generation scheme is generic: there are many different kinds of systems that can realize weakly-nonlinear electromagnetic modes with one- and two-photon drives. In the context of weakly nonlinear optical cavities, the primary implementation challenge is the large cavity displacements required, αb˜κ/U. For typical low-loss silicon micro-resonators, the intrinsic x(3) nonlinearity yields U/κ˜10−8. The x(3) of silicon nitride may be smaller. While the large displacements and intracavity powers required in such systems to achieve αb˜κ/U may be possible given the pulsed nature of the example disclosed scheme, a route of the general ideas in the “Photon blockade with weak drive” section may be followed. Here, displacements much smaller than κ/U may be used, making constraints on power handling much more reasonable. This may result in a perfect blockade and states with vanishingly small g(2)(0). The resulting average photon number may be small but the Fock state so generated have a strong advantage over other existing approaches that only generate Gaussian state and have non-zero population of higher Fock states.
An alternative route for implementation in optical cavities would be to utilize x(2) nonlinearities in materials with broken inversion symmetry like silicon nitride or aluminum nitride. These nonlinearities may be parametrically larger than the corresponding x(3). While the scheme above is suitable for a Kerr-type four-wave mixing nonlinearity, this can be achieved starting with three-wave mixing x(2) processes that generates a nonlinear coupling to a detuned auxiliary mode. To second order in this coupling, a desired self-Kerr interaction U needed for our scheme may be generated. Despite being second order, this can still be orders-of-magnitude larger than an intrinsic x(3) nonlinearity.
While optical cavities are one possible domain of application, they are not the only candidate. The implementations above can also be exploited in parametrically-driven nanomechanical systems with weak intrinsic Duffing nonlinearities, as well as in microwave cavity systems. A current trend in quantum information processing with superconducting circuits is to store and process information in high-Q microwave cavities. In such schemes, detuned qubits are often used to induce weak nonlinearities in the principle bosonic modes. A key limitation in these approaches is that the qubit also induces new loss mechanisms. The scheme above provides a path to circumvent this limitations in existing approaches. One could use extremely large qubit-cavity detunings, resulting in both very weak induced cavity nonlinearities, but also weak induced dissipation. The disclosure above shows that such weak nonlinearities could still be harnessed to produce non-classical states using the disclosed implementations.
As such, a new basic route to generating photonic states that are blockaded is described: they have a sharp cutoff in their photon-number distribution, having zero probability to have more than r photons in the state. This is accomplished by using standard tools (a weak Kerr nonlinearity, one and two photon drives) to realize an effective non-linear drive (Equation (5)). In stark contrast to the well-studied unconventional photon blockade mechanism, the scheme and implementations described above can generate truly blockaded states, and states that do not need to be infinitely close to being vacuum. In principle, the basic mechanism is effective even for arbitrarily weak nonlinearities U«κ. In practice, limitations will arise from the inability to perfectly match the one and two photon drive amplitudes, and the inability to apply the required displacement transformations perfectly. The scheme disclosed above can be effective even if these imperfections are present.
While the above description focuses on generating states that approximate single-photon Fock states, the underlying principles are much more general. By picking the parameter r in Equation (5) to be an integer larger than one (which then influences the choice of drive amplitudes via Equations ((9a)) and ((9b))), one can generate higher-order blockaded states: states that are confined to the manifold spanned by Fock states |0, |1
, . . . , r
. Further, the same basic idea can used to generate non-classical, multi-mode entangled states. One again realizes the nonlinear driving Hamiltonian in Equation (5) in a displaced frame, but now the mode a is actually a collective mode of 2 or more distinct cavity modes. Generating â Fock state in this collective mode directly corresponds to a W-style entangled state.
The mechanism disclosed herein may provide a tool for generating non classical photonic states in a variety of platforms where only weak nonlinearities are achievable. It could also conceivably be harnessed as a tool for quantum simulation, i.e., to realize models of strongly interacting photons. The underlying principles described above are compatible with a wide variety of bosonic systems, including optical and microwave cavities, as well as more general superconducting circuit QED setups. For example, the resonant cavity in the optical spectral range may be implemented as Fabry-Pérot resonator, optical ring resonator, and the like.
The description and accompanying drawings above provide specific example embodiments and implementations. Drawings containing device structure and composition, for example, are not necessarily drawn to scale unless specifically indicated. Subject matter may, however, be embodied in a variety of different forms and, therefore, covered or claimed subject matter is intended to be construed as not being limited to any example embodiments set forth herein. A reasonably broad scope for claimed or covered subject matter is intended. Among other things, for example, subject matter may be embodied as methods, devices, components, or systems. Accordingly, embodiments may, for example, take the form of hardware, software, firmware or any combination thereof.
Throughout the specification and claims, terms may have nuanced meanings suggested or implied in context beyond an explicitly stated meaning. Likewise, the phrase “in one embodiment/implementation” as used herein does not necessarily refer to the same embodiment and the phrase “in another embodiment/implementation” as used herein does not necessarily refer to a different embodiment. It is intended, for example, that claimed subject matter includes combinations of example embodiments in whole or in part.
Unless defined otherwise, all technical and scientific terms used herein have the same meaning as commonly understood by one of skill in the art to which the invention pertains. Although any methods and materials similar to or equivalent to those described herein can be used in the practice or testing of the present invention, the preferred methods and materials are described herein.
In general, terminology may be understood at least in part from usage in context. For example, terms, such as “and”, “or”, or “and/or,” as used herein may include a variety of meanings that may depend at least in part on the context in which such terms are used. Typically, “or” if used to associate a list, such as A, B or C, is intended to mean A, B, and C, here used in the inclusive sense, as well as A, B or C, here used in the exclusive sense. In addition, the term “one or more” as used herein, depending at least in part upon context, may be used to describe any feature, structure, or characteristic in a singular sense or may be used to describe combinations of features, structures or characteristics in a plural sense. Similarly, terms, such as “a,” “an,” or “the,” may be understood to convey a singular usage or to convey a plural usage, depending at least in part upon context. In addition, the term “based on” may be understood as not necessarily intended to convey an exclusive set of factors and may, instead, allow for existence of additional factors not necessarily expressly described, again, depending at least in part on context.
Further, the terms “cavity,” “resonant cavity,” “resonance cavity,” and “resonator” are used interchangeably in the disclosure above. A cavity or resonator may be characterized by multiple resonant modes. The implementations above may be applied to generation of Fock states in one of the multiple resonance modes independent of other resonant modes. The terms “resonance mode,” “resonant electromagnetic mode,” “resonant mode,” and the like are used interchangeably.
Reference throughout this specification to features, advantages, or similar language does not imply that all of the features and advantages that may be realized with the present solution should be or are included in any single implementation thereof. Rather, language referring to the features and advantages is understood to mean that a specific feature, advantage, or characteristic described in connection with an embodiment is included in at least one embodiment of the present solution. Thus, discussions of the features and advantages, and similar language, throughout the specification may, but do not necessarily, refer to the same embodiment.
Furthermore, the described features, advantages and characteristics of the present solution may be combined in any suitable manner in one or more embodiments. One of ordinary skill in the relevant art will recognize, in light of the description herein, that the present solution can be practiced without one or more of the specific features or advantages of a particular embodiment. In other instances, additional features and advantages may be recognized in certain embodiments that may not be present in all embodiments of the present solution.
This invention was made with government support under grant number 1420709 awarded by the National Science Foundation, FA9550-19-1-0399 awarded by the U.S. Air Force Office of Scientific Research, and W911NF-19-1-0380 awarded by the U.S. Army Research Office. The government has certain rights in the invention.
| Number | Date | Country | |
|---|---|---|---|
| 63312461 | Feb 2022 | US |
