Abstract
Gyroscope for rotation sensing plays a key role in inertial navigation systems. Developing more precise gyroscopes than the conventional ones bounded by the classical shot-noise limit by using quantum resources has attracted much attention. However, existing quantum gyroscope schemes suffer severe deterioration under the influence of decoherence, which is called the no-go theorem of noisy metrology. Here, by using two quantized optical fields as the quantum probe, we propose a quantum gyroscope scheme breaking through the constraint of the no-go theorem. Our exact analysis of the non-Markovian noise reveals that both the evolution time as a resource in enhancing the sensitivity and the achieved super-Heisenberg limit in the noiseless case are asymptotically recoverable when each optical field forms a bound state with its environment. The result provides a guideline for realizing high-precision rotation sensing in realistic noisy environments.
© 2023 Chinese Laser Press
1. INTRODUCTION
High-performance gyroscopes for rotation sensing are of pivotal significance for navigation in many types of air, ground, marine, and space applications. Based on the Sagnac effect, i.e., two counter-propagating waves in a rotating loop accumulate a rotation-dependent phase difference, gyroscopes have been realized in optical [1–6] and matter–wave [7–14] systems. However, the precision of a purely Sagnac gyroscope, which is proportional to the surface area enclosed by the optical path, is theoretically limited by the classical shot-noise limit (SNL). It dramatically constrains their practical application and further performance improvement. The records for precision and stability of commercial gyroscopes are held by optical gyroscopes [15,16]. To reduce the noise effect, the practical operation of fiber optical gyroscopes generally modulates the optical signal and measures the ratios of harmonics instead of the phase difference [17], where the classical SNL model is not widely used. Building a purely Sagnac optical gyroscope beating the SNL from the fundamental principle is highly desired.
Pursuing more precise measurement to physical quantities than the classical SNL by using quantum resources [18–22], such as squeezing [23–25] and entanglement [26,27], quantum metrology supplies a way toward achieving gyroscopes with ultimate sensitivity limits. Based on this idea, many schemes of quantum gyroscopes have been proposed. It was found that the entanglement in N00N states [28,29], continuous-variable squeezing [30–32], and optical nonlinearity [33] can enhance the sensitivity of optical gyroscopes beyond the SNL. A quantum-enhanced sensitivity can also be achieved in matter–wave gyroscopes [34–37] by using spin squeezing [38–40] or entanglement. However, quantum gyroscopes are still at the stage of the proof-of-principle study, and their superiority over the conventional ones in the absolute value of sensitivity still has not been exhibited [22,36]. One key obstacle is that the stability of the quantum gyroscope is challenged by the decoherence caused by inevitable noise in microscopic world, which generally makes the quantum resources degraded. It was found that the metrology sensitivity using entanglement [41–43] and squeezing [44,45] exclusively returns to or even becomes worse than the SNL; thus, their quantum superiority completely disappears when the photon loss is considered. This is called the no-go theorem of noisy quantum metrology [46,47] and is one difficulty in achieving a high-precision quantum gyroscope in practice.
In this paper, we propose a scheme of quantum gyroscope and discover a mechanism to overcome the constraint of the no-go theorem on our scheme. A super-Heisenberg limit (HL) on the sensitivity is achieved in the ideal case by using two-mode squeezed vacuum state. Our exact analysis on the photon dissipation reveals that the performance of the quantum gyroscope in the dissipative environments intrinsically depends on the energy-spectrum feature of the total system formed by the probe and its environments. The encoding time as resource and the super-HL in the ideal case are asymptotically recovered when each optical field forms a bound state with its environment, which means that the no-go theorem is efficiently avoided. It supplies a guideline to engineer the optimal working condition of our quantum gyroscope in dissipative environments.
2. IDEAL QUANTUM GYROSCOPE SCHEME
To measure a physical quantity of certain systems, three processes, i.e., the initialization of the quantum probe, the quantity encoding via the probe-system coupling, and the measurement, are generally required. In our quantum gyroscope, we choose two beams of quantized optical fields as the quantum probe. They propagate in opposite directions and are input into a 50:50 beam splitter and split into clockwise and counterclockwise prorogating beams [see Fig. 1(a)]. The setup rotates with an angular velocity about the axis perpendicular to its plane. Thus, the two beams accumulate a phase difference when they reencounter the beam splitter after rounds of propagation in the circular path [48]. Here is the wave vector, is the speed of light, and is the radius of the quantum gyroscope. Remembering the standing-wave condition () of the optical fields propagating along the circular path and defining , we have . Therefore, the quantum gyroscope can be equivalently treated as two counter-propagating optical fields with a frequency difference along the circular path. For concreteness, we choose the basic mode . Then the optical fields in the quantum gyroscope can be quantum mechanically described by [49]
where is the annihilation operator of the th field with frequency . The optical fields couple to the beam splitter twice and output in the state , where is the evolution operator of the fields and describes the action of the beam splitter. Thus, the angular velocity is encoded into the state of the optical probe via the unitary evolution.To exhibit the quantum superiority, we employ two-mode squeezed vacuum state as the input state , where is the squeezing operator, with being the squeezing parameter. The total photon number of this input state is , which is the quantum resource of our scheme. The parity operator is measured at the output port [50]. To the output state , we can calculate and , where has been used. Then the sensitivity of sensing can be evaluated via the error propagation formula as
when with . It is remarkable to find that the best sensing error achieved in our scheme is even smaller than the HL , which reflects the quantum superiority of the used squeezing and measured observable in our scheme. It can be verified that this measurement scheme saturates the Cramér–Rao bound governed by quantum Fisher information. We call such a sensitivity surpassing the HL the super-HL [51,52]. It is noted that a phase estimation error smaller than the inverse of the mean photon number was called the sub-HL in Refs. [50,53,54]. The outstanding performance of quantum squeezing has been found in gravitational wave detection [55].3. EFFECTS OF DISSIPATIVE ENVIRONMENTS
The superiority of the quantum sensor is challenged by the decoherence of the quantum probe due to the inevitable interactions with its environment. Depending on whether the probe has energy exchange with the environment or not, the decoherence can be classified into dissipation and dephasing. The main decoherence in our quantum gyroscope is the photon dissipation. The previous works phenomenologically treat the photon dissipation by introducing an imperfect transmission to the beam splitter [41,44,45,56,57], which is equivalent to the Born–Markovian approximate description. Such an approximation is convenient, but it might miss important physics. It has been found that the system–environment interplay caused by the inherent non-Markovian nature would induce diverse characters absent in the Born–Markovian approximation [58–62]. To reveal the practical performance of our quantum gyroscope, we, going beyond the Born–Markovian approximation and paying special attention to the non-Markovian effect, investigate the impact of the photon dissipation on the scheme.
We consider that the encoding process is influenced by the photon dissipation, which is caused by the energy exchange between the two optical fields and two independent environments [48]. The Hamiltonian of the total system is
where is the annihilation operator of the th mode with frequency of the environment felt by the optical field and is their coupling strength. The coupling is further characterized by the spectral density in the continuous limit of the environmental frequencies. We consider the Ohmic-family spectral density for both environments, where is a dimensionless coupling constant, is a cutoff frequency, and is an Ohmicity index. Under the condition that the environments are initially in the vacuum state, we can derive an exact master equation for the encoding process using the Feynman–Vernon influence functional method [63], where is the Lindblad superoperator, is the renormalized frequency, and is the dissipation rate. The time-dependent functions satisfy under , where and is the environmental correlation function. Equation (4) indicates that all the non-Markovian effects induced by the environmental backactions have been incorporated into these time-dependent coefficients self-consistently. Solving Eq. (4), we obtain (see Appendix A) where , , and , with . The analytical form of can then be calculated in a similar manner as the ideal case.In the special case when the probe–environment coupling is weak and the time scale of is smaller than the typical time scale of the probe, we can apply the Born–Markovian approximation in Eq. (5) [41,57]. Their approximate solutions read , with and . Substituting them into Eq. (6) and using the error propagation formula, we obtain (see Appendix B)
where . Here we have chosen . We plot in Fig. 1(b) the evolution of . It can be found that experiences an obvious oscillation with time. However, the best sensitivity manifested by the profile of its local minima tends to be divergent with time. Thus, being in sharp contrast to the ideal case in Eq. (2), the superiority of time as a resource in enhancing the precision of the quantum gyroscope disappears. After optimizing the encoding time, we obtain the global minimum [see the red dot in Fig. 1(b)]. The numerical fitting reveals [see Fig. 1(c)], which is even worse than the SNL. Therefore, being consistent with the previous quantum sensing schemes [41,44,45,56,57], the photon dissipation under the Born–Markovian approximation makes the quantum advantages of our scheme completely vanish. It is called the no-go theorem of noisy quantum metrology [46,47] and is the main obstacle to achieve a high-precision quantum sensing in practice.In the general non-Markovian case, the analytical solution of Eq. (5) can be found by the method of Laplace transform, which converts Eq. (5) into . Then is obtained by making the inverse Laplace transform to , which can be done by finding its poles from
Here, are also the eigenenergies in the single-excitation subspace of the total systems formed by each optical field and its environment. To see this, we expand the eigenstate as . From the stationary Schrödinger equation, we have and with being the eigenenergies. The two equations readily result in Eq. (8) in the continuous limit of the environmental frequencies. It implies that the dissipation of the optical probe is intrinsically determined by the energy-spectrum character of the probe–environment system in the single-excitation subspace, even though the subspaces with any excitation numbers are involved. Due to being decreasing functions in the regime , each of Eq. (8) has one isolated root in this regime provided . While are ill-defined when due to the poles in the integrand, they have infinite roots in this regime, which form a continuous energy band. We call the eigenstates of the isolated eigenenergies bound states [58]. Making the inverse Laplace transform, we obtain , where and . The integral in is from the energy band and tends to zero in the long-time limit due to the out-of-phase interference. Thus, when the bound state is formed, we have , characterizing the suppressed dissipation; otherwise, we have , meaning a complete dissipation. It can be determined that the bound state is formed for the Ohmic-family spectral density when , where is the Euler’s function.We have three parameter regimes where zero, one, and two bound states are formed, respectively. It is natural to expect that in the former two regimes is qualitatively consistent with the Born–Markovian approximate result Eq. (7) due to the complete dissipation in either two or one optical field. Focusing on the case in the presence of two bound states and substituting the asymptotic solution into Eq. (6), we obtain (see Appendix C)
where with . We have used derived from Eq. (8). Equation (9) exhibits a -dependence on time, which is as perfect as the ideal result Eq. (2). Another finding from Eq. (9) is that it tends to the exactly same form as the ideal result Eq. (2) in the limit tending to 1. Therefore, the formation of two bound states overcomes the problem of no-go theorem and asymptotically retrieves the ideal sensitivity.4. NUMERICAL RESULTS
We now numerically verify our general result by choosing the Ohmic spectral density. Figure 2(a) shows the energy spectrum of the total system consisting of the optical fields and their environments. It can be seen that the two branches of bound states divide the energy spectrum into three regimes: without bound state when , one bound state when , and two bound states when . The result confirms our analytical criterion that the bound states are formed when . Numerically solving Eq. (5) and using Eq. (6), we obtain the exact evolution of in the three regimes. When no or one bound state is formed, the local minima profile of tends to divergence in the long-time limit, and the quantum superiority of the scheme completely disappears [see Figs. 2(b) and 2(c)], which is qualitatively similar to the Born–Markovian result. However, as long as two bound states are formed, the profile of the local minima becomes a decreasing function of the encoding time. The matching of the numerical result with the long-time behavior Eq. (9) verifies the validity of the result in Eq. (9). Thus, the encoding time as a resource in sensing is recovered as perfectly as the ideal case by the formation of the two bound states. Our above result also gives a direct proof on that whether the Born–Markovian approximation is applicable or not depends sensitively on the feature of the energy spectrum of the total probe–environment system. Whenever a bound state is formed in the energy spectrum, the decoherence would be suppressed, and the Born–Markovian approximation would no long be valid anymore. This result refreshes our general belief on applicability of the Born–Markovian approximation.
Figure 3(a) shows the evolution of the local minima of in Eq. (9) at different when two bound states are formed. The formation of the bound state causes an abrupt increase of the corresponding from zero to a finite value exactly matching with [see the inset of Fig. 3(b)]. It is interesting to find that not only the encoding time as a resource is retrieved, but also the ideal precision is asymptotically recovered. This is double confirmed by the long-time behavior of as a function of the photon number in Fig. 3(b). A similar performance is found by changing the coupling constant (see Fig. 4). All the results demonstrate the constructive role played by the two bound states and the non-Markovian effect in retrieving the quantum superiority of our quantum gyroscope. It offers us a guideline to achieve a noise-tolerant rotation sensing by manipulating the formation of the bound states. It is noted that, according to the condition of forming the bound states, we see that what really matters is the relative value . The equivalent result is achievable by tuning for given and .
5. DISCUSSION AND CONCLUSION
Our scheme is independent of the form of the spectral density. Although only the Ohmic form is considered, our scheme can be generalized to other spectra. Given the rich way of controlling the spectral density in the setting of quantum reservoir engineering [64,65], we deem that our scheme is realizable in state-of-the-art quantum-optical experiments. The Ohmic-family spectral densities for the electromagnetic noise are well cotrolled in circuit QED systems [66–68]. Actually, the non-Markovian effect has been observed in the linear optical systems [69,70]. The bound state and its dynamical effect have been observed in circuit QED [71] and ultracold atom [72] systems. A squeezing parameter , which corresponds to , has been realized [73]. Inspired by these experimental achievements in circuit QED systems, we can design a realizable microwave Sagnac interferometer to test our result. We prepare two quantized optical fields propagating in opposite directions in a 1D transmission line, which is coupled via a capacitance or inductance to another wide-band transmission line acting as a structured environment [71]. The squeezed state of the two fields can be generated by the Josephson traveling-wave parametric amplifier [73]. When the fields reencounter after several rounds of rotation, a phase difference depending on the measured angular velocity is accumulated. The structured environments, on one hand, exert a strong non-Markovian effect on the dynamics of the fields and, on the other hand, protect the scheme from the photon loss according to our mechanism.
In summary, we have proposed a quantum gyroscope scheme by using two quantized fields as the quantum probe, which achieves a super-HL sensitivity in measuring the angular velocity. However, the photon dissipation under the conventional Born–Markovian approximation forces this sensitivity to even be worse than the classical SNL. To overcome this problem, we have presented a mechanism to retrieve the ideal sensitivity by relaxing this approximation. It is found that the ideal sensitivity is asymptotically recoverable when each optical field forms a bound state with its environment, which can be realized by the technique of quantum reservoir engineering. Exhibiting the optimal working condition of the quantum gyroscope, our mechanism breaks through the constraint of the no-go theorem of noisy quantum metrology and supplies a guideline for developing high-precision rotation sensing for next-generation inertial navigation systems.
APPENDIX A: EXPECTATION VALUE OF PARITY OPERATOR
In this section, we give the derivation of Eq. (6). The Feynman and Vernon’s influence-functional theory enables us to derive the evolution of the reduced density matrix of the quantum probe formed by two quantized optical fields exactly. By expressing the forward and backward evolution operators of the density matrix of the probe and the environments as a double path integral in the coherent-state representation and performing the integration over the environmental degrees of freedom, we incorporate all the environmental effects on the probe in a functional integral named influence functional. The reduced density matrix fully describing the encoding dynamics of the probe is given by [63]
where is the reduced density matrix expressed in coherent-state representation and is the propagating function. In the derivation of Eq. (A1), we have used the coherent-state representation with , which are the eigenstates of annihilation operators, i.e., , and obey the resolution of identity with . denotes the complex conjugate of . The propagating function is expressed as the path integral governed by an effective action, which consists of the free actions of the forward and backward propagators of the optical probe and the influence functional obtained from the integration of environmental degrees of freedom. After evaluation of the path integral, its final form reads where satisfies with and . We have assumed that the spectral density of the two environments are identical.The input state of the probe is a two-mode squeezed vacuum state , where is the squeezing parameter. After passing the first beam splitter of the quantum optical gyroscope, the state changes into , with , which acts as the initial state of the encoding dynamics. In the coherent-state representation, this initial state is given by
The time-dependent reduced density matrix is obtained by integrating the propagating function over the initial state of Eq. (A1). It reads where , , and , with . Remembering and , we obtain Then the expectation value of the parity operator can be calculated as where has been used. The sensing sensitivity of is calculated by .In the ideal limit, the solution of Eq. (A3) reads ; thus, , , and . Then Eq. (A7) reduces to with .
APPENDIX B: SENSITIVITY UNDER THE BORN–MARKOVIAN APPROXIMATION
Defining , we can rewrite Eq. (A3) as
When the probe–environment coupling is weak and the time scale of the environmental correlation function is much smaller than the one of the probe, we can apply the Born–Markovian approximation to Eq. (B1) by neglecting the memory effect, i.e., , and extending the upper limit of the integral to infinity, i.e., . The utilization of the identity , with being the Cauchy principal value, results in , where and . We, thus, have the Born–Markovian approximate solution of as .Substituting into Eq. (A7) and using the error propagation formula, we analytically obtain the sensitivity under the Born–Markovian approximation as
where . Here we have chosen and neglected the constant , which is generally renormalized into . We readily see from Eq. (B2) that the sensitivity under the Born–Markovian approximation tends to be divergent in the long-time limit.APPENDIX C: SENSITIVITY IN THE NON-MARKOVIAN DYNAMICS
In the non-Markovian case, Eq. (A3) can be analytically solvable by the method of Laplace transform, which converts Eq. (A3) into . Then is obtained by applying the inverse Laplace transform on , we obtain
where and is chosen to be larger than all the poles of the integrand. We find the pole of Eq. (C1) from It is noted that also represents the eigenenergy in the single-excitation subspace of the total systems formed by each optical field and its environment. To see this, we expand the eigenstate as . From the stationary Schrödinger equation, we have and , with being the eigenenergy. These two equations readily result in Eq. (C2) in the continuous limit of the environmental frequencies. According to the residue theorem, we have where and . The first and the second terms of Eq. (C3) are the residues contributed from the poles of Eq. (C2) in the regime and , respectively. It can be found that are decreasing functions in the regime , and each of Eq. (C2) has one isolated root in this regime provided . While are not well analytic in the regime , they have infinite roots in this regime, which form a continuous energy band. We call the eigenstates of the isolated eigenenergies bound states. We plot in Fig. 5 the solution of Eq. (C2) obtained by the graphical method. It verifies the three typical features on the solutions, i.e., no bound state when in Fig. 5(a), one bound state when but in Fig. 5(b), and two bound states when in Fig. 5(c).Due to the out-of-phase interference, the second term in Eq. (C3) tends to vanish in the long-time limit. Therefore, we have the asymptotical solution of Eq. (C3) as
Due to the dominated role played by in the encoding dynamics of the quantum probe, the two qualitatively different asymptotical behaviors in Eq. (C5) manifest the significance of the bound states in determining the photon dissipation. The former case of Eq. (C3) characterizes a complete dissipation, while the latter case denotes a suppressed dissipation. Figures 5(d)–5(f) show the evolution of . This verifies that, as long as the bound state is formed, the corresponding tends to a finite value, which exactly matches .Focusing on the case in the presence of two bound states and substituting the asymptotic solution into Eq. (A7), we obtain
where with . We have used and derived from Eq. (C2). It can be observed that Eq. (C6) is dominated by the second term in the large-time limit due to the time-independence of its first and the third terms. We, thus, have Equation (C7) behaves as . Therefore, the problem that the sensing error asymptotically tends to be divergent in the Born–Markovian approximation is overcome.On the one hand, Eq. (C7) in the large- limit behaves as
On the other hand, Eq. (C7) in the limit of tending to 1 behaves as when . It is just the ideal sensitivity. Thus, the limits of the steady-state under different orders of and have different scaling relations with . Equations (C8) and (C9) imply that, in the general case , the scaling relation of with must have a threshold. Taking the same set of parameter values as the blue dashed line in Fig. 4(b), we plot in Fig. 6(a) the global picture of the steady-state as a function of . There indeed exists a threshold at . Below the threshold , takes the form as Eq. (C9). Above the threshold , it takes the form as Eq. (C8). Figure 6(b) shows the threshold as a function of . It reveals that the closer is to 1, the larger is. In the limit tending to 1, tends to an infinity, and the scaling behavior Eq. (C8) is completely avoided.Experimentally, a squeezing parameter , which corresponds to , has been realized [73]. We see from Figs. 3(b) and 4(b) that the threshold is still absent even when is as large as 100. Therefore, although we have to face the balance between the bound-state favored restoring superiority and the photon-dissipation caused destruction to the sensitivity, our mechanism still supplies us with a sufficient space to rescue the ideal sensitivity from the noise using the experimentally accessible numbers of quantum resources.
Funding
National Natural Science Foundation of China (11834005, 11875150, 12047501).
Disclosures
The authors declare no conflicts of interest.
Data Availability
The data that support the plots within this paper and other findings of this paper are available from the corresponding author upon reasonable request.
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