### 56121 results for qubit oscillator frequency

Contributors: Whittaker, J. D., da Silva, F. C. S., Allman, M. S., Lecocq, F., Cicak, K., Sirois, A. J., Teufel, J. D., Aumentado, J., Simmonds, R. W.

Date: 2014-08-08

We describe a tunable-cavity QED architecture with an rf SQUID phase **qubit** inductively coupled to a single-mode, resonant cavity with a tunable **frequency** that allows for both microwave readout of tunneling and dispersive measurements of the **qubit**. Dispersive measurement is well characterized by a three-level model, strongly dependent on **qubit** anharmonicity, **qubit**-cavity coupling and detuning. A tunable cavity **frequency** provides a way to strongly vary both the **qubit**-cavity detuning and coupling strength, which can reduce Purcell losses, cavity-induced dephasing of the **qubit**, and residual bus coupling for a system with multiple **qubits**. With our **qubit**-cavity system, we show that dynamic control over the cavity **frequency** enables one to avoid Purcell losses during coherent **qubit** evolutions and optimize state readout during **qubit** measurements. The maximum **qubit** decay time $T_1$ = 1.5 $\mu$s is found to be limited by surface dielectric losses from a design geometry similar to planar transmon **qubits**....The two possible flux values at the readout spot leads to two possible **frequencies** for the tunable cavity coupled to the **qubit** loop. Similar microwave readout schemes have been used with other rf-SQUID phase **qubits** . For our circuit design, the size of this **frequency** difference is proportional to the slope d f c / d φ c of the cavity **frequency** versus flux curve at a particular cavity flux φ c = Φ c / Φ o . The transmission of the cavity can be measured with a network analyzer to resolve the **qubit** flux (or circulating current) states. The periodicity of the rf SQUID phase **qubit** can be observed by monitoring the cavity’s resonance **frequency** while sweeping the **qubit** flux. This allows us to observe the single-valued and double-valued regions of the hysteretic rf SQUID. In Fig. Fig4(a), we show the cavity response to such a flux sweep for design A . Two data sets have been overlaid, for two different **qubit** resets ( φ q = ± 2 ) and sweep directions (to the left or to the right), allowing the double-valued or hysteretic regions to overlap. There is an overall drift in the cavity **frequency** due to flux crosstalk between the **qubit** bias line and the cavity’s rf SQUID loop that was not compensated for here. This helps to show how the **frequency** difference in the overlap regions increases as the slope d f c / d φ c increases....(Color online) (a) **Qubit** spectroscopy (design A ) overlaid with cavity spectroscopy at two **frequencies**, f c = 6.58 GHz and 6.78 GHz. (b) Zoom-in of the split cavity spectrum in (a) when f c = 6.78 GHz with corresponding fit lines. (c) Zoom-in of the split cavity spectrum in (a) when f c = 6.58 GHz with corresponding fit lines. (d) Cavity spectroscopy (design B ) while sweeping the **qubit** flux with f c = 7.07 GHz showing a large normal-mode splitting when the **qubit** is resonant with the cavity. All solid lines represent the uncoupled **qubit** and cavity **frequencies** and the dashed lines show the new coupled normal-mode **frequencies**. Notice in (d) the additional weak splitting from a slot-mode just below the cavity, and in (c) and (d), **qubit** tunneling events are visible as abrupt changes in the cavity spectrum....(Color online) (a) Cavity spectroscopy (design A ) while sweeping the cavity flux bias with the **qubit** far detuned, biased at its maximum **frequency**. The solid line is a fit to the model including the junction capacitance. (b) Zoom-in near the maximum cavity **frequency** showing a slot-mode. (c) Line-cut on resonance along the dashed line in (b) with a fit to a skewed Lorentzian (solid line)....In general, rf SQUID phase **qubits** have lower T 2 * (and T 2 ) values than transmons, specifically at lower **frequencies**, where d f 01 / d φ q is large and therefore the **qubit** is quite sensitive to bias fluctuations and 1/f flux noise . For example, 600 MHz higher in **qubit** **frequency**, at f 01 = 7.98 GHz, Ramsey **oscillations** gave T 2 * = 223 ns. At this location, the decay of on-resonance Rabi **oscillations** gave T ' = 727 ns, a separate measurement of **qubit** energy decay after a π -pulse gave T 1 = 658 ns, and so, T 2 ≈ 812 ns, or T 2 ≈ 3.6 × T 2 * , a small, but noticeable improvement over the lower **frequency** results displayed Fig. Fig6. The current device designs suffer from their planar geometry, due to a very large area enclosed by the non-gradiometric rf SQUID loop (see Fig. Fig1). Future devices will require some form of protection against flux noise , possibly gradiometric loops or replacing the large geometric inductors with a much smaller series array of Josephson junctions ....(Color online) Coupling rate 2 g / 2 π (design A ) as a function of cavity **frequency** ω c / 2 π . The solid red (blue) line is the prediction from Eq. ( eq:g) (including L x and C J ’s). The (dotted) dashed line is the prediction for capacitive coupling with C = 15 fF ( C = 5 fF). The solid circles were measured spectroscopically (see text). At lowest cavity **frequency**, the solid ⋆ results from a fit to the Purcell data, discussed later in section TCQEDC. The gray region highlights where the phase **qubit** (design A ) remains stable enough for operation (see text)....Next, we carefully explore the size of the dispersive shifts for various cavity and **qubit** **frequencies**. In order to capture the maximum dispersive **frequency** shift experienced by the cavity, we applied a π -pulse to the **qubit**. A fit to the phase response curve allows us to extract the cavity’s amplitude response time 2 / κ , the **qubit** T 1 , and the full dispersive shift 2 χ . Changing the cavity **frequency** modifies the coupling g and the detuning Δ 01 , while changes to the **qubit** **frequency** change both Δ 01 and the ** qubit’s** anharmonicity α . In Fig. Fig9(a), we show the phase

**anharmonicity as a function of its transition**

**qubit**’s**frequency**ω 01 / 2 π extracted from the spectroscopic data shown in Fig. Fig5 from section QBB for design A . The solid red line is a polynomial fit to the experimental data, used to calculate the three-level model curves in Fig. Fig9(b–d), while the blue line is a theoretical prediction of the relative anharmonicity (including L x , but neglecting C J ) using perturbation theory and the characteristic

**qubit**parameters extracted section QBB. In Fig. Fig9(b–d), we find that the observed dispersive shifts strongly depend on all of these factors and agree well with the three-level model predictions . For comparison, in Fig. Fig9(b), we show the results for the two-level system model (bold dashed line) when f c = 6.58 GHz, which has a significantly larger amplitude for all detunings (outside the “straddling regime”). Notice that it is possible to increase the size of the dispersive shifts for a given | Δ 01 | / ω 01 by decreasing the cavity

**frequency**f c , which increases the coupling rate 2 g / 2 π (as seen in Fig. Fig2 in section TCQED). Also, notice that decreasing the ratio of | Δ 01 | / ω 01 also significantly increases the size of the dispersive shifts, even when the phase

**relative anharmonicity α r decreases as ω 01 increases. Essentially, the ability to reduce | Δ 01 | helps to counteract any reductions in α r . These results clearly demonstrate the ability to tune the size of the dispersive shift through selecting the relative**

**qubit**’s**frequency**of the

**qubit**and the cavity. This tunability offers a new flexibility for optimizing dispersive readout of

**qubits**in cavity QED architectures and provides a way for rf SQUID phase

**qubits**to avoid the destructive effects of tunneling-based measurements....(Color online) (a) Time domain measurements (design A ). Rabi

**oscillations**for

**frequencies**near f 01 = 7.38 GHz. (b) Line-cut on-resonance along the dashed line in (a). The fit (solid line) yields a Rabi

**oscillation**decay time of T ' = 409 ns. (c) Ramsey

**oscillations**versus

**qubit**flux detuning near f 01 = 7.38 GHz. (d) Line-cut along the dashed line in (c). The fit (solid line) yields a Ramsey decay time of T 2 * = 106 ns. With T 1 = 600 ns, this implies a phase coherence time T 2 = 310 ns....(Color online) (a) Pulse sequence. (b) Rabi

**oscillations**(design A ) for various pulse durations obtained using dispersive measurement at f 01 = 7.18 GHz, with Δ 01 = + 10 g . (c) A single, averaged time trace along the vertical dashed line in (b). (d) Rabi

**oscillations**extracted from the final population at the end of the drive pulse, along the dashed diagonal line in (b). (e) Zoom-in of dashed box in (b) showing Rabi

**oscillations**observed during continuous driving....We can explore the coupled

**qubit**-cavity behavior described by Eq. ( eq:H) by performing spectroscopic measurements on either the

**qubit**or the cavity near the resonance condition, ω 01 = ω c . Fig. Fig7(a) shows

**qubit**spectroscopy for design A overlaid with cavity spectroscopy for two cavity

**frequencies**, f c = 6.58 GHz and 6.78 GHz. Fig. Fig7(d) shows cavity spectroscopy for design B with the cavity at its maximum

**frequency**of f c m a x = 7.07 GHz while sweeping the

**qubit**flux bias φ q . In both cases, when the

**qubit**

**frequency**f 01 is swept past the cavity resonance, the inductive coupling generates the expected spectroscopic normal-mode splitting....The weak additional splitting just below the cavity in Fig. Fig7(d) is from a resonant slot-mode. We can determine the coupling rate 2 g / 2 π between the

**qubit**and the cavity by extracting the splitting size as a function of cavity

**frequency**f c from the measured spectra. Three examples of fits are shown in Fig. Fig7(b–d) with solid lines representing the bare

**qubit**and cavity

**frequencies**, whereas the dashed lines show the new coupled normal-mode

**frequencies**. For design A ( B ), at the maximum cavity

**frequency**of 6.78 GHz (7.07 GHz), we found a minimum coupling rate of 2 g m i n / 2 π = 78 MHz (104 MHz). Notice that the splitting size is clearer bigger in Fig. Fig7(c) than for Fig. Fig7(b) by about 25 MHz. The results for the coupling rate 2 g / 2 π as a function of ω c / 2 π for design A were shown in Fig. Fig2 in section TCQED. Also visible in Fig. Fig7(c–d) are periodic, discontinuous jumps in the cavity spectrum. These are indicative of

**qubit**tunneling events between adjacent metastable energy potential minima, typical behavior for hysteretic rf SQUID phase

**qubits**. Moving away from the maximum cavity

**frequency**increases the flux sensitivity, with the

**qubit**tunneling events becoming more visible as steps. This behavior is clearly visible in Fig. Fig7(c) and was already shown in Fig. Fig4 in Sec. QBA and, as discussed there, provides a convenient way to perform rapid microwave readout of traditional tunneling measurements . Next, we describe dispersive measurements of the phase

**qubit**for design A . These results agree with the tunneling measurements across the entire

**qubit**spectrum. ... We describe a tunable-cavity QED architecture with an rf SQUID phase

**qubit**inductively coupled to a single-mode, resonant cavity with a tunable

**frequency**that allows for both microwave readout of tunneling and dispersive measurements of the

**qubit**. Dispersive measurement is well characterized by a three-level model, strongly dependent on

**qubit**anharmonicity,

**qubit**-cavity coupling and detuning. A tunable cavity

**frequency**provides a way to strongly vary both the

**qubit**-cavity detuning and coupling strength, which can reduce Purcell losses, cavity-induced dephasing of the

**qubit**, and residual bus coupling for a system with multiple

**qubits**. With our

**qubit**-cavity system, we show that dynamic control over the cavity

**frequency**enables one to avoid Purcell losses during coherent

**qubit**evolutions and optimize state readout during

**qubit**measurements. The maximum

**qubit**decay time $T_1$ = 1.5 $\mu$s is found to be limited by surface dielectric losses from a design geometry similar to planar transmon

**qubits**.

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Contributors: Catelani, G., Schoelkopf, R. J., Devoret, M. H., Glazman, L. I.

Date: 2011-06-04

A further test of the theory presented in Sec. sec:th_s is provided by the measurement of the **qubit** resonant **frequency**. In the semiclassical regime of small E C , the **qubit** can be described by the effective circuit of Fig. fig1(b), with the junction admittance Y J of Eq. ( YJ), Y C = i ω C , and Y L = 1 / i ω L [the inductance is related to the inductive energy by E L = Φ 0 / 2 π 2 / L ]. As discussed in Ref. ...As a second example of a strongly anharmonic system, we consider here a flux **qubit**, i.e., in Eq. ( Hphi) we assume E J > E L and take the external flux to be close to half the flux quantum, Φ e ≈ Φ 0 / 2 . Then the potential has a double-well shape and the flux **qubit** ground states | - and excited state | + are the lowest tunnel-split eigenstates in this potential, see Fig. fig:fl_q. The non-linear nature of the sin ϕ ̂ / 2 **qubit**-quasiparticle coupling in Eq. ( HTle) has a striking effect on the transition rate Γ + - , which vanishes at Φ e = Φ 0 / 2 due to destructive interference: for flux biased at half the flux quantum the **qubit** states | - , | + are respectively symmetric and antisymmetric around ϕ = π , while the potential in Eq. ( Hphi) and the function sin ϕ / 2 in Eq. ( wif_gen) are symmetric. Note that the latter symmetry and its consequences are absent in the environmental approach in which a linear phase-quasiparticle coupling is assumed....The transmon low-energy spectrum is characterized by well separated [by the plasma **frequency** ω p , Eq. ( pl_fr)] and nearly degenerate levels whose energies, as shown in Fig. fig:trans, vary periodically with the gate voltage n g . Here we derive the asymptotic expression (valid at large E J / E C ) for the energy splitting between the nearly degenerate levels. We consider first the two lowest energy states and then generalize the result to higher energies....Schematic representation of the transmon low energy spectrum as function of the dimensionless gate voltage n g . Solid (dashed) lines denotes even (odd) states (see also Sec. sec:cpb). The amplitudes of the **oscillations** of the energy levels are exponentially small, see Appendix app:eosplit; here they are enhanced for clarity. Quasiparticle tunneling changes the parity of the **qubit** sate. The results of Sec. sec:semi are valid for transitions between states separated by energy of the order of the plasma **frequency** ω p , Eq. ( pl_fr), and give, for example, the rate Γ 1 0 . For the transition rates between nearly degenerate states of opposite parity, such as Γ o e 1 , see Appendix app:eorate....As an application of the general approach described in the previous section, we consider here a weakly anharmonic **qubit**, such as the transmon and phase **qubits**. We start with the the semiclassical limit, i.e., we assume that the potential energy terms in Eq. ( Hphi) dominate the kinetic energy term proportional to E C . This limit already reveals a non-trivial dependence of relaxation on flux. Note that assuming E L ≠ 0 we can eliminate n g in Eq. ( Hphi) by a gauge transformation. In the transmon we have E L = 0 and the spectrum depends on n g , displaying both well separated and nearly degenerate states, see Fig. fig:trans. The results of this section can be applied to the single-junction transmon when considering well separated states. The transition rate between these states and the corresponding **frequency** shift are dependent on n g . However, since E C ≪ E J this dependence introduces only small corrections to Γ n n - 1 and δ ω ; the corrections are exponential in - 8 E J / E C . By contrast, the leading term in the rate of transitions Γ e ↔ o between the even and odd states is exponentially small. The rate Γ e ↔ o of parity switching is discussed in detail in Appendix app:eorate....Potential energy (in units of E L ) for a flux **qubit** biased at Φ e = Φ 0 / 2 with E J / E L = 10 . The horizontal lines represent the two lowest energy levels, with energy difference ϵ ̄ given in Eq. ( e0_eff)....As low-loss non-linear elements, Josephson junctions are the building blocks of superconducting **qubits**. The interaction of the **qubit** degree of freedom with the quasiparticles tunneling through the junction represent an intrinsic relaxation mechanism. We develop a general theory for the **qubit** decay rate induced by quasiparticles, and we study its dependence on the magnetic flux used to tune the **qubit** properties in devices such as the phase and flux **qubits**, the split transmon, and the fluxonium. Our estimates for the decay rate apply to both thermal equilibrium and non-equilibrium quasiparticles. We propose measuring the rate in a split transmon to obtain information on the possible non-equilibrium quasiparticle distribution. We also derive expressions for the shift in **qubit** **frequency** in the presence of quasiparticles....which has the same form of the Hamiltonian for the single junction transmon [i.e., Eq. ( Hphi) with E L = 0 ] but with a flux-dependent Josephson energy, Eq. ( EJ_flux). Therefore the spectrum follows directly from that of the single junction transmon (see Fig. fig:trans) and consists of nearly degenerate and well separated states. The energy difference between well separated states is approximately given by the flux-dependent **frequency** [cf. Eq. ( pl_fr)]...(a) Schematic representation of a **qubit** controlled by a magnetic flux, see Eq. ( Hphi). (b) Effective circuit diagram with three parallel elements – capacitor, Josephson junction, and inductor – characterized by their respective admittances. ... As low-loss non-linear elements, Josephson junctions are the building blocks of superconducting **qubits**. The interaction of the **qubit** degree of freedom with the quasiparticles tunneling through the junction represent an intrinsic relaxation mechanism. We develop a general theory for the **qubit** decay rate induced by quasiparticles, and we study its dependence on the magnetic flux used to tune the **qubit** properties in devices such as the phase and flux **qubits**, the split transmon, and the fluxonium. Our estimates for the decay rate apply to both thermal equilibrium and non-equilibrium quasiparticles. We propose measuring the rate in a split transmon to obtain information on the possible non-equilibrium quasiparticle distribution. We also derive expressions for the shift in **qubit** **frequency** in the presence of quasiparticles.

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Contributors: Shahriar, M. S., Pradhan, Prabhakar

Date: 2002-12-19

Left: Schematic illustration of an experimental arrangement for measuring the phase dependence of the population of the excited state | 1 : (a) The microwave field couples the ground state ( | 0 ) to the excited state ( | 1 ). A third level, | 2 , which can be coupled to | 1 optically, is used to measure the population of | 1 via fluorescence detection. (b) The microwave field is turned on adiabatically with a switching time-constant τ s w , and the fluorescence is monitored after a total interaction time of τ . Right: Illustration of the Bloch-Siegert **Oscillation** (BSO): (a) The population of state | 1 , as a function of the interaction time τ , showing the BSO superimposed on the conventional Rabi **oscillation**. (b) The BSO **oscillation** (amplified scale) by itself, produced by subtracting the Rabi **oscillation** from the plot in (a). (c) The time-dependence of the Rabi **frequency**. Inset: BSO as a function of the absolute phase of the field with fixed τ ....We show that if the Rabi **frequency** is comparable to the Bohr **frequency** so that the rotating wave approximation is inappropriate, an extra **oscillation** is present with the Rabi **oscillation**. We discuss how the sensitivity of the degree of excitation to the phase of the field may pose severe constraints on precise rotations of quantum bits involving low-**frequency** transitions. We present a scheme for observing this effect in an atomic beam. ... We show that if the Rabi **frequency** is comparable to the Bohr **frequency** so that the rotating wave approximation is inappropriate, an extra **oscillation** is present with the Rabi **oscillation**. We discuss how the sensitivity of the degree of excitation to the phase of the field may pose severe constraints on precise rotations of quantum bits involving low-**frequency** transitions. We present a scheme for observing this effect in an atomic beam.

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Contributors: Vierheilig, Carmen, Bercioux, Dario, Grifoni, Milena

Date: 2010-10-22

We consider a **qubit** coupled to a nonlinear quantum **oscillator**, the latter coupled to an Ohmic bath, and investigate the **qubit** dynamics. This composed system can be mapped onto that of a **qubit** coupled to an effective bath. An approximate mapping procedure to determine the spectral density of the effective bath is given. Specifically, within a linear response approximation the effective spectral density is given by the knowledge of the linear susceptibility of the nonlinear quantum **oscillator**. To determine the actual form of the susceptibility, we consider its periodically driven counterpart, the problem of the quantum Duffing **oscillator** within linear response theory in the driving amplitude. Knowing the effective spectral density, the **qubit** dynamics is investigated. In particular, an analytic formula for the **qubit**'s population difference is derived. Within the regime of validity of our theory, a very good agreement is found with predictions obtained from a Bloch-Redfield master equation approach applied to the composite **qubit**-nonlinear **oscillator** system....model We consider a composed system built of a **qubit**, -the system of interest-, coupled to a nonlinear quantum **oscillator** (NLO), see Fig. linearbath. To read-out the **qubit** state we couple the **qubit** linearly to the **oscillator** with the coupling constant g ¯ , such that via the intermediate NLO dissipation also enters the **qubit** dynamics....Jeff The effective spectral density follows from Eqs. ( gl20) and ( chilarger). It reads: J s i m p l J e f f ω e x = g ¯ 2 γ ω e x n 1 0 4 2 Ω 1 | ω e x | + Ω 1 M γ 2 Ω 1 2 2 n t h Ω 1 + 1 2 n 1 0 4 + 4 M Ω 2 | ω e x | - Ω 1 2 . As in case of the effective spectral density J e f f H O , Eq. ( linearspecdens), we observe Ohmic behaviour at low **frequency**. In contrast to the linear case, the effective spectral density is peaked at the shifted **frequency** Ω 1 . Its shape approaches the Lorentzian one of the linear effective spectral density, but with peak at the shifted **frequency**, as shown in Fig. CompLorentz....Schematic representation of the complementary approaches available to evaluate the **qubit** dynamics: In the first approach one determines the eigenvalues and eigenfunctions of the composite **qubit** plus **oscillator** system (yellow (light grey) box) and accounts afterwards for the harmonic bath characterized by the Ohmic spectral density J ω . In the effective bath description one considers an environment built of the harmonic bath and the nonlinear **oscillator** (red (dark grey) box). In the harmonic approximation the effective bath is fully characterized by its effective spectral density J e f f ω . approachschaubild...mapping The main aim is to evaluate the ** qubit’s** evolution described by q t . This can be achieved within an effective description using a mapping procedure. Thereby the

**oscillator**and the Ohmic bath are put together, as depicted in Figure approachschaubild, to form an effective bath. The effective Hamiltonian...The transition

**frequencies**in Eqs. ( rc1) and ( rc2) coincide, and in Figs. Plowg and Flowg there is no deviation observed when comparing the three different approaches....where the trace over the degrees of freedom of the bath and of the

**oscillator**is taken. In Fig. approachschaubild two different approaches to determine the

**qubit**dynamics are depicted. In the first approach, which is elaborated in Ref. [...Corresponding Fourier transform of P t shown in Fig. CompNLLP. The effect of the nonlinearity is to increase the resonance

**frequencies**with respect to the linear case. As a consequence the relative peak heights change. CompNLLF...Schematic representation of the composed system built of a

**qubit**, an intermediate nonlinear

**oscillator**and an Ohmic bath. linearbath ... We consider a

**qubit**coupled to a nonlinear quantum

**oscillator**, the latter coupled to an Ohmic bath, and investigate the

**qubit**dynamics. This composed system can be mapped onto that of a

**qubit**coupled to an effective bath. An approximate mapping procedure to determine the spectral density of the effective bath is given. Specifically, within a linear response approximation the effective spectral density is given by the knowledge of the linear susceptibility of the nonlinear quantum

**oscillator**. To determine the actual form of the susceptibility, we consider its periodically driven counterpart, the problem of the quantum Duffing

**oscillator**within linear response theory in the driving amplitude. Knowing the effective spectral density, the

**qubit**dynamics is investigated. In particular, an analytic formula for the

**qubit**'s population difference is derived. Within the regime of validity of our theory, a very good agreement is found with predictions obtained from a Bloch-Redfield master equation approach applied to the composite

**qubit**-nonlinear

**oscillator**system.

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Contributors: Wallraff, A., Schuster, D. I., Blais, A., Frunzio, L., Majer, J., Girvin, S. M., Schoelkopf, R. J.

Date: 2005-02-27

(color online) (a) Rabi **oscillations** in the **qubit** population P vs. Rabi pulse length Δ t (blue dots) and fit with unit visibility (red line). (b) Measured Rabi **frequency** ν R a b i vs. pulse amplitude ϵ s (blue dots) and linear fit....(color online) Measurement response φ (blue lines) and theoretical prediction (red lines) vs. time. At t = 6 μ s (a) a π pulse, (b) a 2 π pulse, and (c) a 3 π pulse is applied to the **qubit**. In each panel the dashed lines correspond to the expected measurement response in the ground state φ , in the saturated state φ = 0 , and in the excited state φ ....The extracted **qubit** population P is plotted versus Δ t in Fig. fig:rabioscillationsa. We observe a visibility of 95 ± 6 % in the Rabi **oscillations** with error margins determined from the residuals of the experimental P with respect to the predicted values. Thus, in a measurement of Rabi **oscillations** in a superconducting **qubit**, a visibility in the population of the **qubit** excited state that approaches unity is observed for the first time. Moreover, we note that the decay in the Rabi **oscillation** amplitude out to pulse lengths of 100 n s is very small and consistent with the long T 1 and T 2 times of this charge **qubit**, see Fig. fig:rabioscillationsa and Ramsey experiment discussed below. We have also verified the expected linear scaling of the Rabi **oscillation** **frequency** ν R a b i with the pulse amplitude ϵ s ∝ n s , see Fig. fig:rabioscillationsb....In our circuit QED architecture , a Cooper pair box , acting as a two level system with ground and excited states and level separation E a = ℏ ω a = E e l 2 + E J 2 is coupled capacitively to a single mode of the electromagnetic field of a transmission line resonator with resonance **frequency** ω r , see Fig. fig:setupa. As demonstrated for this system, the electrostatic energy E e l and the Josephson energy E J of the split Cooper pair box can be controlled in situ by a gate voltage V g and magnetic flux Φ , see Fig. fig:setupa. In the resonant ( ω a = ω r ) strong coupling regime a single excitation is exchanged coherently between the Cooper pair box and the resonator at a rate g / π , also called the vacuum Rabi **frequency** . In the non-resonant regime ( Δ = ω a - ω r > g ) the capacitive interaction gives rise to a dispersive shift g 2 / Δ σ z in the resonance **frequency** of the cavity which depends on the **qubit** state σ z , the coupling g and the detuning Δ . We have suggested that this shift in resonance **frequency** can be used to perform a quantum non-demolition (QND) measurement of the **qubit** state . With this technique we have recently measured the ground state response and the excitation spectrum of a Cooper pair box ....(color online) (a) Measured Ramsey fringes (blue dots) observed in the **qubit** population P vs. pulse separation Δ t using the pulse sequence shown in Fig. fig:setupb and fit of data to sinusoid with gaussian envelope (red line). (b) Measured dependence of Ramsey **frequency** ν R a m s e y on detuning Δ a , s of drive **frequency** (blue dots) and linear fit (red line)....(color online) (a) Simplified circuit diagram of measurement setup. A Cooper pair box with charging energy E C and Josephson energy E J is coupled through capacitor C g to a transmission line resonator, modelled as parallel combination of an inductor L and a capacitor C . Its state is determined in a phase sensitive heterodyne measurement of a microwave transmitted at **frequency** ω R F through the circuit, amplified and mixed with a local **oscillator** at **frequency** ω L O . The Cooper pair box level separation is controlled by the gate voltage V g and flux Φ . Its state is coherently manipulated using microwaves at **frequency** ω s with pulse shapes determined by V p . (b) Measurement sequence for Rabi **oscillations** with Rabi pulse length Δ t , pulse **frequency** ω s and amplitude ∝ n s with continuous measurement at **frequency** ω R F and amplitude ∝ n R F . (c) Sequence for Ramsey fringe experiment with two π / 2 -pulses at ω s separated by a delay Δ t and followed by a pulsed measurement....In a Rabi **oscillation** experiment with a superconducting **qubit** we show that a visibility in the **qubit** excited state population of more than 90 % can be attained. We perform a dispersive measurement of the **qubit** state by coupling the **qubit** non-resonantly to a transmission line resonator and probing the resonator transmission spectrum. The measurement process is well characterized and quantitatively understood. The **qubit** coherence time is determined to be larger than 500 ns in a measurement of Ramsey fringes....In the experiments presented here, we coherently control the quantum state of a Cooper pair box by applying to the **qubit** microwave pulses of **frequency** ω s , which are resonant with the **qubit** transition **frequency** ω a / 2 π ≈ 4.3 G H z , through the input port C i n of the resonator, see Fig. fig:setupa. The microwaves drive Rabi **oscillations** in the **qubit** at a **frequency** of ν R a b i = n s g / π , where n s is the average number of drive photons within the resonator. Simultaneously, we perform a continuous dispersive measurement of the **qubit** state by determining both the phase and the amplitude of a coherent microwave beam of **frequency** ω R F / 2 π = ω r / 2 π ≈ 5.4 G H z transmitted through the resonator . The phase shift φ = tan -1 2 g 2 / κ Δ σ z is the response of our meter from which we determine the **qubit** populat...We have determined the coherence time of the Cooper pair box from a Ramsey fringe experiment, see Fig. fig:setupc, when biased at the charge degeneracy point where the energy is first-order insensitive to charge noise . To avoid dephasing induced by a weak continuous measurement beam we switch on the measurement beam only after the end of the second π / 2 pulse. The resulting Ramsey fringes **oscillating** at the detuning **frequency** Δ a , s = ω a - ω s ∼ 6 M H z decay with a long coherence time of T 2 ∼ 500 n s , see Fig. fig:Ramseya. The corresponding **qubit** phase quality factor of Q ϕ = T 2 ω a / 2 ∼ 6500 is similar to the best values measured so far in **qubit** realizations biased at such an optimal point . The Ramsey **frequency** is shown to depend linearly on the detuning Δ a , s , as expected, see Fig. fig:Ramseyb. We note that a measurement of the Ramsey **frequency** is an accurate time resolved method to determine the **qubit** transition **frequency** ω a = ω s + 2 π ν R a m s e y . ... In a Rabi **oscillation** experiment with a superconducting **qubit** we show that a visibility in the **qubit** excited state population of more than 90 % can be attained. We perform a dispersive measurement of the **qubit** state by coupling the **qubit** non-resonantly to a transmission line resonator and probing the resonator transmission spectrum. The measurement process is well characterized and quantitatively understood. The **qubit** coherence time is determined to be larger than 500 ns in a measurement of Ramsey fringes.

Data types:

Contributors: Chen, Yu, Sank, D., O'Malley, P., White, T., Barends, R., Chiaro, B., Kelly, J., Lucero, E., Mariantoni, M., Megrant, A.

Date: 2012-09-09

(Color online) (a) Schematic representation of **qubit** projective measurement, where a current pulse allows a **qubit** in the excited state | e to tunnel to the right well ( R ), while a **qubit** in the ground state | g stays in the left well ( L ). (b) Readout circuit, showing lumped-element L R - C R readout resonator inductively coupled to the **qubit**, with Josephson junction effective inductance L J and capacitance C , with loop inductance L . **Qubit** control is through the differential flux bias line ( F B ). The readout resonator is capacitively coupled through C c to the readout line, in parallel with the other readout resonators. The readout line is connected through a cryogenic circulator to a low-noise cryogenic amplifier and to a room temperature microwave source. (c) Photomicrograph of four-**qubit** sample. F B 1 - 4 are control lines for each **qubit** and R R is the resonator readout line. Inset shows details for one **qubit** and its readout resonator. Scale bar is 50 μ m in length; fig.setup...(Color online) (a) Phase of signal reflected from readout resonator, as a function of the probe microwave **frequency** (averaged 900 times), for the **qubit** in the left ( L , blue) and right ( R , red) wells. Dashed line shows probe **frequency** for maximum visibility. (b) Reflected phase as a function of **qubit** flux bias, with no averaging. See text for details. fig.phase...(Color online) Setup for **frequency**-multiplexed readout. Multiplexed readout signals I p and Q p from top FGPA-DAC board are up-converted by mixing with a fixed microwave tone, then pass through the circulator into the **qubit** chip. Reflected signals pass back through the circulator, through the two amplifiers G 1 and G 2 , and are down-converted into I r and Q r using the same microwave tone, and are then processed by the bottom ADC-FPGA board. Data in the shadowed region are the down-converted I r and Q r spectra output from the ADC-FPGA board; probe signals from the FPGA-DAC board have the same **frequency** spectrum. D C indicates the digital demodulation channels, each processed independently and sent to the computer. fig.measure...With the bias points chosen for each **qubit**, we demonstrated the **frequency**-multiplexed readout by performing a multi-**qubit** experiment. To minimize crosstalk, we removed the coupling capacitors between **qubits** used in Ref. 6. In this experiment, we drove Rabi **oscillations** on each ** qubit’s** | g ↔ | e transition and read out the

**qubit**states simultaneously. We first calibrated the pulse amplitude needed for each

**qubit**to perform a | g → | e Rabi transition in 10 ns. The drive amplitude was then set to 1, 2/3, 1/2 and 2/5 the calibrated Rabi transition amplitude for

**qubits**Q 1 to Q 4 respectively, so that the Rabi period was 20 ns, 30 ns, 40 ns and 50 ns for

**qubits**Q 1 to Q 4 . We then drove each

**qubit**separately using an on-resonance Rabi drive for a duration τ , followed immediately by a projective measurement and

**qubit**state readout. This experiment yielded the measurements shown in Fig. fig.rabi(a)-(d) for

**qubits**Q 1 - Q 4 respectively....We introduce a

**frequency**-multiplexed readout scheme for superconducting phase

**qubits**. Using a quantum circuit with four phase

**qubits**, we couple each

**qubit**to a separate lumped-element superconducting readout resonator, with the readout resonators connected in parallel to a single measurement line. The readout resonators and control electronics are designed so that all four

**qubits**can be read out simultaneously using

**frequency**multiplexing on the one measurement line. This technology provides a highly efficient and compact means for reading out multiple

**qubits**, a significant advantage for scaling up to larger numbers of

**qubits**....(Color online) (a)-(d) Rabi

**oscillations**for

**qubits**Q 1 - Q 4 respectively, with the

**qubits**driven with 1, 2/3, 1/2 and 2/5 the on-resonance drive amplitude needed to perform a 10 ns Rabi | g → | e transition. (e) Rabi

**oscillations**measured simultaneously for all the

**qubits**, using the same color coding and drive amplitudes as for panels (a)-(d). fig.rabi...We demonstrated the multiplexed readout using a quantum circuit comprising four phase

**qubits**and five integrated resonators, shown in Fig. fig.setup(c). The design of this chip is similar to that used for a recent implementation of Shor’s algorithm,, but here the

**qubits**were read out off a single line using microwave reflectometry, replacing the SQUID readout used Ref 6. . This dramatically simplifies the chip design and significantly reduces the footprint of the quantum circuit. We designed the readout resonators so that they resonated at

**frequencies**of 3-4 GHz (far de-tuned from the

**qubit**| g ↔ | e transition

**frequency**of 6-7 GHz), with loaded resonance linewidths of a few hundred kilohertz. This allows us to use

**frequency**multiplexing, which has been successfully used in the readout of microwave kinetic inductance detectors as well as other types of

**qubits**. Combined with custom GHz-

**frequency**signal generation and acquisition boards, this approach provides a compact and efficient readout scheme that should be applicable to systems with 10-100

**qubits**using a single readout line, with sufficient measurement bandwidth for microsecond-scale readout times....The calibration of the readout process was done in two steps. We first optimized the microwave probe

**frequency**to maximize the signal difference between the left and right well states. This was performed by measuring the reflected phase φ as a function of the probe

**frequency**, with the

**qubit**prepared first in the left and then in the right well. In Fig. fig.phase(a), we show the result with the

**qubit**flux bias set to 0.15 Φ 0 , where the difference in L J in two well states was relatively large. The probe

**frequency**that maximized the signal difference was typically mid-way between the loaded resonator

**frequencies**for the

**qubit**in the left and right wells, marked by the dashed line in Fig. fig.phase(a). We typically obtained resonator

**frequency**shifts as large as ∼ 150 kHz for the

**qubit**between the two wells, as shown in Fig. fig.phase(a), significantly larger than the resonator linewidth....With the probe

**frequency**set in the first step, the flux bias was then set to optimize the readout. As illustrated in Fig. fig.phase(b), the optimization was performed by measuring the resonator’s reflected phase as a function of

**qubit**bias flux, at the optimal probe

**frequency**, 3.70415 GHz in this case. The

**qubit**was initialized by setting the flux to its negative “reset” value (position I), where the

**qubit**potential has only one minimum. The flux was then increased to an intermediate value Φ , placing the

**qubit**state in the left well, and the reflection phase measured with a 5 μ s microwave probe signal (blue data). The flux was then set to its positive reset value (position V), then brought back to the same flux value Φ , placing the

**qubit**state in the right well, and the reflection phase again measured with a probe signal (red data). Between the symmetry point III ( Φ = 0.5 ) and the regions with just one potential minimum ( Φ ≤ 0.1 or Φ ≥ 0.9 ), the

**qubit**inductance differs between the left and right well states, which gives rise to the difference in phase for the red and blue data measured at the same flux. This difference increases for the flux bias closer to the single-well region, which can give a signal-to-noise ratio as high as 30 at ambient readout microwave power. The optimal flux bias was then set to a value where the readout had a high signal-to-noise ratio (typically > 5), but with a potential barrier sufficient to prevent spurious readout-induced switching between the potential wells. Several iterations were needed to optimize both the probe

**frequency**and flux bias....With each

**qubit**individually characterized, we then excited and measured all four

**qubits**simultaneously, as shown in Fig. fig.rabi(e). There is no measurable difference between the individually-measured Rabi

**oscillations**in panels (a)-(d) compared to the multiplexed readout in panel (e). ... We introduce a

**frequency**-multiplexed readout scheme for superconducting phase

**qubits**. Using a quantum circuit with four phase

**qubits**, we couple each

**qubit**to a separate lumped-element superconducting readout resonator, with the readout resonators connected in parallel to a single measurement line. The readout resonators and control electronics are designed so that all four

**qubits**can be read out simultaneously using

**frequency**multiplexing on the one measurement line. This technology provides a highly efficient and compact means for reading out multiple

**qubits**, a significant advantage for scaling up to larger numbers of

**qubits**.

Data types:

Contributors: Agarwal, S., Rafsanjani, S. M. Hashemi, Eberly, J. H.

Date: 2012-01-13

Numerical and analytical evaluation of the entanglement dynamics between the two **qubits** for ω 0 = 0.15 ω , β = 0.16 and α = 3 . Entanglement between the **qubits** exhibits collapse and revival. The analytic expression agrees well with the envelope of the numerically evaluated entanglement evolution....Here, within the adiabatic approximation, we extend the examination to the two-**qubit** case. Qualitative differences between the single-**qubit** and the multi-**qubit** cases are highlighted. In particular, we study the collapse and revival of joint properties of both the **qubits**. Entanglement properties of the system are investigated and it is shown that the entanglement between the **qubits** also exhibits collapse and revival. We derive what we believe are the first analytic expressions for the individual revival signals beyond the RWA, as well as analytic expression for the collapse and revival dynamics of entanglement. In the quasi-degenerate regime, the invalidity of the RWA in predicting the dynamical evolution will clearly be demonstrated in Sec. s.collapse_rev (see Figs. f.collapse_revival_double and f.collapse_revival_single)....The three potential wells corresponding to the states | 1 , 1 | N 1 (left), | 1 , 0 | N 0 (middle) and | 1 , - 1 | N -1 (right). The factor Δ X z p is the zero point fluctuation of a harmonic **oscillator**. For an **oscillator** of mass M and **frequency** ω the zero point fluctuation is given by Δ X z p = ℏ / 2 M ω ....Collapse and revival dynamics for ω 0 = 0.1 ω , β = 0.16 and α = 3 . The first two panels show analytic evaluations of (a.) one-**qubit** and (b.) two-**qubit** probability dynamics, and (c.) shows that the two-**qubit** analytic formula matches well to the corresponding numerical evolution. In each case the initial state is a product of a coherent **oscillator** state with the lowest of the S x states. Note the breakup in the main revival peak of the two-**qubit** numerical evaluation, which comes from the ω - 2 ω beat note, not included in the analytic calculation, and not present for a single **qubit**....When squared, the probability shows two **frequencies** of **oscillation**, 2 Ω N ω and 2 2 Ω N ω . Since three new basis states are involved, we could expect three **frequencies**, but two are equal: | E N + - E N 0 | = | E N - - E N 0 | . This is in contrast to the single-**qubit** case where only one Rabi **frequency** determines the evolution . We show below in Fig. f.col_rev the way differences between one and a pair of **qubits** can be seen....If the average excitation of the **oscillator**, n ̄ = α 2 , is large one can evaluate the above sum approximately (see Appendix) and obtain analytic expressions and graphs of the evolution, as shown in Fig. f.col_rev. As expected, because of the double **frequency** in ( eqnP(t)), the revival time for S t 2 ω 0 is half the revival time for S t ω 0 . Thus there are two different revival sequences in the time series. Appropriate analytic formulas, e.g., ( a.Phi), agree well with the numerically evaluated evolution even for the relatively weak coherent excitation, α = 3 ....Entanglement dynamics between the **qubits**. (a) ω 0 = 0.1 ω , β = 0.16 , (b) ω 0 = 0.15 ω , β = 0.16 , (c) ω 0 = 0.1 ω , β = 0.2 and (d) ω 0 = 0.15 ω , β = 0.2 . For all the figures, α = 3 ....Collapse and revival dynamics for P 1 2 , - 1 2 α t , given ω 0 = 0.15 ω , β = 0.16 and α = 3 . Note the single revival sequence. Also, note that there are no breakups in the revival peaks in contrast to the two-**qubit** case (Fig. f.collapse_revival_double). The RWA fails to describe the dynamical evolution even for the single **qubit** case....With recent advances in the area of circuit QED, it is now possible to engineer systems for which the **qubits** are coupled to the **oscillator** so strongly, or are so far detuned from the **oscillator**, that the RWA cannot be used to describe the system’s evolution correctly . The parameter regime for which the coupling strength is strong enough to invalidate the RWA is called the ultra-strong coupling regime . Niemczyk, et al. and Forn-Díaz, et al. have been able to experimentally achieve ultra-strong coupling strengths and have demonstrated the breakdown of the RWA. Motivated by these experimental developments and the importance of understanding collective quantum behavior, we investigate a two-**qubit** TC model beyond the validity regime of RWA. The regime of parameters we will be concerned with is the regime where the **qubits** are quasi-degenerate, i.e., with **frequencies** much smaller than the **oscillator** **frequency**, ω 0 ≪ ω , while the coupling between the **qubits** and the **oscillator** is allowed to be an appreciable fraction of the **oscillator** **frequency**. In this parameter regime, the dynamics of the system can neither be correctly described under the RWA, nor can the effects of the counter rotating terms be taken as a perturbative correction to the dynamics predicted within the RWA by including higher powers of β . For illustration, systems are shown in Fig. f.model for which the RWA is valid, or breaks down, because the condition ω 0 ≈ ω is valid, or is violated. The regime that we will be interested in, for which ω 0 ≪ ω , is shown on the right....There is only one revival sequence for the single **qubit** system as a consequence of having only one Rabi **frequency** in the single **qubit** case. The analytic and numerically exact evolution of P 1 2 , - 1 2 α t is plotted in Fig. f.collapse_revival_single. The single revival sequence is evident from the figure. A discussion on the multiple revival sequences for the K -**qubit** TC model, within the parameter regime where the RWA is valid, can be found in ....The Tavis-Cummings model for more than one **qubit** interacting with a common **oscillator** mode is extended beyond the rotating wave approximation (RWA). We explore the parameter regime in which the **frequencies** of the **qubits** are much smaller than the **oscillator** **frequency** and the coupling strength is allowed to be ultra-strong. The application of the adiabatic approximation, introduced by Irish, et al. (Phys. Rev. B \textbf{72}, 195410 (2005)), for a single **qubit** system is extended to the multi-**qubit** case. For a two-**qubit** system, we identify three-state manifolds of close-lying dressed energy levels and obtain results for the dynamics of intra-manifold transitions that are incompatible with results from the familiar regime of the RWA. We exhibit features of two-**qubit** dynamics that are different from the single **qubit** case, including calculations of **qubit**-**qubit** entanglement. Both number state and coherent state preparations are considered, and we derive analytical formulas that simplify the interpretation of numerical calculations. Expressions for individual collapse and revival signals of both population and entanglement are derived. ... The Tavis-Cummings model for more than one **qubit** interacting with a common **oscillator** mode is extended beyond the rotating wave approximation (RWA). We explore the parameter regime in which the **frequencies** of the **qubits** are much smaller than the **oscillator** **frequency** and the coupling strength is allowed to be ultra-strong. The application of the adiabatic approximation, introduced by Irish, et al. (Phys. Rev. B \textbf{72}, 195410 (2005)), for a single **qubit** system is extended to the multi-**qubit** case. For a two-**qubit** system, we identify three-state manifolds of close-lying dressed energy levels and obtain results for the dynamics of intra-manifold transitions that are incompatible with results from the familiar regime of the RWA. We exhibit features of two-**qubit** dynamics that are different from the single **qubit** case, including calculations of **qubit**-**qubit** entanglement. Both number state and coherent state preparations are considered, and we derive analytical formulas that simplify the interpretation of numerical calculations. Expressions for individual collapse and revival signals of both population and entanglement are derived.

Data types:

Contributors: Hua, Ming, Deng, Fu-Guo

Date: 2013-09-30

(color online) (a) The probability distribution of the two quantum Rabi **oscillations** ROT 0 (the blue solid line) and ROT 1 (the green solid line) of two charge **qubits** coupled to a resonator. (b) The probability distribution of the four quantum Rabi **oscillations** in our cc-phase gate on a three-charge-**qubit** system. Here, the blue-solid, green-solid, red-dashed, and Cambridge-blue-dot-dashed lines represent the quantum Rabi **oscillations** ROT 00 ( | 0 1 | 0 2 | 1 3 | 0 a ↔ | 0 1 | 0 2 | 0 3 | 1 a ), ROT 01 ( | 0 1 | 1 2 | 1 3 | 0 a ↔ | 0 1 | 1 2 | 0 3 | 1 a ), ROT 10 ( | 1 1 | 0 2 | 1 3 | 0 a ↔ | 1 1 | 0 2 | 0 3 | 1 a ), and ROT 11 ( | 1 1 | 1 2 | 1 3 | 0 a ↔ | 1 1 | 1 2 | 0 3 | 1 a ), respectively....The quantum entangling operation based on the SR can also help us to complete a single-step controlled-controlled phase (cc-phase) quantum gate on the three charge **qubits** q 1 , q 2 , and q 3 by using the system shown in Fig. fig2 except for the resonator R b . Here, q 1 and q 2 act as the control **qubits**, and q 3 is the target **qubit**. The initial state of this system is prepared as | Φ 0 = 1 2 2 | 0 1 | 0 2 | 0 3 + | 0 1 | 0 2 | 1 3 + | 0 1 | 1 2 | 0 3 + | 0 1 | 1 2 | 1 3 + | 1 1 | 0 2 | 0 3 + | 1 1 | 0 2 | 1 3 + | 1 1 | 1 2 | 0 3 + | 1 1 | 1 2 | 1 3 | 0 a . In this system, both q 1 and q 2 are in the quasi-dispersive regime with R a , and the transition **frequency** of q 3 is adjusted to be equivalent to that of R a when q 1 and q 2 are in their ground states. The QSD transition **frequency** on R a becomes...(color online) Sketch of a coplanar geometry for the circuit QED with three superconducting **qubits**. **Qubits** are placed around the maxima of the electrical field amplitude of R a and R b (not drawn in this figure), and the distance between them is large enough so that there is no direct interaction between them. The fundamental **frequencies** of resonators are ω r j / 2 π ( j = a , b ), the **frequencies** of the **qubits** are ω q i / 2 π ( i = 1 , 2 , 3 ), and they are capacitively coupled to the resonators. The coupling strengths between them are g i j / 2 π . We can use the control line (not drawn here) to afford the flux to tune the transition **frequencies** of the **qubits**....(color online) (a) The **qubit**-state-dependent resonator transition, which means the **frequency** shift of the resonator transition δ r arises from the state ( | 0 q or | 1 q ) of the **qubit**. (b) The number-state-dependent **qubit** transition, which means the **frequency** shift δ q takes place on the **qubit** due to the photon number n = 1 or 0 in the resonator in the dispersive regime....in which we neglect the direct interaction between the two **qubits** (i.e., q 1 and q 2 ), shown in Fig. fig2. Here σ i + = | 1 i 0 | is the creation operator of q i ( i = 1 , 2 ). g i is the coupling strength between q i and R a . The parameters are chosen to make q 1 interact with R a in the quasi-dispersive regime. That is, the transition **frequency** of R a is determined by the state of q 1 . By taking a proper transition **frequency** of q 2 (which equals to the transition **frequency** of R a when q 1 is in the state | 0 1 ), one can realize the quantum Rabi **oscillation** (ROT) ROT 0 : | 0 1 | 1 2 | 0 a ↔ | 0 1 | 0 2 | 1 a , while ROT 1 : | 1 1 | 1 2 | 0 a ↔ | 1 1 | 0 2 | 1 a occurs with a small probability as q 2 detunes with R a when q 1 is in the state | 1 1 . Here the Fock state | n a represents the photon number n in R a ( n = 0 , 1 ). | 0 i and | 1 i are the ground and the first excited states of q i , respectively....Eq.( stark) means the NSD **qubit** transition and Eq.( kerr) means the QSD resonator transition, shown in Fig. fig1(a) and (b), respectively....(color online) Simulated outcomes for the maximum amplitude value of the expectation about the quantum Rabi **oscillation** varying with the coupling strength g 2 and the **frequency** of the second **qubit** ω 2 . (a) The outcomes for ROT 0 : | 0 1 | 1 2 | 0 a ↔ | 0 1 | 0 2 | 1 a . (b) The outcomes for ROT 1 : | 1 1 | 1 2 | 0 a ↔ | 1 1 | 0 2 | 1 a . Here the parameters of the resonator and the first **qubit** q 1 are taken as ω a / 2 π = 6.0 GHz, ω q 1 / 2 π = 7.0 GHz, and g 1 / 2 π = 0.2 GHz....and it is shown in Fig. fig4 (a). In the simulation of our SR, we choose the reasonable parameters by considering the energy level structure of a charge **qubit**, according to Ref. . Here ω r a / 2 π = 6.0 GHz. The transition **frequency** of two **qubits** between and are chosen as ω 0 , 1 ; 1 / 2 π = E 1 ; 1 - E 0 ; 1 = 5.0 GHz, ω 1 , 2 ; 1 / 2 π = E 2 ; 1 - E 1 ; 1 = 6.2 GHz, ω 0 , 1 ; 2 / 2 π = E 1 ; 2 - E 0 ; 2 = 6.035 GHz, and ω 1 , 2 ; 2 / 2 π = E 2 ; 2 - E 1 ; 2 = 7.335 GHz. Here E i ; q is the energy for the level i of the **qubit** q , and σ i , i ' ; q + ≡ i q i ' . g i , j ; q is the coupling strength between the resonator R a and the **qubit** q in the transition between the energy levels | i q and | j q ( i = 0 , 1 , j = 1 , 2 , and q = 1 , 2 ). For convenience, we take the coupling strengths as g 0 , 1 ; 1 / 2 π = g 1 , 2 ; 1 / 2 π = 0.2 GHz and g 0 , 1 ; 2 / 2 π = g 1 , 2 ; 2 / 2 π = 0.0488 GHz....We numerically simulate the maximal expectation values (MAEVs) of ROT 0 and ROT 1 based on the Hamiltonian H 2 q , shown in Fig. fig3(a) and (b), respectively. Here, the expectation value is defined as | ψ | e - i H 2 q t / ℏ | ψ 0 | 2 . | ψ 0 and | ψ are the initial and the final states of a quantum Rabi **oscillation**, respectively. The MAEV s vary with the transition **frequency** ω 2 and the coupling strength g 2 . It is obvious that the amplified QSD resonator transition can generate a selective resonance (SR) when the coupling strength g 2 is small enough....We present a fast quantum entangling operation on superconducting **qubits** assisted by a resonator in the quasi-dispersive regime with a new effect --- the selective resonance coming from the amplified **qubit**-state-dependent resonator transition **frequency** and the tunable period relation between a wanted quantum Rabi **oscillation** and an unwanted one. This operation does not require any kind of drive fields and the interaction between **qubits**. More interesting, the non-computational third excitation states of the charge **qubits** can play an important role in shortening largely the operation time of the entangling gates. All those features provide an effective way to realize much faster quantum entangling gates on superconducting **qubits** than previous proposals. ... We present a fast quantum entangling operation on superconducting **qubits** assisted by a resonator in the quasi-dispersive regime with a new effect --- the selective resonance coming from the amplified **qubit**-state-dependent resonator transition **frequency** and the tunable period relation between a wanted quantum Rabi **oscillation** and an unwanted one. This operation does not require any kind of drive fields and the interaction between **qubits**. More interesting, the non-computational third excitation states of the charge **qubits** can play an important role in shortening largely the operation time of the entangling gates. All those features provide an effective way to realize much faster quantum entangling gates on superconducting **qubits** than previous proposals.

Data types:

Contributors: Hausinger, Johannes, Grifoni, Milena

Date: 2010-07-30

Population difference for zero static bias. Further parameters are Δ / Ω = 0.5 , ℏ β Ω = 10 and g / Ω = 1.0 . The adiabatic approximation and VVP are compared to numerical results. The first one only covers the longscale dynamics, while VVP also returns the fast **oscillations**. With increasing time small differences between numerical results and VVP become more pronounced. Fig::P_e=0_D=0.5_g=1.0...As a first case, we consider in Fig. Fig::PF_e=Sqrt0.5_D=Sqrt0.5_g=1.0 a weakly biased **qubit** ( ε / Ω = 0.5 ) being at resonance with the **oscillator** ( Δ b = Ω ). For a coupling strength of g / Ω = 1.0 , we find a good agreement between the numerics and VVP. The adiabatic approximation, however, conveys a slightly different picture: Looking at the time evolution it reveals collapse and rebirth of **oscillations** after a certain interval. This feature does not survive for the exact dynamics. Like in the unbiased case, the adiabatic approximation gives only the first group of **frequencies** between the quasidegenerate subspaces, and thus yields a wrong picture of the dynamics. In order to cover the higher **frequency** groups, we need again to go to higher-order corrections by using VVP. For the derivation of our results we assumed that ε is a multiple of the **oscillator** **frequency** Ω , ε = l Ω . In this case we found that the levels E ↓ , j 0 and E ↑ , j + l 0 form a degenerate doublet, which dominates the long-scale dynamics through the dressed **oscillations** **frequency** Ω j l . For l being not an integer those doublets cannot be identified unambiguously anymore. For instance, we examine the case ε / Ω = 1.5 in Fig. Fig::PF_e=1.5_D=0.5_g=1.0. Here, it is not clear which levels should be gathered into one subspace: j and j + 1 or j and j + 2 . Both the dressed **oscillation** **frequencies** Ω j 1 and Ω j 2 influence the longtime dynamics....Fourier transform of the population difference in Fig. Fig::P_e=0_D=0.5_g=1.0. The left-hand graph shows the whole **frequency** range. The lowest **frequency** peaks originate from transitions between levels of a degenerate subspace and are determined through the dressed **oscillation** **frequency** Ω j 0 . Numerical calculations and VVP predict group of peaks located around ν / Ω = 0 , 1.0 , 2.0 , 3.0 . The first group at ν / Ω = 0 is shown in the middle graph. One can identify **frequencies** Ω 0 0 and Ω 2 0 , which fall together, and Ω 1 0 . The small peak comes from the **frequency** Ω 3 0 . This first gr...We examine a two-level system coupled to a quantum **oscillator**, typically representing experiments in cavity and circuit quantum electrodynamics. We show how such a system can be treated analytically in the ultrastrong coupling limit, where the ratio $g/\Omega$ between coupling strength and **oscillator** **frequency** approaches unity and goes beyond. In this regime the Jaynes-Cummings model is known to fail, because counter-rotating terms have to be taken into account. By using Van Vleck perturbation theory to higher orders in the **qubit** tunneling matrix element $\Delta$ we are able to enlarge the regime of applicability of existing analytical treatments, including in particular also the finite bias case. We present a detailed discussion on the energy spectrum of the system and on the dynamics of the **qubit** for an **oscillator** at low temperature. We consider the coupling strength $g$ to all orders, and the validity of our approach is even enhanced in the ultrastrong coupling regime. Looking at the Fourier spectrum of the population difference, we find that many **frequencies** are contributing to the dynamics. They are gathered into groups whose spacing depends on the **qubit**-**oscillator** detuning. Furthermore, the dynamics is not governed anymore by a vacuum Rabi splitting which scales linearly with $g$, but by a non-trivial dressing of the tunneling matrix element, which can be used to suppress specific **frequencies** through a variation of the coupling....Population difference and Fourier spectrum for a biased **qubit** ( ε / Ω = 0.5 ) at resonance with the **oscillator** Δ b = Ω in the ultrastrong coupling regime ( g / Ω = 1.0 ). Concerning the time evolution VVP agrees well with numerical results. Only for long time weak dephasing occurs. The inset in the left-hand figure shows the adiabatic approximation only. It exhibits death and revival of **oscillations** which are not confirmed by the numerics. For the Fourier spectrum, VVP covers the various **frequency** peaks, which are gathered into groups like for the unbiased case. The adiabatic approximation only returns the first group. Fig::PF_e=Sqrt0.5_D=Sqrt0.5_g=1.0...respectively. Concerning the population difference, we see a relatively good agreement between the numerical calculation and VVP for short timescales. In particular, VVP also correctly returns the small overlaid **oscillations**. For longer timescales, the two curves get out of phase. The adiabatic approximation only can reproduce the coarse-grained dynamics. The fast **oscillations** are completely missed. To understand this better, we turn our attention to the Fourier transform in Fig. Fig::F_e=0_D=0.5_g=1.0. There, we find several groups of **frequencies** located around ν / Ω = 0 , ν / Ω = 1.0 , ν / Ω = 2.0 and ν / Ω = 3.0 . This can be explained by considering the transition **frequencies** in more detail. We have from Eq. ( VVEnergies)...with ζ k , j l = 1 8 ε ↓ , k 2 - ε ↓ , j 2 + ε ↑ , j + l 2 - ε ↑ , k + l 2 being the second-order corrections. For zero bias, ε = 0 , the index l vanishes. The term k - j Ω determines to which group of peaks a **frequency** belongs and Ω j 0 its relative position within this group. The latter has Δ as an upper bound, so that the range over which the peaks are spread within a group increases with Δ . The dynamics is dominated by the peaks belonging to transitions between the same subspace k - j = 0 , while the next group with k - j = 1 yields already faster **oscillations**. To each group belong theoretically infinite many peaks. However, under the low temperature assumption only those with a small **oscillator** number play a role. For the used parameter regime, the adiabatic approximation does not take into account the connections between different manifolds. It therefore covers only the first group of peaks with k - j = 0 , providing the long-scale dynamics. For ε = 0 , the dominating **frequencies** in this first group are given by Ω 0 0 = | Δ e - α / 2 | , Ω 1 0 = | Δ 1 - α e - α / 2 | and Ω 2 0 = | Δ L 2 0 α e - α / 2 | , where Ω 0 0 and Ω 2 0 coincide. A small peak at Ω 3 0 = | Δ L 3 0 α e - α / 2 | can also be seen. Notice that for certain coupling strengths some peaks vanish; like, for example, choosing a coupling strength of g / Ω = 0.5 makes the peak at Ω 1 0 vanish completely, independently of Δ , and the Ω 0 0 and Ω 2 0 peaks split. The JCM yields two **oscillation** peaks determined by the Rabi splitting and fails completely to give the correct dynamics, see the left-hand graph in Fig. Fig::F_e=0_D=0.5_g=1.0. Now, we proceed to an even stronger coupling, g / Ω = 2.0 , where we also expect the adiabatic approximation to work better. From Fig. Fig::EnergyVSg_e=0_W=1_D=0.5 we noticed that at such a coupling strength the lowest energy levels are degenerate within a subspace. Only for **oscillator** numbers like j = 3 , we see that a small splitting arises. This splitting becomes larger for higher levels. Thus, only this and higher manifolds can give significant contributions to the long time dynamics; that is, they can yield low **frequency** peaks. Also the adiabatic approximation is expected to work better for such strong couplings . And indeed by looking at Figs. Fig::P_e=0_D=0.5_g=2.0 and Fig::F_e=0_D=0.5_g=2.0, we notice that both the adiabatic approximation and VVP agree quite well with the numerics. Especially the first group of Fourier peaks in Fig. Fig::F_e=0_D=0.5_g=2.0 is also covered almost correctly by the adiabatic approximation. The first manifolds we can identify with those peaks are the ones with j = 3 and j = 4 . This is a clear indication that even at low temperatures higher **oscillator** quanta are involved due to the large coupling strength. Also **frequencies** coming from transitions between the energy levels from neighboring manifolds are shown enlarged in Fig. Fig::F_e=0_D=0.5_g=2.0. The adiabatic approximation and VVP can cover the main structure of the peaks involved there, while the former shows stronger deviations. If we go to higher values Δ / Ω 1 , the peaks in the individual groups become more spread out in **frequency** space, and for the population difference dephasing already occurs at a shorter timescale. For Δ / Ω = 1 , at least VVP yields still acceptable results in Fourier space but gets fast out of phase for the population difference....Fourier spectrum of the population difference in Fig. Fig::P_e=0_D=0.5_g=2.0. In the left-hand graph a large **frequency** range is covered. Peaks are located around ν / Ω = 0 , 1.0 , 2.0 , 3.0 etc. Even the adiabatic approximation exhibits the higher **frequencies**. The upper right-hand graph shows the first group close to ν / Ω = 0 . The two main peaks come from Ω 3 0 and Ω 4 0 and higher degenerate manifolds. **Frequencies** from lower manifolds contribute to the peak at zero. The adiabatic approximation and VVP agree well with the numerics. The lower right-hand graph shows the second group of peaks around ν / Ω = 1.0 . This group is also predicted by the adiabatic approximation and VVP, but they do not fully return the detailed structure of the numerics. Interestingly, there is no peak exactly at ν / Ω = 1.0 indicating no nearest-neighbor transition between the low degenerate levels. Fig::F_e=0_D=0.5_g=2.0...Population difference and Fourier spectrum for ε / Ω = 1.5 , Δ / Ω = 0.5 and g / Ω = 1.0 . Van Vleck perturbation theory is confirmed by numerical calculations, while results obtained from the adiabatic approximation deviate strongly. In Fourier space, we find pairs of **frequency** peaks coming from the two dressed **oscillation** **frequencies** Ω j 1 and Ω j 2 . The spacings in between those pairs is about 0.5 Ω . The adiabatic approximation only returns one of those dressed **frequencies** in the first pair. Fig::PF_e=1.5_D=0.5_g=1.0...Fourier transform of the population difference in Fig. Fig::P_e=0_D=0.5_g=1.0. The left-hand graph shows the whole **frequency** range. The lowest **frequency** peaks originate from transitions between levels of a degenerate subspace and are determined through the dressed **oscillation** **frequency** Ω j 0 . Numerical calculations and VVP predict group of peaks located around ν / Ω = 0 , 1.0 , 2.0 , 3.0 . The first group at ν / Ω = 0 is shown in the middle graph. One can identify **frequencies** Ω 0 0 and Ω 2 0 , which fall together, and Ω 1 0 . The small peak comes from the **frequency** Ω 3 0 . This first group of peaks is also covered by the adiabatic approximation. The other groups come from transitions between different manifolds. The adiabatic approximation does not take them into account, while VVP does. A blow-up of the peaks coming from transitions between neighboring manifolds is given in the right-hand graph. In the left-hand graph additionally the Jaynes-Cummings peaks are shown, which, however, fail completely. Fig::F_e=0_D=0.5_g=1.0 ... We examine a two-level system coupled to a quantum **oscillator**, typically representing experiments in cavity and circuit quantum electrodynamics. We show how such a system can be treated analytically in the ultrastrong coupling limit, where the ratio $g/\Omega$ between coupling strength and **oscillator** **frequency** approaches unity and goes beyond. In this regime the Jaynes-Cummings model is known to fail, because counter-rotating terms have to be taken into account. By using Van Vleck perturbation theory to higher orders in the **qubit** tunneling matrix element $\Delta$ we are able to enlarge the regime of applicability of existing analytical treatments, including in particular also the finite bias case. We present a detailed discussion on the energy spectrum of the system and on the dynamics of the **qubit** for an **oscillator** at low temperature. We consider the coupling strength $g$ to all orders, and the validity of our approach is even enhanced in the ultrastrong coupling regime. Looking at the Fourier spectrum of the population difference, we find that many **frequencies** are contributing to the dynamics. They are gathered into groups whose spacing depends on the **qubit**-**oscillator** detuning. Furthermore, the dynamics is not governed anymore by a vacuum Rabi splitting which scales linearly with $g$, but by a non-trivial dressing of the tunneling matrix element, which can be used to suppress specific **frequencies** through a variation of the coupling.

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Contributors: Du, Lingjie, Yu, Yang

Date: 2010-12-13

(Color online). (a) and (b). Schematic energy diagram of Rabi **oscillation** induced interference. (a) describes the transition from state | 1 to | 0 . (b) describes the transition from state | 0 to | 1 . (c), (d) and (e). The interference pattern of population in state | 0 obtained from Eqs. (45), (46), and (47), respectively. The parameters used here are ω ~ / 2 π = 2 GHz, Ã / ω ~ = 0.9 , Γ 01 / 2 π = 0.000008 GHz and the temperature is 20 mK. Other parameters of the **qubit** are identical with Fig. 4 (a)....(Color online). Calculated final **qubit** population versus energy detuning and microwave amplitude. (a). The stationary interference pattern in the weak relaxation situation. The parameters we used are the driving **frequency** ω / 2 π = 0.6 GHz, the dephasing rate Γ 2 / 2 π = 0.06 GHz, the couple tunneling Δ / 2 π = 0.013 GHz, φ 2 α = 0.0002 GHz, the temperature is 20 mK, and the characteristic **frequency** ω c / 2 π = 0.05 GHz. The periodical patterns of RII can be seen, although not clear. (b). The stationary interference pattern in the strong relaxation situation with α φ 2 = 0.02 GHz and ω c / 2 π = 0.05 GHz. Since the relaxation strength is stronger, the periodical interference patterns are more notable. (c). The stationary interference pattern in the weak relaxation situation with α φ 2 = 0.000002 GHz and ω c / 2 π = 6 GHz. (d). The stationary interference pattern in the strong relaxation situation with α φ 2 = 0.0002 GHz and ω c / 2 π = 6 GHz. (e). The unsaturated interference pattern in the weak relaxation situation. The system dynamics time t = 0.5 μ s. The characteristic **frequency** ω c / 2 π = 0.05 GHz. α φ 2 = 0.0002 GHz. (f). The unsaturated interference pattern in the weak relaxation situation. The system dynamics time t = 0.5 μ s. The characteristic **frequency** ω c / 2 π = 6 GHz. α φ 2 = 0.000002 GHz. (g). The unsaturated interference pattern in the strong relaxation situation. The system dynamics time t = 0.5 μ s. The characteristic **frequency** ω c / 2 π = 0.05 GHz. α φ 2 = 0.02 GHz. (h). The unsaturated interference pattern in the strong relaxation situation. The dynamics time t = 0.5 μ s. The characteristic **frequency** ω c / 2 π = 6 GHz. α φ 2 = 0.0002 GHz. The other parameters used in these Figures are the same with those in Fig. 4 (a)....(Color online). (a). Schematic energy diagram of a flux **qubit**. The dotted curve represents the strong driving field A cos ω t . The field through the tunnel coupling Δ forms a LZS interference, exchanging photons with the **qubit**. (b). Quantum tunnel coupling exists between states | 0 and | 1 . The interaction between a **qubit** and an electromagnetic system (such as the environment bath or a single-mode electromagnetic field) would form new couplings between the two states....(Color online). The stationary population of relaxation induced interference. The pattern is obtained from Eq. (34). (a). The characteristic **frequency** ω c / 2 π = 0.05 GHz with the temperature 20 mK. Features of population inversion and periodical modulation are notable. (b). The characteristic **frequency** ω c / 2 π = 6 GHz with the temperature 20 mK. (c). The characteristic **frequency** ω c / 2 π = 0.05 GHz with the temperature 2 × 10 -5 mK. In above figures, the driving **frequency** ω / 2 π = 0.6 GHz....(Color online). (a). Schematic energy diagram of a strongly driven flux **qubit** interacting with a weak single-mode field. The green solid curve represents the weak field, forming effective coupling between states | 0 and | 1 ....We study electromagnetically induced interference at superconducting **qubits**. The interaction between **qubits** and electromagnetic fields can provide additional coupling channels to **qubit** states, leading to quantum interference in a microwave driven **qubit**. In particular, the interwell relaxation or Rabi **oscillation**, resulting respectively from the multi- or single-mode interaction, can induce effective crossovers. The environment is modeled by a multi-mode thermal bath, generating the interwell relaxation. Relaxation induced interference, independent of the tunnel coupling, provides deeper understanding to the interaction between the **qubits** and their environment. It also supplies a useful tool to characterize the relaxation strength as well as the characteristic **frequency** of the bath. In addition, we demonstrate the relaxation can generate population inversion in a strongly driving two-level system. On the other hand, different from Rabi **oscillation**, Rabi **oscillation** induced interference involves more complicated and modulated photon exchange thus offers an alternative means to manipulate the **qubit**, with more controllable parameters including the strength and position of the tunnel coupling. It also provides a testing ground for exploring nonlinear quantum phenomena and quantum state manipulation, in not only the flux **qubit** but also the systems with no crossover structure, e.g. phase **qubits**. ... We study electromagnetically induced interference at superconducting **qubits**. The interaction between **qubits** and electromagnetic fields can provide additional coupling channels to **qubit** states, leading to quantum interference in a microwave driven **qubit**. In particular, the interwell relaxation or Rabi **oscillation**, resulting respectively from the multi- or single-mode interaction, can induce effective crossovers. The environment is modeled by a multi-mode thermal bath, generating the interwell relaxation. Relaxation induced interference, independent of the tunnel coupling, provides deeper understanding to the interaction between the **qubits** and their environment. It also supplies a useful tool to characterize the relaxation strength as well as the characteristic **frequency** of the bath. In addition, we demonstrate the relaxation can generate population inversion in a strongly driving two-level system. On the other hand, different from Rabi **oscillation**, Rabi **oscillation** induced interference involves more complicated and modulated photon exchange thus offers an alternative means to manipulate the **qubit**, with more controllable parameters including the strength and position of the tunnel coupling. It also provides a testing ground for exploring nonlinear quantum phenomena and quantum state manipulation, in not only the flux **qubit** but also the systems with no crossover structure, e.g. phase **qubits**.

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